Equations of Mathematical Diffraction Theory

EQUATIONS OF MATHEMATICAL DIFFRACTION THEORY © 2005 by CRC Press LLC Differential and Integral Equations and Their Ap...

Author: Mezhlum A. Sumbatyan | Antonio Scalia

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EQUATIONS OF MATHEMATICAL DIFFRACTION THEORY

© 2005 by CRC Press LLC

Differential and Integral Equations and Their Applications A series edited by: A.D. Polyanin Institute for Problems in Mechanics, Moscow, Russia Volume 1 Handbook of First Order Partial Differential Equations A.D. Polyanin, V.F. Zaitsev and A. Moussiaux Volume 2 Group-Theoretic Methods in Mechanics and Applied Mathematics D.M. Klimov and V. Ph. Zhuravlev Volume 3 Quantization Methods in the Theory of Differential Equations V.E. Nazaikinskii, B.-W. Schulze and B. Yu. Sternin Volume 4 Hypersingular Integral Equations and Their Applications I.K. Lifanov, L.N. Poltavskii and G.M. Vainikko Volume 5 Equations of Mathematical Diffraction Theory Mezhlum A. Sumbatyan and Antonio Scalia


EQUATIONS OF MATHEMATICAL DIFFRACTION THEORY Mezhlum A. Sumbatyan Antonio Scalia

CHAPMAN & HALL/CRC A CRC Press Company Boca Raton London New York Washington, D.C.


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Library of Congress Cataloging-in-Publication Data Sumbatyan, Mezhlum A. Equations of mathematical diffraction theory / Mezhlum A. Sumbatyan, Antonio Scalia. p. cm. Includes bibliographical references and index. ISBN 0-415-30849-6 (alk. paper) 1. Diffraction—Mathematics. I. Scalia, A. II. Title. QC415.S95 2004 535'.42'0151—dc22

2004051957

This book contains information obtained from authentic and highly regarded sources. Reprinted material is quoted with permission, and sources are indicated. A wide variety of references are listed. Reasonable efforts have been made to publish reliable data and information, but the author and the publisher cannot assume responsibility for the validity of all materials or for the consequences of their use. Neither this book nor any part may be reproduced or transmitted in any form or by any means, electronic or mechanical, including photocopying, microfilming, and recording, or by any information storage or retrieval system, without prior permission in writing from the publisher. The consent of CRC Press LLC does not extend to copying for general distribution, for promotion, for creating new works, or for resale. Specific permission must be obtained in writing from CRC Press LLC for such copying. Direct all inquiries to CRC Press LLC, 2000 N.W. Corporate Blvd., Boca Raton, Florida 33431. Trademark Notice: Product or corporate names may be trademarks or registered trademarks, and are used only for identification and explanation, without intent to infringe.

Visit the CRC Press Web site at www.crcpress.com © 2005 by CRC Press LLC No claim to original U.S. Government works International Standard Book Number 0-415-30849-6 Library of Congress Card Number 2004051957 Printed in the United States of America 1 2 3 4 5 6 7 8 9 0 Printed on acid-free paper


PREFACE The connection between heuristic and strictly formal methods is seemingly one of the most interesting and debatable questions in modern mathematics. Each of these two different approaches, whose foundations were laid by Socrates and Aristotle, respectively, and in the new history are reflected in discussions and written papers by Descartes, Leibnitz and Bacon, has its own intrinsic merits and restrictions. Moreover, a large number of discoveries in science were made owing to a combination of strict and heuristic methods of investigation. Apparently, one of the brightest examples in modern mathematical physics is diffraction theory, where the combination of the two approaches would lead so efficiently to such impressive results. Many important and interesting solutions and even some classical theories appeared from heuristic ideas, and the most impressive example was given by Kirchhoff’s physical diffraction theory, which is based upon a clear “light and shadow” concept for diffracted wave fields. Subsequently, many of these heuristic results were rigorously substantiated and proved as theorems. On the other hand, unsuccessful attempts to prove some other heuristic ideas caused significant progress in the development of formal methods that yielded correct solutions, different sometimes from those prompted by someone’s intuition. The above specific features have affected the style of presentation of the book. Each section deals with a discussion of heuristic ideas, which as a rule are substantiated (or disproved) with the use of rigorous mathematical methods. Due to limited volume of the book, at some places we give only a brief sketch of the substantiation, referring the reader to the original literature for more details. One more specific feature of the presented material is connected with the rapid progress in computer technology over the last 20 years, which has significantly changed our viewpoint on what could be accepted as efficient methods of investigation. Only recently, expansion of unknown functions into series in terms of special functions, when a problem reduced to infinite system of linear algebraic equations with respect to coefficients of the expansion, was regarded as a standard method. Such a “semi-analytical” approach was efficient 15–20 years ago, when the evaluation of regularity of the obtained infinite systems seemed to be very important, since this could guarantee accuracy of a solution by retaining only few first equations, which was acceptable for first-generations computers. Nowadays, when there is not much difference between 10 × 10 and 500 × 500 systems even for home personal computers, such a viewpoint looks archaic, since the time required to convert the system to a form appropriate for “fast computations” is much greater than that for “slow computations” based on modern direct numerical methods like boundary element method and finite element method. Apparently, it should be agreed that in the cases where direct numerical techniques provide reliable results in an acceptable computational time, they should be regarded as most efficient for the problem in question. It is very important to recognize the cases where one has a priori to reject direct numerical methods. These are listed below. 1◦ . Problems where an exact analytical solution or a good approximation to it can be obtained. Diffraction theory shows many examples of this kind. 2◦ . Studying dynamic processes with high frequencies. Here, one has to take at least


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10 nodes per wavelength to obtain reliable results by any direct numerical method. As the wavelength decreases (i.e., the frequency increases) within a given frequency range, the total number of nodes increases very rapidly, which results in too large algebraic systems. An impressive example is given by room acoustics. Suppose a sound wave of frequency f = 2 kHz, whose wavelength in air is 17 cm, propagates in a 17-m long room. For reasonable numerical accuracy, one should hence take at least 1000 nodes along the room length. If the room has a width of 8 m and a height of 5.1 m, one has to consider 1000 × 500 × 300 ≈ 108 finite-element nodes and perform complex-valued arithmetic. This cannot be implemented even on the most powerful super computers. Here, a reasonable criterion for acceptability of a numerical approach is its implementability on a PC or similar computer. So, obtaining solutions to such high-frequency problems in exact formulation by direct numerical techniques does not seem to be feasible in the visible future. 3◦ . Studying phenomena of complex qualitative nature. Since direct numerical methods provide only numbers, which are usually tabulated and plotted, it is often very difficult to extract such complex qualitative effects from numerous tables and graphs. Instead, it is preferable to construct an approximate analytical solution, from which qualitative effects may be extracted explicitly. 4◦ . Cases where an exact analytical solution has been obtained but its representation is inapplicable to practice for specific calculations. An example of this kind is considered in Section 6.1. In such interesting cases, one should look for an alternative approach, which is often the construction of an approximate solution that would be more appropriate for fast computations than the exact analytical solution obtained. The above situations are not widespread but, when met, are very difficult to cope with efficiently, especially if the researcher does not have sufficient experience in tackling them. This prompted us to conclude each section with a special subsection titled “Helpful Remarks,” which may help the reader to build up his or her own less formal conception and allow the creation of a more complete picture of the issue under consideration. Note that the application of numerical methods in regular cases is well described in the classical literature. For this reason, we only consider numerical methods for some irregular operator problems; see Chapter 9. To summarize, the main purpose of the present book is to show the close connection between heuristic and rigorous methods in mathematical diffraction theory. We focus on differential and integral equations that can easily be utilized in practical applications. Such an approach is accounted for by the choice of our potential readers. The book presents clear and elegant methods and is aimed at graduate and post-graduate students, so that they could quickly examine the state of the art in a specific field of interest. At the same time, researchers with considerable expertise in dealing with diffraction theory will hopefully discover that the time of clear explicit solutions in unsolved complex problems has not passed yet—this is demonstrated by the authors’ original results in Sections 4.5, 4.6, 5.4–5.7, and 6.3–6.6 as well as in many sections of Chapters 7–9. Furthermore, we hope that an experienced reader will be able to discover for him- or herself new helpful methods, both analytical and numerical. The reader will see in what follows that we prefer to rely upon classical results of the founders of modern science unlike a rather widespread (mistaken) point of view that only very complicated recent “abstract” theories can provide further progress in contemporary science. We strongly recommend the younger reader to operate with classical mathematical theories, and the present book will demonstrate that the fruitful ideas of Hilbert, Cauchy, Fourier, Abel, Poisson, Weyl, Riemann, Green, Kirchhoff, Rayleigh, Helmholtz, Neumann, and others can guide the reader very efficiently around present-day problems in diffraction theory. It should also be stressed that we tried to avoid too formal presentation, since we


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believe that wielding thorough knowledge in any mathematical theory implies applying it effectively and successfully to practice rather than operating with the formal apparatus of the theory. Due to its limited volume, any monograph cannot cover all important questions, and the present book is no exception. For example, the reader will not find here transient problems at all. The presentation is confined to boundary problems for elliptic operators only, and only those with constant coefficients (except Section 3.6). Moreover, the main focus is on methods that provide solutions without too cumbersome mathematical manipulations. For example, the reader will not find the structure of the wave field in the “semi-shadow” zone in diffraction by convex obstacles, and in the method of “edge waves” in diffraction from linear segments, the reader will only find the leading high-frequency asymptotic term, which is constructed by a simple and elegant technique. The sections, displayed formulas, and figures are enumerated independently within each chapter with the chapter number in front. The book is intended for the reader familiar with fundamentals of real, complex-valued, and functional analysis within a standard course on calculus in the first three years of any university program of mathematical, physical, or engineering departments. The style and content of this book have been influenced by the authors’ friends, teachers, and colleagues, Alexander Vatulyan (Rostov State University, Russia), Mauro Fabrizio (University of Bologna, Italy), and Dorin Iesan (University of Iasi, Romania). The authors are grateful to Alexander Manzhirov and Alexei Zhurov for their helpful discussions and comments. The first author is thankful to his wife, Angela Sumbatyan, and to his daughters, Laura, Carina, and Angelica, who assisted and inspired him in the writing of the book. M. A. Sumbatyan A. Scalia


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AUTHORS Mezhlum A. Sumbatyan, Ph.D., D.Sc., is a noted scientist in the field of wave dynamics, diffraction theory, and ill-posed problems. Mezhlum Sumbatyan graduated from the Faculty of Mechanics and Mathematics, Rostov State University, Russia, in 1969 and received his Ph.D. degree in 1980 at the Institute for Problems in Mechanics, Russian Academy of Sciences, Moscow. His Ph.D. thesis was devoted to asymptotic methods for solving integral equations arising in some problems of mechanics and acoustics with mixed boundary conditions. In 1995, Professor Sumbatyan received his Doctor of Sciences degree; his D.Sc. thesis was dedicated to direct and inverse problems of diffraction theory, with applications to reconstruction of obstacles by ultrasonic techniques. In 1985–2000, Mezhlum Sumbatyan worked in the Research Institute of Mechanics and Applied Mathematics of the Rostov State University. Since 2001, Professor Sumbatyan has been a member of the staff of the Faculty of Mechanics and Mathematics, Rostov State University. Professor Sumbatyan has made important contributions to new methods in the theory and analytical methods applied to direct and inverse diffraction problems. He is an author of more than 120 scientific publications, a member of the Russian Acoustical Society, and a member of the Acoustical Society of America. Address: Faculty of Mechanics and Mathematics Zorge Street 5 344090 Rostov-on-Don, Russia

E-mail: [email protected]

Antonio Scalia, Ph.D., D.Sc., is a prominent scientist in the field of mathematical physics. Antonio Scalia graduated from the Faculty of Mathematics, Physics and Natural Sciences, University of Catania, Italy, in 1972 and received his Ph.D. degree in 1980 at the University of Catania. His thesis was devoted to solvability and uniqueness in mathematical formulation of mechanical and acoustical problems for media with micro-structure. Since 1989, Professor Scalia has been a member of the staff of the Department of Mathematics and Informatics at the University of Catania as an Associate Professor, and since 2001, as a Full Professor. Professor Scalia has great expertise in the theory of viscoelasticity, electromagnetism, and acoustics. He has significant achievements in the mixture theory in solids, with applications to wave processes in elastic and acoustic media. Recently, within the theory of porous media, he devised a new approach to the analysis of wave propagation in elastic solids with pores. He explicitly studied a characteristic equation of the porous composite media applied to the layered geometry. Professor Scalia is an author of nearly 100 scientific publications. Address: Dipartimento di Matematica e Informatica Universitá di Catania, Viale A. Doria n. 6 95125 Catania, Italy

E-mail: [email protected]


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CONTENTS Preface Authors Contents 1. Some Preliminaries from Analysis and the Theory of Wave Processes 1.1. Fourier Transform, Line Integrals of Complex-Valued Integrands, and Series in Residues 1.2. Convolution Integral Equations and the Wiener–Hopf Method 1.3. Summation of Divergent Series and Integrals 1.4. Asymptotic Estimates of Integrals 1.5. Fredholm Theory for Integral Equations of the Second Kind 1.6. Fredholm Integral Equations of the First Kind 1.7. Singular Integral Equations with a Cauchy-Type Singularity in the Kernel 1.8. Hyper-Singular Integrals and Integral Equations 1.9. Governing Equations of Hydroaeroacoustics, Electromagnetic Theory, and Dynamic Elasticity 2. Integral Equations of Diffraction Theory for Obstacles in Unbounded Medium 2.1. Properties of the Potentials of Single and Double Layers 2.2. Basic Integral Equations of the Diffraction Theory 2.3. Properties of Integral Operators of Diffraction Theory: General Case and Low Frequencies 2.4. Full Low-Frequency Solution for Spherical Obstacle 2.5. Application: Scattering Diagram for Obstacles of Canonical Shape 2.6. Asymptotic Character of the Kirchhoff Physical Diffraction Theory 3. Wave Fields in a Layer of Constant Thickness 3.1. Wave Operator in Acoustic Layer: Mode Expansion, Homogeneous and Inhomogeneous Waves 3.2. Principles of Selection of Unique Solution in Unbounded Domain 3.3. Waves in Elastic Layer 3.4. Generalized Riemann’s Zeta Function and Summation of Some Oscillating Series 3.5. Application: Efficient Calculation of Wave Fields in a Layer of Constant Thickness 3.6. Waves in the Stratified Half-Plane


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4. Analytical Methods for Simply Connected Bounded Domains 4.1. General Spectral Properties of the Interior Problem for Laplacian 4.2. Explicit Formulas for Eigenfrequencies of Round Disc 4.3. Some Variational Principles for Eigenvalues 4.4. Weyl–Carleman Theory of Asymptotic Distribution of Large Eigenvalues 4.5. Exact Explicit Results for Some Polygons 4.6. Explicit Analytical Results for Some Polyhedra 5. Integral Equations in Diffraction by Linear Obstacles 5.1. Integral Operators in Diffraction by Linear Screen and by a Gap in the Screen 5.2. Operator Equation in Diffraction Problem on a Crack in Unbounded Elastic Medium 5.3. High-Frequency Asymptotics in Diffraction by Linear Obstacles in Unbounded Medium 5.4. High-Frequency Asymptotics for Diffraction by Linear Obstacles in Open Waveguides 5.5. High-Frequency Diffraction by a Linear Discontinuity in the Waveguide 5.6. Waves in Elastic Half-Space. Factorization of the Rayleigh Function 5.7. Integral Equation of the Mixed Boundary Value Problem for Elastic Layer 6. Short-Wave Asymptotic Methods on the Basis of Multiple Integrals 6.1. Schoch’s Method: Exact Representation of 3D Wave Fields by One-Dimensional Quadratures 6.2. High-Frequency Wave Fields in Elastic Half-Space 6.3. Asymptotic Nature of the Geometrical Diffraction Theory 6.4. High-Frequency Diffraction with Re-Reflections 6.5. Application: Examples of High-Frequency Multiple Diffraction 6.6. Application: Physical Diffraction Theory for Nonconvex Obstacles 6.7. Short-Wave Integral Operator in Diffraction by a Flaw in Elastic Medium 6.8. High-Frequency Asymptotics of Integral Operator in a Three-Dimensional Diffraction Theory 7. Inverse Problems of the Short-Wave Diffraction 7.1. Some Basic Results in a Local Differential Geometry of Smooth Convex Surfaces 7.2. Reducing Inverse Problem of the Short-Wave Diffraction to Minkowski Problem 7.3. Explicit Results for a Differential Operator of the 2D Inverse Problem 7.4. Exact Explicit Inversion of the Basic Operator in the Case of Axial Symmetry 7.5. Nonlinear Differential Operator of the Three-Dimensional Inverse Problem 7.6. Reconstruction of Nonconvex Obstacles in the High-Frequency Range: 2D Case 7.7. Reconstruction of Nonconvex Obstacles in the High-Frequency Range: 3D Case


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8. Ill-Posed Equations of Inverse Diffraction Problems for Arbitrary Boundary 8.1. Ill-Posed Problems for Operator Equations of the First Kind: General Properties 8.2. Regularization of Ill-Posed Problems with the Help of Smoothing Functional 8.3. Iterative Methods for Operator Equations of the First Kind 8.4. Comparison of Various Methods for Reconstruction of the Scatterer Geometry 8.5. General Inverse Diffraction Problem: Combination of Iterations and Smoothing 8.6. A Correct Treatment of Ill-Posed Boundary Equations in Acoustics of Closed Regions 8.7. Ill-Posed Method of Auxiliary Sources in Diffraction Theory 8.8. A Method of Global Random Search in Inverse Problems 8.9. Ill-Posed Problem on Reconstruction of Convex Hull of the Obstacle in Acoustic Medium 9. Numerical Methods for Irregular Operator Equations 9.1. Steepest Descent Method: Stability and Improvement of the Convergence 9.2. Galerkin Methods for Integral Equations of the First Kind with Weakly Singular Kernels 9.3. Integral Equations of the Physical Diffraction Theory in the Case of Nonconvex Obstacles 9.4. Numerical Methods in Singular Integral Equations with the Cauchy-Type Kernel 9.5. Numerical Methods for Hyper-Singular Integral Equations References


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Chapter 1

Some Preliminaries from Analysis and the Theory of Wave Processes 1.1. Fourier Transform, Line Integrals of Complex-Valued Integrands, and Series in Residues Let a function f (x) be integrable on the real axis: f (x) ∈ L 1 (–∞, ∞). Then its Fourier transform F (s ) is defined as Z ∞ F (s ) = f (x) eisx dx, F (s ) ∈ L1 (–∞, ∞), (1.1) –∞

and in the case when f (x) is continuous, the following inversion formula is valid: Z a Z ∞ 1 1 lim F (s ) e–isx ds , x ∈ (–∞, ∞). F (s ) e–isx ds = f (x) = 2π –∞ 2π a→∞ –a

(1.2)

The function f (x) will be called an original and the function F (s ) the Fourier image of f (x). The fact that the original and the image are related by formulas (1.1) and (1.2) will be denoted f (x) =⇒ F (s ). Many important and helpful properties of the Fourier transform are well known (e.g., see Titchmarsh, 1948; Bremermann, 1965). We will use only the following two of them: 1◦ . Fourier image of the derivative. Let f (x) =⇒ F (s ) and f (n) (x) ∈ L1 (–∞, ∞), then f (n) (x) =⇒ (–is )n F (s ).

(1.3)

2 . Fourier image of the convolution. Let f (x) ∈ L1 (–∞, ∞), g (x) ∈ L1 (–∞, ∞), and f (x) =⇒ F (s ), g (x) =⇒ G(s ). Then the convolution of f (x) and g (x) is given by Z ∞ h(x) = (f ∗g )(x) = f (ξ ) g (x – ξ ) dξ ∈ L1 (–∞, ∞), and h(x) =⇒ F (s ) G(s ). (1.4) ◦

–∞

The first property (1.3) can be obtained by the direct differentiation of Eq. (1.2), and the second property (1.4) follows from a change of variable when applying the Fourier transform to Eq. (1.4). The Fourier transform can also be defined for functions from the Hilbert space L 2 : f (x) ∈ L2 (–∞, ∞). The classical Plancherel theorem asserts the existence of a Fourier transform F (s ) (Wiener, 1934): Z ∞ Z a F (s ) = f (x) eisx ds = lim F (s ) eisx ds , x ∈ (–∞, ∞). (1.5) –∞

a→∞

–a

The convergence here is implied in the mean-square sense, i.e., as a convergence in L 2 . In this case, F (s ) ∈ L2 (–∞, ∞), and the inverse Fourier transform is valid in the same sense, Z ∞ Z a 1 1 F (s ) e–isx ds = lim F (s ) e–isx ds , x ∈ (–∞, ∞), (1.6) f (x) = 2π –∞ 2π a→∞ –a


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of mean-square convergence. For L2 functions f (x), g (x) ∈ L2 (–∞, ∞) the Parseval identity states that if f (x) =⇒ F (s ) and g (x) =⇒ G(s ), then Z

∞

f (x) g (x) dx =

–∞

Z

∞

F (s ) G(s ) ds ,

in particular,

–∞

Z

∞

|f (x)| dx = 2

–∞

Z

∞

|F (s )|2 ds ,

–∞

(1.7) where the bar over a symbol denotes a complex conjugate. The convolution theorem also remains valid in L2 . Let H (D ) denote a set of complex-valued analytic functions f (z ) of the complex variable z = Re(z ) + i Im(z ) defined over a domain D : f (z ) ∈ H (D ), z ∈ D . Recall that this implies that f (z ) is analytic and single-valued together with all its derivatives: f (n) (z ) ∈ H (D ), ∀n = 0, 1, 2, . . . (see Markushevich, 1963). Then the Cauchy theorem declares that the value of the line integral I( Γ ) =

Z

zB

f (z ) dz ,

z A , zB ∈ D ,

zA

(1.8)

along a curve Γ ⊂ D of finite length with endpoints z A , zB is the same for any Γ, no matter how Γ connects zA and zB . This is equivalent to the statement that I ( Γ ) = 0 for any closed contour Γ ⊂ D of finite length. It is clear from the previous consideration that I ( Γ ) in Eq. (1.8) is contour dependent only in the case when f (z ) has singular points in D . In the present book, we will consider only poles and branching points out of the whole variety of singular points. A point z0 ∈ D is a pole of the function f (z ) if and only if z0 is a zero of g (z ) = 1/f (z ), i.e., g (z0 ) = 0. The multiplicity n of the zero z0 of g (z ) is, at the same time, the multiplicity of the pole z0 of f (z ). It can be proved that the leading term in the Laurent series of the function f (z ) in a neighborhood of z0 is (z – z0 )–n , i.e., f (z ) =

∞ X

am (z – z0 )m ,

(1.9)

m=–n

where the coefficient a–1 is called the residue of the function f (z ) at the pole z 0 and denoted a–1 = Res [f (z ), z0 ]. If n = 1 in Eq. (1.9), then a–1 is the leading coefficient in the Laurent expansion, and such a pole is called a simple pole. There is quite a simple way to calculate the residue at a simple pole z0 : Res [f (z ), z0 ] =

h(z0 ) g 0 (z0 )

if

f (z ) =

h(z ) , g (z )

(1.10)

which is very efficient with any natural fractional decomposition (as in the case tan z = sin z/cos z ). Residues at poles play a key role in the calculation of integrals in the complex plane. This fact is represented by the Cauchy integral formula valid for any closed contour Γ ⊂ D traced counterclockwise Z X f (z ) dz = 2πi Res[f (z ), zm ], (1.11) Γ

m

where the residues are taken at all poles zm (of arbitrary multiplicity) inside Γ.


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This formula is very helpful for the calculation of integrals of the type (1.8). Quite often I ( Γ ) in (1.8) may be easily calculated for a certain simple path I ( Γ s ) passing through the endpoints zA , zB . Then the difference between I ( Γ ) and I ( Γs ), is equal to the sum of the residues at the poles between Γ and Γs , taken with appropriate sign. This strategy can be applied to integrals written along infinite lines also. In particular, let f (z ) in the Fourier transform (1.1) be analytic in a finite-width strip | Im(z )| ≤ δ . Then the integration contour Γ = (–∞, ∞) may be arbitrarily shifted up or down within this strip. Indeed, if the integral (1.1) is finite under the integration along the initial path (–∞, ∞), this implies that f (z ) → 0 as Re(z ) → ∞. Consequently, the integral of the same integrand (1.1) along any closed contour (–∞, ∞) ∪ (∞, ∞ + iε) ∪ (∞ + iε, –∞ + iε) ∪ (–∞ + iε, –∞), |ε| < δ , is zero. Since the two integrals over far finite vertical intervals vanish, because f (z ) decays in a far-zone, this proves our simple statement. In the case when there is a number of poles between the real axis and the line Im(z ) = ε parallel to it, it is evident that the same shift of the contour is possible if we add the residues at these poles. Sometimes this technique permits explicit calculation of Fourier transforms. In order to shift the integration contour Γ more up (or down), outside of a finite-width strip, we need to apply the following LEMMA (JORDAN). Let IR =

Z

f (z ) eisz dz ,

(1.12)

CR

where f (z ) is analytic everywhere in the upper half-plane Im(z ) ≥ 0, except perhaps a finite number of poles; Re(s ) > 0; f (z ) → 0 as z → ∞ uniformly over 0 ≤ arg(z ) ≤ π ; and C R is an upper semi-circle of radius R: |z | = R, Im(z ) ≥ 0. Then I R → 0 as R → ∞. The proof of this lemma is simple and can be found in the classical literature. Corollary. Under the same conditions, the Fourier transform F (s ) of (1.1) can be explicitly expressed as Z ∞ X F (s ) = f (x) eisx dx = 2πi Res [f (z ), zm ] eiszm . (1.13) –∞

Im(zm )>0

This result directly follows from the Cauchy integral formula (1.11) if you apply it to the function f (z ) exp(isz ) along the contour Γ = (–R, R) ∪ C R , with R → ∞.

LEMMA (GENERALIZED JORDAN LEMMA). Let f (z ) have a countable set of poles z m , m = 1, 2, . . ., Im(zm ) > 0, zm → ∞, m → ∞; and f (z ) vanishes uniformly on semicircles CRm of radius Rm as Rm → ∞; and each CRm passes somewhere between zm and zm+1 . Then for any s such that Re(s ) > 0, we have Z I Rm = f (z ) eisz dz → 0 as m → ∞. (1.14) C Rm

The proof of this less known result repeats the one for the classical Jordan lemma. Corollary. Under the same conditions, the Fourier transform can be explicitly calculated as an infinite series: Z ∞ ∞ X F (s ) = f (x) eisx dx = 2πi Res [f (z ), zm ] eiszm . (1.15) –∞

m=1

This corollary is very helpful when f (z ) is meromorphic, i.e., is the ratio of two entire functions: f (z ) = h(z )/g (z ). Recall that entire functions are defined as analytic over the


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whole complex plane, so countable sets of zeros and poles of any meromorphic function f (z ) are given by zeros of the entire functions h(z ) and g (z ), and both the resulting sets are finite if and only if f (z ) is rational. This is clearly demonstrated by the function f (z ) = tanh(z )/z , where h(z ) = sinh(z )/z and g (z ) = cosh(z ). It is also clear that in this example the set of upper semi-circles C Rm , m = 1, 2, . . . , can be determined from the condition tanh(iR m ) = 0 ∼ tan(Rm ) = 0 ∼ Rm = πm, which causes respective semi-circles to pass through the imaginary points iR m = πmi (see Fig. 1.1).

Im z CR m p p/2

-Rm

Re z Rm

Figure 1.1. Alternate poles and zeros of a meromorphic function Very often we will encounter below, in diffraction problems, some functions of a complex-valued argument that have branching points and hence are not single-valued. A typical representative here is the root square difference √ (1.16) γ (z ) = z 2 – k 2 , with a certain constant positive parameter k > 0 (see Mittra and Lee, 1971). Usually, in order to operate with a single-valued function, one has to arrange some cuts that become boundaries between different branches. For the root square difference (1.16) there are two branching points: z = k and z = –k , and it is quite natural to make such cuts that allow operating with the arithmetic value of the root square difference, i.e., the branch with Re(z ) ≥ 0. This can be provided by the cuts shown in Fig. 1.2, one of which passes totally in the upper half-plane √ half-plane Im(z ) < 0. Note that √ Im(z ) > 0 and the other in the lower for real z , γ (z ) = z 2 – k 2 ≥ 0 if |z | ≥ k , and γ (z ) = –i k 2 – z 2 if |z | ≤ k .

Im z

+k -k

Re z

Figure 1.2. Cuts in the complex plane z making the function γ (z ) =

√ z 2 – k 2 single-valued

For example, with such cuts the integral representation of the Hankel function (Abramowitz and Stegun, 1965) Z ∞ Z ∞ cos(xs ) exp(ixs ) ds = 2 ds = πiH0(1) (k |x|) (1.17) γ ( s ) γ ( s ) 0 –∞


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implies integration along the real axis, when the integration path lies between the two cuts. Note that the singularities s = ±k are integrable in the classical sense. It should be noted that branching functions, like the root square difference (1.16), generally are not analytic. However, some combinations of such functions can yield analytic and even entire functions, as can be seen by the example of the function sin[bγ (z )]/γ (z ) (b is constant). It is certainly an entire function, since it can be represented by a Taylor series, ∞ sin[bγ (z )] X (z 2 – k 2 )m b2m+1 = (–1)m , (1.18) γ (z ) (2 m + 1)! m=0

that is analytic and convergent for all finite z .

Helpful remarks 1◦ . Interestingly, quite often the “shortest way” between two real points “lies” in the complex plane. To illustrate this, let us consider the following integral over an interval of the real axis and with real integrands: Z ∞ cos(ax) dx (a ≥ 0, b > 0). (1.19) J1 = x2 + b 2 0 With the help of the Jordan lemma, we obtain iaz Z e e–ab π –ab 1 ∞ eiax dx = πi Res , ib = πi = e , (1.20) J1 = 2 2 2 2 2 –∞ x + b z +b 2bi 2b where the residue at the simple pole z = ib has been calculated with the method described above. 2◦ . The same approach is applicable to a meromorphic function if you use the generalized Jordan lemma (a, b > 0) : Z Z ∞ 1 ∞ iax sinh(bx) tanh(bx) cos(ax) dx = e dx J2 = x 2 –∞ x cosh(bx) 0 ∞ ∞ X X πi 1 e–πa(m–1/2)/b iaz sinh(bz ) , m– = πi Res e = πi (1.21) z cosh(bz ) b 2 π (m – 1/2) i m=1

m=1

∞ h πa i X 1 + e–πa/2b e–πa(m–1/2)/b = ln = ln coth , = m – 1/2 1 – e–πa/2b 4b m=1

where, in order to calculate the residues at simple poles, we have put f (z ) = h(z )/g (z ) with entire functions h(z ) = eiaz sinh(bz )/z and g (z ) = cosh(bz ). The following tabulated series has also been taken into account here (Gradshteyn and Ryzhik, 1994): ∞ X 1 1+x x2m+1 = ln , (|x| < 1). (1.22) 2m + 1 2 1 – x m=0

3◦ . Another remarkable phenomenon is related to the question why the Fourier transform, which (as a rule) converts real-valued functions to complex-valued, is so helpful when solving real boundary value problems. The answer can be seen from property 1 ◦ of the Fourier transform, since any derivative of an unknown function is converted to the same image with a factor containing Fourier parameter s . Therefore, if you solve any boundary value problem in a domain where the Cartesian coordinate x varies from –∞ to ∞, then the application of the Fourier transform will allow you to reduce the dimension of the problem by 1. This can reduce ordinary differential equations to algebraic ones, and a partial differential equation in two variables to an ordinary differential equation. Property 2◦ of the Fourier transform allows you to solve integral equations with convolution kernels explicitly. Both techniques will be demonstrated in detail below.


Page 5

1.2. Convolution Integral Equations and the Wiener–Hopf Method Generally, a convolution integral equation has the following form: Z b α ϕ(x) + K (x – ξ ) ϕ(ξ ) dξ = f (x), a < x < b,

(1.23)

a

which is evidently an equation of the second kind. In the case α = 0 it becomes an equation of the first kind. The function K (x) is a (known) kernel of the equation, and f (x) is a (known) right-hand side. The function ϕ(x) is unknown and is to be determined from Eq. (1.23). There is a special, unique case when equation (1.23) generally admits exact analytical solution. This is the case of b = ∞, where we get the Wiener–Hopf equation. In this case, a can be made equal to zero by a linear change of variable and only the first-kind equation (α = 0) will be important to us in this case. The solution of this equation is based upon some evident properties of the Fourier transform in the complex plane (see Bremermann, 1965; Mittra and Lee, 1971; Noble, 1958): 1◦ . If |f (x)| ≤ A eτ– x , x → +∞, then the function Z ∞ F+ (s ) = f (x) eisx dx

(1.24)

0

is analytic in the upper half-plane Im(s ) > τ– .

2◦ . If |f (x)| ≤ B eτ+ x , x → –∞, then the function Z 0 F– (s ) = f (x) eisx dx

(1.25)

–∞

is analytic in the lower half-plane Im(s ) < τ+ .

3◦ . If both properties 1◦ and 2◦ are satisfied and τ+ > τ– , then the full Fourier transform Z ∞ F (s ) = f (x) eisx dx (1.26) –∞

is analytic in the strip τ– < Im(s ) < τ+ , and the inverse Fourier transform may be calculated as follows: Z ∞+iτ 1 F (s ) e–isx ds , τ – < τ < τ+ . (1.27) f (x) = 2π –∞+iτ It is obvious from the previous section that you may arbitrarily deform the infinite integration contour Γ in (1.27) within the marked strip, if necessary. Now we are ready to apply the Wiener–Hopf method to the equation Z ∞ K (x – ξ ) ϕ(ξ ) dξ = f (x), 0 < x < ∞. (1.28) 0

Equation (1.28) is equivalent to Z ∞ K (x – ξ ) ϕ+ (ξ ) dξ = f+ (x) + f– (x), –∞

where

ϕ+ (x) =

ϕ(x), x > 0, 0, x < 0,

f+ (x) =

|x| < ∞, f (x), x > 0, 0, x < 0,

(1.29)

(1.30)


Page 6

and f– (x) is a certain additional unknown function. Equation (1.29) contains two unknown functions, ϕ+ (x) and f– (x); however both of them can be determined, as we will see soon, from a single equation. It should be noted that we assume the functions f + (x) and f– (x) to satisfy properties 1◦ and 2◦ , respectively, and the function K (x) the property 3 ◦ , with τ+ > τ– . Then, by applying the Fourier transform to Eq. (1.29) and using the convolution theorem (see the previous section), we arrive at the following relation: L(s ) Φ+ (s ) = F+ (s ) + F– (s ),

τ – < τ < τ+ ,

(1.31)

where L(s ) is the Fourier image of K (x). The key step of the method is the so-called factorization of L(s ), i.e., its representation in the form L(s ) = L+ (s ) L– (s ), where L+ (s ) is analytic with no singularity and no zero in the upper half-plane Im(s ) > τ– , and L– (s ) possesses similar properties for Im(s ) < τ+ . Then Eq. (1.31) becomes F+ (s ) F– (s ) + . (1.32) L+ (s ) Φ+ (s ) = L– (s ) L– (s ) The next step is decomposition, according to which the first fraction on the right-hand side of Eq. (1.28) is represented as F+ (s ) = N+ (s ) + N– (s ), L– (s )

(1.33)

which yields

F– (s ) , (1.34) L– (s ) where the function on the left is analytic for Im(s ) > τ – and that on the right is analytic for Im(s ) < τ+ . Since there is a common strip of analyticity, identity (1.34) represents a unique entire function, which coincides with the left-hand side in the upper half-plane and with the right-hand side in the lower half-plane. Typically, it can be proved that this entire function vanishes at infinity, so this implies that it is identically equal to zero. Hence, L+ (s ) Φ+ (s ) – N+ (s ) = N– (s ) +

L+ (s ) Φ+ (s ) – N+ (s ) = 0

∼

Φ+ (s ) =

N+ (s ) , L+ (s )

(1.35)

i.e., the Fourier image of the main unknown function ϕ(x) is defined explicitly, and so is its origin.

Helpful remarks 1◦ . In many cases factorization and decomposition can be performed in a simple natural way. For instance, some problems have kernels whose Fourier transform L(s ) is qualitatively like Pm (s ) 1 , (1.36) L(s ) = √ s 2 + d2 Qm (s ) where d is a positive parameter, and Pm (s ) and Qm (s ) are some polynomials of the same order. Such a function L(s ) admits very clear factorization: m + + m – = m,

L(s ) = L+ (s ) L– (s ),

1 L+ (s ) = √ d – is

ap aq

m Q+

k=1 m+ Q

(s – sp–k )

k=1

(s – sq–k )

,

m– Q

(s – sp+k ) 1 k=1 L– (s ) = √ , m– d + is Q (s – s + ) qk

(1.37)

k=1


Page 7

where ap and aq are the leading coefficients of the polynomials P m and Qm , m± is the number of zeros of Pm (s ) and Qm (s ) with positive (negative) imaginary part (see Fig. 1.3, where all zeros are indicated by crosses), and the cuts to arrange single-valued arithmetic value of branching functions are shown by zigzag lines. Irregular cases, where at least one of the zeros is real, will not be considered here. Let us denote τ– = max{–d; Im(sp–k ); Im(sq–k )},

τ+ = min{d; Im(sp+k ); Im(sq+k )}.

(1.38)

Then the function L+ (s ) is analytic and has no zeros in the upper half-plane Im(s ) > τ – , and L– (s ) is analytic and has no zeros in the lower half-plane Im(s ) < τ + .

Im z s+

sp+

qk

k

id

t+

G Re z t-

sp-

-id

k

sqk

Figure 1.3. The strip of common analyticity for the “+” and “–” functions √ Similar idea is applicable to decomposition also. For example, if L(s ) = 1/ s 2 + d2 , f (x) = e–βx (d, β > 0), then √ 1 1 F+ (s ) d + is 1 √ = L+ (s ) L– (s ), F+ (s ) = , = . (1.39) L(s ) = √ β – is L– (s ) (β – is ) d – is d + is The only singular point of this √ function in the lower half-plane Im(s ) < d is the simple pole s = –iβ , with the factor h = d + β . In order to get rid of this pole, you may subtract and add the quantity h/(β – is ): √ 1 p d+β F+ (s ) √ = + = N– (s ) + N+ (s ). (1.40) d + is – d + β L– (s ) β – is β – is

Here the first term is analytic in the lower half-plane Im(s ) < d (since it no longer contains the simple pole), and the second term is evidently analytic in the upper half-plane Im(s ) > –β . Once we have performed factorization and decomposition, the solution of the Wiener– Hopf equation is given by Eq. (1.35) √ √ d + β d – is N+ (s ) = , (1.41) Φ+ (s ) = L+ (s ) β – is whose inversion can be found in the tables of inverse Laplace transforms (Bateman and Erdelyi, 1954), since in our problem the Fourier transform is closely connected with the Laplace transform: Z ∞ Z ∞ Φ+ (s ) = ϕ(x) eisx dx = ϕ(x) e–px dx, p = –is , (1.42) 0

0


Page 8

so our Fourier image in terms of the Laplace argument p becomes √ √ d+β d+p , (1.43) Φ+ (p) = β+p whose Laplace original is –dx hp i p p e –βx + d – β e Erf (d – β )x , (1.44) ϕ(x) = d + β √ πx where Erf(x) is the probability integral (error function). 2◦ . If the kernel is too complex to be treated in such a straightforward way as in 1 ◦ , you may apply general formulas of factorization and decomposition, and a good survey of various representations of this kind can be found in (Mittra and Lee, 1971). But this general approach is hardly reasonable for practical purposes! General formulas are expressed in terms of some line integrals along infinite contours with too complex integrand depending on the variable s . Further, you have to substitute these numerically calculated functions into the infinite integral for the inverse Fourier transform to obtain the originals of your unknown quantities. Implementation of such integrals on a computer is a much more difficult task than your initial convolution integral equation. Instead, it is much more efficient to apply a numerical technique directly to the initial equation. Another possibility is to apply, prior to solving the equation, a uniform approximation to the image of the kernel by a certain function L(x) that admits a simple clear factorization based upon intuition.

1.3. Summation of Divergent Series and Integrals The concept of divergent series is quite natural in mathematical physics. It also arises in diffraction theory in many boundary value problems. Quite often (and we will see this soon from the forthcoming consideration) the structure of the solution near some boundary lines (like y = h in the example of a layer of constant thickness h) has qualitatively the following behavior: ∞ X am (y–h) α cos p(x, y ) = e m (bm x); sin (1.45) m=1

am ∼ am, bm ∼ bm, m → ∞, (a, b > 0), which in the case α ≥ 0 yields a series convergent inside the strip 0 < y < h only, and divergent on the boundary y = h. It is quite natural to treat the series (1.45) also on the boundary line y = h, implying that the limit of the sum (1.45) at y → h, if finite, may be admitted as its boundary value. This idea generates a vast theory of generalized summation (see, for example, Hardy, 1956; Rees et al., 1981). P∞ Poisson–Abel summation. If a series P∞ S = m m=1 am is divergent in the classical sense and there exists a limit of S (x) = m=1 am x as x → 1 – 0, then the (Poisson–Abel) generalized value of S is SPA = limx→1–0 S (x). It follows from the classical Cauchy–Hadamard formula for the convergence radius R–1 = sup |am |1/m of a power series (see Smirnov, 1964) in the case where |a m | grows not faster than any finite power mα , m → ∞, that R = 1. Hence there is a good chance that the generalized value SPA is finite. It is clear that the Poisson–Abel sum of a series can also be defined alternatively as ∞ ∞ X X SPA = lim am xm = lim e–εm am , (x = e–ε ), (1.46) x→1–0

m=1

ε→+0

m=1

which is in agreement with our heuristic idea discussed above.


Page 9

A similar technique can be applied to divergent integrals. If the integral Z ∞ f (x) dx S=

(1.47)

a

is divergent at infinity, then its Poisson–Abel generalized value (if exists) is defined as the limit Z ∞ e–εx f (x) dx.

SPA = lim

ε→+0

(1.48)

a

There are a number of alternative regularization methods for divergent series and integrals (for instance, the Cesaro method is very famous among others), but the Poisson– Abel summation is the most powerful and general. Examples. 1◦ . The series S1 =

∞ X

m cos(am)

(0 < a < 2π )

(1.49)

m=1

is divergent in the classical sense, since its terms increase without bound as m → ∞. The Poisson–Abel technique gives S1 = lim

x→1–0

∞ X

x→1–0

m=1

= Re lim

x→1–0

2◦ . The integral

xm m cos(am) = Re lim

∞ X

m xe

m=1

S2 =

Z

∞

ia m

√

∞ X

mxm eiam

m=1

1 eia =– . = Re 2 2 4 sin (a/2) (1 – eia )

x cos(ax) dx

(a > 0)

(1.50)

(1.51)

0

is divergent at infinity. Its Poisson–Abel generalized sum is Z ∞ Z ∞ √ –(ε+ia)x √ –εx x cos(ax) dx = Re lim xe dx S2 = lim e ε→+0 0 ε→+0 0 √ √ √ πi/2 –3/2 1 π π π = Re = Re (e ) = Re lim ε→+0 2(ε + ia)3/2 2 (ia)3/2 2a3/2 √ √ π π . = 3/2 cos – 43 π = – 2a (2a)3/2

(1.52)

A completely different approach to regularization of divergent series and integrals is based on the theory of generalized functions, or distributions (see Gel’fand and Shilov, 1964; Bremermann, 1965). In these monographs you can find a detailed comparative analysis of methods for generalized summation and regularization of generalized functions. In the ambit of what we are doing in the present section, the following technique gives an appropriate treatment for this sort of series. Let us consider a series ∞ X S (α) = am (α), (1.53) m=1

where all terms are analytic functions of a complex-valued parameter α in a domain D , α = Re(α) + i Im(α) ∈ D , and S (α) is analytic in the same domain D . Then a (regularized)


Page 10

value of the series for α = α0 ∈ / D is (if exists) an analytic continuation of the function S (α) from the domain D , where it is analytic, to the value α 0 . The same idea is applicable to divergent integrals. If f (x, α) is analytic with respect to a complex-valued parameter α ∈ D , and the integral, Z b f (x, α) dx, (1.54) S (α) = a

represents a function analytic in D (any limit in this integral may be equal to ±∞), then an analytic continuation from the domain D to a value α = α 0 ∈ / D is called a regularized value of the integral S (α0 ). Let us demonstrate these definitions by examples 1◦ and 2◦ considered above. 1◦ . Let us consider the function S1 (α) =

∞ X

mα cos(am),

0 < a < 2π ,

(1.55)

m=1

which is analytic in the infinite domain D = {Re(α) < 0}. In this domain the series has an ordinary sum (Gradshteyn and Ryzhik, 1994) h a a i (2π )–α ζ 1 + α, + ζ 1 + α, 1 – , (1.56) S1 (α) = 4 Γ(–α) cos(πα/2) 2π 2π P∞ where ζ (s , v ) = k=0 1/(k + v )s is a generalized Riemann zeta function. As follows from Eq. (1.56), the function S1 (α) can be continued analytically up to the value α = 1. Here we should take into account the limits (Bateman and Erdelyi, 1953) α–1 1 2 1 ∼ ∼ – →– , Γ(–α) cos(πα/2) cos(πα/2) (π/2) sin(πα/2) π

α → 1,

(1.57)

and the limit value ζ (2, z ) + ζ (2, 1 – z ) =

∞ X k=0 0

∞

X 1 1 + 2 (k + z ) (k + 1 – z )2 k=0

= ψ (z ) + ψ 0 (1 – z ) = [ψ (z ) – ψ (1 – z )]0 π2 , = –π [cot(πz )]0 = sin2 (πz )

(1.58)

where ψ (z ) is the logarithmic derivative of Euler’s Gamma function Γ(z ). Collecting formulas (1.55)–(1.58) together, we arrive at the same result (1.50): S 1 (α = 1) = –1/[4 sin2 (a/2)]. For other values of the parameter a, outside the interval (0, 2π ), this identity can be continued by periodicity. 2◦ . The function S2 (α) =

Z

∞

xα cos(ax) dx

(a > 0)

(1.59)

0

is analytic in the domain D = {–1 < Re(α) < 0}, where it is equal to (see Gradshteyn and Ryzhik, 1994) πα Γ(α + 1) sin . (1.60) S2 (α) = – α+1 a 2


Page 11

The analytic continuation of S2 (α) to the point α = 1/2 is S2 (1/2) = –

√ π π Γ(3/2) sin = – , a3/2 4 (2a)3/2

(1.61)

which coincides with (1.52). Combination of the Fourier transform with the concept of generalized functions permits the application of the Fourier transform to those functions for which it does not exist in the ordinary sense. As a rule, this applies to functions that do not vanish at infinity, and so are not integrable. This automatically implies that the classical Fourier transform for such functions is not finite. A detailed theory of the Fourier transform for generalized functions can be found, for example, in Bremermann (1965) and Zemanian (1969). Here we only cite some rather classical results of this theory, which may be helpful, in one or another way, for further theories. The most important applications are related to the properties of Dirac’s delta function. First of all, the Fourier transform of the unit step function is Dirac’s delta and some helpful properties of this function and its relations to other known functions are listed below: d |x| = sign(x), 1(x) =⇒ 2πδ (s ), dx Z ∞ (1.62) d sign(x) = 2 δ (x), f (ξ ) δ (x – ξ ) dξ = f (x). dx –∞

Helpful remarks If you are an applied mathematician and trust more in concrete results when implemented on your home PC, rather than in abstract theory of generalized functions, then you may want to arrange computations to test correctness of these abstract theories. You may write a code, using an algorithmic language (such as Fortran, C, Pascal, etc.), for computing some formulas like those given by Eq. (1.50). P∞ Then, for example, in the case a = π you will obtain for the sum of the series S1 (x) = m=1 (–1)m mxm numbers for various x similar to those obtained by us (the upper limit in the infinite series was taken 10 4 and double-precision Fortran computations were carried out): 0.89 0.91 0.93 0.95 0.97 0.99 –0.24915 –0.24944 –0.24967 –0.24983 –0.24994 –0.24999 P∞ As x → 1 – 0, the series approaches the value S1 (1) = m=1 (–1)m m = –1/[4 sin2 (π/2)] = –0.25. You may also test the accuracy of formulas (1.52), (1.61). Results of computations for a = 1/2 and various ε are shown here: x S1 (x)

ε S2

0.011 0.009 0.007 0.005 0.003 0.001 –1.710 –1.722 –1.738 –1.742 –1.751 –1.776

These results were calculated with a double-precision Fortran code√as an integral with the upper limit 104 . These values should be compared with S2 (1/2) = π = –1.772.

1.4. Asymptotic Estimates of Integrals It is a rather typical situation that you cannot construct any explicit analytic solution for your complex problem of mathematical physics, but you succeed in finding an approximate


Page 12

solution in some specific range of a certain physical parameter (usually, small or large). The idea of asymptotic estimates is very fruitful in such cases. We will say that an infinite set of functions {gn (λ)}∞ n=0 forms an asymptotic scale as λ → λ0 if gn+1 (λ) = o(gn (λ)), λ → λ0 . It can easily be proved that if {gn (λ)} is an asymptotic scale, then the functions {gnδ (λ)} also form a scale for any δ > 0. In practice we will use only λ0 = 0 or λ0 = ∞, where the natural scales are formed by powers and logarithmic functions. In particular, the {λ–µn }, where µn → +∞ as n → ∞, form a scale as λ → +∞, and so do the functions {(ln λ)–µn }. Certain combinations of power and logarithmic functions of the type {λ–δn (ln λ)–µn } can also form asymptotic scales as λ → +∞. By analogy, both sets {λµn } and {[ln(1/λ)]–µn }, as well as some of their combinations, form asymptotic scales as λ → +0. P∞If {gn (λ)} forms an asymptotic scale as λ → λ0 , then the formal expansion f (λ) ∼ n=0 an gn (λ) is called a full asymptotic expansion of the function f (x) at λ → x 0 if for any N = 0, 1, 2, . . . f (λ) =

N X

an gn (λ) + o(gN (λ)),

n=0

λ → λ0 .

(1.63)

In the case where relation (1.63) is valid only for some N , it is an asymptotic expansion of order N . As a rule, in actual complex problems only the leading asymptotic term (N = 0) can be constructed in a direct elegant way. Further terms usually require too cumbersome transformations. The most widespread type of integrals admitting natural asymptotic estimates contains a large asymptotic parameter in the argument of exponential function. In the case where the phase function S (x) is real-valued, we arrive at the Laplace integral (Erdelyi, 1956) Z b I (λ) = f (x) eλS(x) dx, (a < b ≤ ∞). (1.64) a

If the density f (x) and the phase S (x) are smooth—more precisely if f (x), S (x) ∈ C ∞ [a, b]—and maxx∈[a,b] S (x) = S (x0 ), where x0 is an internal point of the interval [a, b], then the full power asymptotic expansion for I (λ) is expressed as follows (Erdelyi, 1956; Fedorjuk, 1977): Z b ∞ X λS(x) λS(x0 ) I (λ) = f (x) e dx ∼ e cn λ–n–1/2 , λ → +∞, where a

Γ(n + 1/2) lim cn = (2n – 1)! x→x0

d dx

n (

n=0

f (x)

S (x0 ) – S (x) (x – x0 )2

–n–1/2 )

(1.65)

.

Heuristically, the principal asymptotic contribution to the integral I (λ) is given by a small neighborhood of the point x0 (where S 0 (x0 ) = 0), and the relative contribution of all other points is exponentially small. It is obvious from Eq. (1.65) that the leading asymptotic term is r 2π f (x0 ) eλS(x0 ) + O λ–3/2 , λ → +∞, (1.66) – 00 I (λ) ∼ λS (x0 ) where, of course, we have S 00 (x0 ) < 0 at the point x0 of (local) maximum. The estimate (1.66) is also valid for f (x) ∈ C1 [a, b].


Page 13

In the case where the point x0 of global maximum coincides with a boundary point (x0 = a or x0 = b), the considered integral (1.64) usually looks like a Laplace-transform integral: Z a

I (λ) =

xβ–1 f (x) exp(–λ xα ) dx

0

(α, β > 0, 0 < a ≤ ∞).

(1.67)

If f (x) ∈ C ∞ [0, a] here, then the asymptotic expansion is given by the following Watson lemma (see Erdelyi, 1956; Fedorjuk, 1977). LEMMA (WATSON). Z

a

α

xβ–1 f (x) e–λ x dx 0 ∞ n + β f (n) (0) –(n+β)/α 1X λ , Γ ∼ α n=0 α n!

I (λ) =

(1.68) λ → +∞.

Estimates like (1.65) and (1.68) can be derived, after some appropriate preliminary treatment, either with the help of integration by parts or using Taylor series expansion for the density f (x). For a general analytic phase function the Laplace transform (1.67) admits the same integration by parts, which yields the following asymptotic expansion: Z a ∞ X cn , λ → +∞ (0 < a ≤ ∞), I (λ) = f (x) e–λ S(x) dx ∼ e–λ S(0) λn+1 0 n=0 (1.69) n d f (x) 1 , f (x), S (x) ∈ C ∞ [0, a], cn = S 0 (x) dx S 0 (x) x=0 if x0 = 0 is a point of global maximum of the phase S (x).

Example. This example demonstrates an asymptotic expansion with logarithmic scale. The integral Z ∞ dx (1.70) J (λ) = e–λx x ln2 x 0 is finite in the classical sense, since the singularity at the origin (x → 0) is integrable, which can be proved by integration by parts: ∞ Z ∞ –λx Z ∞ –λx e e –λx 1 –λ dx = –λ dx . (1.71) J (λ) = –e ln x x=0 ln x ln x 0 0

If we are interested in an asymptotic estimate of J (λ) for large λ, then we can use a Taylor expansion: Z ∞ Z ∞ Z ∞ –x e–x dx 1 e–x dx e dx = = J (λ) = – ln(x/λ) ln λ – ln x ln λ 0 1 – (ln x/ln λ) 0 0 Z (1.72) ∞ ∞ ∞ X X Γ(n) (1) 1 n –x e ln ( x ) dx = , ∼ (ln λ)n+1 0 (ln λ)n+1 n=0 n=0 where the value of the last integral has been taken from the tables (see Gradshteyn and Ryzhik, 1994).


Page 14

More often, in diffraction theory, the exponential kernel in integrals is not a positive (as in Eqs. (1.64), (1.67)) but an oscillating function. The following lemma is analogous to the Watson lemma for integrals with positive kernels (Erdelyi, 1956; Fedorjuk, 1977). LEMMA (ERDELYI). Let a ≤ ∞; α, β > 0; f (x) ∈ C ∞ [0, a]; and f (n) (a) = 0 (∀n = 0, 1, 2, . . . ). Then Z a ∞ n + β f (n) (0) 1X β–1 iλxα Γ (–iλ)–(n+β)/α , λ → +∞. (1.73) dx ∼ x f (x) e α α n ! 0 n=0 The more general Fourier integral Z b I (λ) = f (x) eiλS(x) dx,

S (x) ≠ const,

(1.74)

a

where the phase function S (x) is real-valued, can also be estimated. As the parameter λ increases the exponential function in (1.74) becomes rapidly oscillating; however the integral itself is asymptotically small, since the contribution of its positive and negative parts almost cancel each other out. The only unclear question is how small it is. If f (x), S (x) ∈ C ∞ [a, b] and S 0 (x) ≠ 0, x ∈ [a, b] (a < b < ∞), then Z b f (x) eiλS(x) dx I (λ) = a

∞ X en dn iλS(b) – e , λ → +∞, n+1 (iλ) (iλ)n+1 n=0 n=0 n n d f (x) 1 d f (x) 1 , en = . dn = S 0 (x) dx S 0 (x) x=a S 0 (x) dx S 0 (x) x=b

∼ eiλS(a)

∞ X

(1.75)

It follows from the last expansion that, with exponentially small error, the principal contribution to the asymptotics of I (λ), λ → +∞ is made by small neighborhoods of the boundary points a and b, and the leading term has the order of O (1/λ), λ → +∞. If f (x) ∈ C 1 [a, b] only, then the estimate (1.75) remains valid, containing only the leading contributions from the ends. In the case where the density f (x) vanishes at the ends of the interval [a, b] together with all its derivatives, f (n) (a) = f (n) (b) = 0, ∀n, the value of I (λ) is exponentially small. It should be noted that the formal structures of Eqs. (1.69) and (1.75) are quite similar. This is because both expansions have been derived with the help of integration by parts. However, the intrinsic essence of these expansions is totally different. The former is a consequence of the fact that the integrand decays exponentially far away from the endpoint x0 = 0. By contrast, the latter is determined by mutual cancellation of the positive and negative contributions, far away from the endpoints. DEFINITION. A point x0 ∈ (a, b) where S 0 (x0 ) = 0 is called a stationary point. Intuitively, since the stationary point is at the same time a point of local extremum, the phase function S (x) is almost constant in a small neighborhood of any stationary point. We thus expect that its contribution to the integral I (λ) of (1.75) is asymptotically more significant than those from the endpoints. A precise result here is given by the following statement, which represents a stationary phase method. Let f (x), S (x) ∈ C ∞ [a, b] and let there exist only one internal stationary point x 0 : a < x0 < b, S 0 (x0 ) = 0, S 00 (x0 ) ≠ 0. Then if f (n) (a) = f (n) (b) = 0, ∀n = 0, 1, 2, . . . , we have Z b ∞ X an iλS(x) , λ → +∞, (1.76) I (λ) = f (x) e dx ∼ n+1/2 λ a n=0


Page 15

where the leading asymptotic term is s 1 2π i[λS(x0 )+πδ/4] e f (x0 ) + O , λ → +∞, δ = sign[S 00 (x0 )]. I (λ) ∼ 00 λ| S (x0 ) | λ (1.77) We can see that the contribution of the stationary point is of the order of O (λ –1/2 ) (compare with the contribution given by an endpoint, which is O (1/λ)). Note that if there are several stationary points on the interval x ∈ (a, b), then the complete asymptotics is given by adding all the contributions together. This stationary phase method can be extended to multiple integrals (Fedorjuk, 1962; 1977). Let Ω ⊂ RN be a bounded domain and let the density f (x) ∈ C ∞ (Ω) be compactly supported (i.e., it vanishes when approaching the boundary of Ω, together with all its partial derivatives). Let the real-valued phase S (x), x ∈ Ω, have a unique stationary point x 0 ∈ Ω, such that the gradient S 0 (x0 ) = 0 and det S 00 (x0 ) ≠ 0, where S 00 (x) is the Hessian of the function S (x), i.e., the matrix of its second-order partial derivatives ∂ 2 S/∂xn ∂xk . Then the multiple stationary-phase method determines the contribution of the stationary point x 0 as follows: Z ∞ X an , λ → +∞, (1.78) I (λ) = f (x)eiλS(x) dx ∼ eiλS(x0 ) n+N/2 λ Ω n=0

with the leading asymptotic term being N/2 f (x0 ) 2π , ei[λS(x0 )+πiδ/4] √ I (λ) ∼ λ | det S 00 (x0 )|

λ → +∞,

(1.79)

where δ = ν+ [S 00 (x0 )] – ν– [S 00 (x0 )] is the difference between the number of positive and negative eigenvalues of the Hessian S 00 (x) at the stationary point x0 . Sometimes it is very important to know the behavior of the Laplace integral (1.67) not only for large but also for small λ. Unexpectedly, this turns out to be a harder task. An important result is stated as follows (Handelsman and Lew, 1970; Riekstinsh, 1977). If ∞ X cn , x → +∞, n < µn < n + 1 (c0 ≠ 0), (1.80) f (x) ∼ x µn n=0 then

I (λ) =

Z

0

∞

e–λx f (x) dx ∼

where

∞ X

cn Γ(1 – µn ) λµn –1 +

n=0

dn =

Z

0

∞

∞ X (–1)n dn λn , n ! n=0

λ → +0, (1.81)

n X cj dx. xn f (x) – µj x j=0

Note that the µn in Eqs. (1.80), (1.81) are noninteger. Otherwise, in the case of integer µ n we have: if ∞ X cn , x → +∞ (c1 ≠ 0), (1.82) f (x) ∼ xn n=1 then

I (λ) ∼ ln λ

∞ ∞ X X (–1)n (–1)n cn+1 λn + en λn , n ! n ! n=0 n=0

λ → +0,

(1.83)


Page 16

where en =

Z

1

x 0

n

f (x) –

n X

cj x

n–j

j=1

dx +

Z

n

x

1

n

n n+1 X X 1 cj 0 dx + cn+1 Γ (1) + . f (x) – xj j j=1 j=1

These results can be extended to Fourier-type integrals. If the behavior of the function f (x) at infinity is given by Eq. (1.80), then Z ∞ I (λ) = eiλx f (x) dx 0

∼

∞ X

cn Γ(1 + n – µn )(–1)n (–iλ)µn –1 +

n=0

∞ X dn (iλ)n , n ! n=0

(1.84)

λ → +0,

where dn is the same as in Eq. (1.81). In the case where the µn are positive integers, i.e., if f (x) possesses for large x asymptotics of the type (1.82), we have (Grosjean, 1965; Riekstinsh, 1981) I (λ) ∼ – ln λ

∞ ∞ X X hn cn+1 (iλ)n + (iλ)n , n ! n ! n=0 n=0

λ → +0,

h n = en +

πi , 2n!

(1.85)

where en is given by Eq. (1.83).

Helpful remarks 1◦ . If you compare the properties of asymptotic series with those of usual convergent series, you may find that these properties, according to their definitions, are absolutely different. If we treat the expansion (1.63) as an infinite series (with N = ∞), then the convergence of an ordinary series implies that the difference between the left and the right hand sides in (1.63) vanishes as N → ∞. By contrast, the expansion (1.63) treated in the asymptotic sense means that this discrepancy vanishes as λ → λ 0 . We can conclude from these arguments that typically convergent series do not represent any asymptotic series, and vice versa, many asymptotic series, which are very helpful in efficient calculations with λ around λ0 , do not converge in the classical sense. 2◦ . If you compare the asymptotic behavior of the Laplace transform Z ∞ F (λ) = e–λx f (x) dx

(1.86)

0

for large (Eq. (1.68)) and small (Eq. (1.81)) values of the parameter λ in the case when the principal term of the density is a noninteger power, you will be surprised that the leading asymptotic terms coincide in the sense that (A = const) +0 µ if f (x) ∼ A x (–1 < µ < 0), x → , +∞ (1.87) +0 –µ–1 then F (λ) ∼ A Γ(1 + µ) λ , λ → . +∞ Remarkably, the respective leading terms of the Fourier integrals have the same form (see Eqs. (1.73), (1.83) with n = 0) if you formally substitute –iλ for λ in the Laplace integral.


Page 17

3◦ . We considered above only real-valued and imaginary phases. In some problems the phase S (x) appears complex-valued, which leads to the so-called steepest descent method, described in detail in the literature (e.g., see Fedorjuk, 1962). This method is thus designed to deal with more general cases than Laplace or oscillating integrals. Amazingly, after some refined manipulation, almost always such problems can be reduced to a combination of Laplace-type and Fourier-type integrals. For example, the integral Z ∞ √ (k > 0), (1.88) J= F (α) e–iαx e–γ(α) y dα, γ (s ) = α2 – k 2 –∞

where y > 0 and we assume the function √ F (s ) to be even for simplicity (so J is even with respect to x), is usually studied for r = x2 + y 2 → ∞ by the steepest descent method (Mittra and Lee, 1971) with the help of a certain transformation of the integration contour in √ the complex plane (recall the analytic properties of the branching function γ (α) = α2 – k 2 , Section 1.1). Let us show that the integral (1.88) can be correctly estimated just over the real ray Im(α) = 0, Re(α) ≥ 0: Z ∞ J= F (α) eiα|x| + e–iα|x| e–γ(α) y dα 0 Z ∞ Z k √ 2 2 i√k2 –α2 y (1.89) iα|x| –iα|x| dα + F (α) eiα|x| + e–iα|x| e– α –k y dα F (α) e +e = e 0

k

= J1 + J2 .

The second integral in (1.89) admits estimation of the leading term from (1.67), where α = 1/2, β = 1, λ = y , n = 0: Z ∞ √ √ |J2 | = F (α + k ) ei(α+k)|x| + e–i(α+k)|x| e– α+2k y e– α y dα 0 (1.90) Z ∞ √ 1 – αy , y → +∞. dα ∼ O ≤ |F (α + k )| e y2 0 This relation holds uniformly over |x| < ∞. For any fixed√y , J 2 can be estimated for x → ∞ by (1.75) as O (1/|x|). Therefore, J2 ∼ O (1/r ) as r = x2 + y 2 → ∞. The first integral J1 can be treated by the stationary phase method. Let |x| = r sin ϕ, y = r cos ϕ, r → ∞. Then in the integral J1 , Z k i h √ √ 2 2 2 2 (1.91) J1 = F (α) eir( k –α cos ϕ–α sin ϕ) + eir( k –α cos ϕ+α sin ϕ) dα, √

0

the phase S (α) = k 2 – α2 cos ϕ – α sin ϕ of the first term has no stationary points (S 0 (α) ≠ 0, 0 ≤ α ≤√k ), so its behavior is again of the order of O (1/r ). The other term has the phase S (α) = k 2 – α2 cos ϕ + α sin ϕ with the stationary point α0 = k sin ϕ, S 0 (α0 ) = 0. Taking into account that S 00 (α0 ) = –1/(k cos2 ϕ), δ = –1, S (α0 ) = k , we thus arrive at the final result, in accordance with Eq. (1.77): r 1 2πk i(kr–π/4) cos ϕ e f (k sin ϕ) + O , r → ∞. (1.92) J ∼ r r 4◦ . In practice, you may always try to extract the leading asymptotic term by letting λ tend to infinity. Such an approach can be, as a rule, proved strictly as a theorem. Thus, the following integral: Z ∞ Z ∞ J (t) dt J0 (λx) dx √ √ 20 = , (1.93) 2 2 t + (aλ)2 x +a 0 0


Page 18

after the change of variable λx = t, can be estimated by letting λ ∼ +∞, which implies the first term under the square root to be negligibly small compared with (aλ). This yields the leading asymptotic term in the form Z ∞ Z ∞ Z ∞ 1 1 J0 (λx) dx √ ∼ J0 (t) dt = , λ → +∞ J0 (t) dt = 1 , (1.94) aλ 0 aλ x 2 + a2 0 0 where we have used the value of a tabulated integral (Gradshteyn and Ryzhik, 1994). In the cases where you do not succeed in applying such a simple and clear procedure, it is quite probable that your asymptotic expansion involves logarithmic or fractional power functions rather than integer powers. For example, the integral Z ∞ Z ∞ 2 J 2 (t) dt J0 (λx) dx √ √ 20 , (1.95) = t + (aλ)2 x 2 + a2 0 0

which isRallied to (1.93), cannot be estimated for large λ in the same way because the ∞ integral 0 J02 (t) dt diverges at infinity. This is a sure sign for a logarithmic scale to appear. The same feature becomes apparent for both integrals (1.93), (1.95) as λ → 0. Indeed, with λ → +0, all arising integrals, Z ∞ 2 Z ∞ Z ∞ J0 (t) dt J0 (t) dt dx √ , , (1.96) (J0 (0) = 1), 2 2 t t x +a 0 0 0

diverge, the first one at infinity and the others at the origin.

5◦ . The phenomenon indicated in 4◦ is related to a very difficult case where the integrand does not contain any logarithmic function but, at the same time, the asymptotic expansion yields a logarithmic term in the scale. In this case it is advised that you reduce your integral to any combination of standard special functions, which have been well studied over the last 100 years and whose asymptotics are described in detail. For example, the integral (1.93) is Z ∞ aλ aλ J0 (λx) dx √ = I0 K0 , (1.97) 2 2 2 2 x +a 0 where I0 is the modified Bessel function of order 0 and K0 is the McDonald function of the same order. First of all, this leads to a full asymptotics as λ → +∞ (Abramowitz and Stegun, 1965): Z ∞ ∞ n 1 X (2n – 1)!! Y J0 (λx) dx √ ∼ + (–1)n (2j – 1)2 , λ → +∞, (1.98) 2n+1 2 2 aλ (2 n )!! ( aλ ) x +a 0 n=1 j=1 which, in particular, gives our leading term (1.94). Representation (1.97) permits also a small-λ expansion, which contains logarithmic functions. It can be obtained from the series: n ∞ ∞ X X (z/2)2n X 1 (z/2)2n , K0 (z ) = –[ln(z/2) + γ ] I0 (z ) + , (1.99) I0 (z ) = (n!)2 (n!)2 j n=1 j=1 n=0 where γ = 0.577216 is the Euler constant. In particular, the first two leading terms are: Z ∞ J0 (λx) dx √ ∼ – ln λ + [ln(4/a) – γ ] + O (λ2 ln λ), λ → +0. (1.100) 2 2 x +a 0


Page 19

The same idea can be applied to the integral (1.95). To this end, we use the integral representation (Abramowitz and Stegun, 1965) Z 2 π/2 2 J0 (2t cos ψ ) dψ , (1.101) J0 (t) = π 0

which, with the help of (1.97), gives Z Z ∞ 2 2 π/2 J0 (λx) dx √ = I0 (aλ cos ψ )K0 (aλ cos ψ ) dψ , π 0 x2 + a 2 0

and, by using series (1.99), for small λ we finally obtain Z ∞ 2 J0 (λx) dx √ ∼ – ln λ + [ln(4/a) – 2γ ] + O (λ2 ln λ), 2 2 x +a 0 Z π/2 since ln(cos ψ ) dψ = π (γ – ln 2)/2

λ → +0,

(1.102)

(1.103)

0

(see Gradshteyn and Ryzhik, 1994). The most difficult problem is to construct the asymptotics of the integral (1.80) as λ → +∞, which requires a more refined approach. With the help of tabulated integrals (Gradshteyn and Ryzhik, 1994) one can observe that Z ∞ Z ∞ Z ∞ 2 2 J0 (λx) dx 2 √ = J0 (x) dx cos(ξx) K0 (|α|ξ ) dξ π 0 x 2 + a2 0 0 ) ( Z ∞ 0, ξ > 2 2 dξ K0 (|α|ξ ) 1 = (1.104) 1 2 π 0 – 1 , 0 < ξ < 2 P ξ –1/2 2 2 2 Z 2|α| ξ 1 K0 (ξ ) P–1/2 – 1 dξ , = π |α| 0 2α2

where Pν (x) is the Legendre function, whose asymptotic behavior near the singular point x = –1 is (Bateman and Erdelyi, 1953) 1+x sin(πν ) + γ + 2ψ (ν + 1) + π cot(πν ) + o(1). (1.105) ln Pν (x) ∼ π 2

Now asymptotic results of the theorem directly follow from the exponential decay of the McDonald function at infinity, which yields the first two leading asymptotic terms in the following form: Z ∞ Z ∞ 2 1 ξ2 1 J (x) dx √0 ∼– 2 K0 (ξ ) ln 2 + γ + 2ψ dξ 2 2 π | α | 4 α 2 α +x 0 0 1 γ 1 (1.106) ln |α| – ψ + 2 ln 2 + = π |α| 2 2 3 1 ln |α| + 4 ln 2 + γ , α → ∞, = π |α| 2

since ψ (1/2) = –(γ + 2 ln 2). Here the following tabulated integrals have been taken into account (Gradshteyn and Ryzhik, 1994): Z ∞ Z ∞ π π K0 (ξ ) ln ξ dξ = – (ln 2 + γ ). (1.107) K0 (ξ ) dξ = , 2 2 0 0


Page 20

1.5. Fredholm Theory for Integral Equations of the Second Kind The theory of integral equations of the second kind Z u(x) – µ G(x, ξ ) u(ξ ) ds (ξ ) = f (x), S

x ∈ S,

(1.108)

was devised by Fredholm and further developed by Riesz (see, for example, Riesz and Sz.-Nagy, 1972). Here the right-hand side f (x) and the kernel G(x, ξ ) are some known functions, µ is a known parameter, S is a piecewise-smooth surface (or line), and u(x) is an unknown function to be determined from Eq. (1.108). Note that, in order to specify the variable (x or ξ ) with respect to which the integral operator is applied, we explicitly indicate that the element ds depends on the variable ξ . The following results of the Fredholm theory can be regarded as classical. Let us rewrite Eq. (1.108) as a linear functional equation Z (I – µG) u = f , (Gu)(x) = G(x, ξ ) u(ξ ) ds (ξ ). (1.109) S

We will consider it in the normal space of continuous functions X = C (S ) with the norm ||ϕ||C(S) = maxx∈S |ϕ(x)|. The set U ⊂ X is called compact if an arbitrary sequence {u n} ⊂ U contains a convergent subsequence.

THEOREM 1 (ARZELA–ASCOLI). A set U ⊂ X of functions u(x), x ∈ S is compact in C (S ) if and only if: 1) there exists a constant A such that |u(x)| ≤ A for all x ∈ S and all u(x) ∈ U ; 2) for any ε > 0 there exists a constant δ > 0 such that for all x, y ∈ S : |x – y | < δ and all u(x) ∈ U the inequality |u(x) – u(y )| < ε holds.

The first property means that the set U is uniformly bounded in C (S ), and the second one means that the functions from U are equicontinuous. An operator G : X → X is called compact if it maps any bounded set to a compact one. It is well known that any linear compact operator is continuous (i.e., bounded). Below in this section we will consider only compact operators G. A simple theorem states that if the kernel of Eq. (1.108) is continuous with respect to both its variables, G(x, ξ ) ∈ C (S × S ), then the operator G in (1.109) is compact. The proof directly follows from the Arzela–Ascoli theorem. The same result holds if the operator G is weakly singular, i.e., if the kernel G(x, ξ ) is continuous for x ≠ ξ and G(x, ξ ) ≤ B |x – ξ |α–n (0 < α < 1) (n is the dimension of S , i.e., n = 2 when S is a surface, and n = 1 when S is a line). The kernel G(x, ξ ) with weak singularity is integrable over S . Note that the logarithmic singularity, when G(x, ξ ) ∼ A ln |x – ξ |, ξ → x, is also weak. It should also be noted that the same properties of compactness, as in C (S ), take place in the Hilbert space L2 (S ), too. Two operators A, B : X → X are called conjugate to each other if for all u, v we have (Au, v ) = (u, Bv ). It is evident that if G(x, ξ ) in (2) is continuous or ¯ (ξ , x) is conjugate weakly singular, then the integral operator G∗ with kernel G∗ (x, ξ ) = G to the operator G. Thus, any symmetric real-valued kernel yields some self-conjugate (self-adjoint) operator. THEOREM 2 (FIRST FREDHOLM THEOREM). The subspaces of solutions to the homogeneous equations (I – µG) u = 0, (I – µG∗ ) v = 0, (1.110) have equal finite dimension n < ∞ in X . In particular, these subspaces may be empty (i.e., n = 0).


Page 21

THEOREM 3 (SECOND FREDHOLM THEOREM). Nonhomogeneous equations (I – µG) u = f ,

(I – µG∗ ) v = g ,

(1.111)

have solutions if and only if (f , vj ) = 0, (uj , g ) = 0 for all solutions of the conjugate homogeneous equations (I – µG) uj = 0,

(I – µG∗ ) vj = 0 .

(1.112)

From these two theorems you can directly come to the conclusion stated by the following THEOREM 4 (FREDHOLM’S ALTERNATIVE, OR THIRD FREDHOLM THEOREM). If the homogeneous equations (1.110) have (simultaneously) only the trivial solution u = 0, v = 0, then the nonhomogeneous equations (1.111) have unique solutions for all f , g ∈ C (S ). In this case there exist continuous inverse operators (I – µG) –1 , (I – µG∗ )–1 . In the case where equations (1.110) have (simultaneously) nontrivial solutions, the nonhomogeneous equations (1.111) have a solution if and only if relations (1.112) hold. DEFINITION. The set of irregular values of the parameter λ related to the second case in Fredholm’s alternative, when Guj = λuj , uj ≠ 0, is called the spectrum of the operator G, or the set of its eigenvalues . The corresponding values µ = 1/λ are called characteristic values of G. THEOREM 5 (FOURTH FREDHOLM THEOREM). The set of characteristic values of integral operators with continuous or weakly singular kernels consists of a (possibly complexvalued) countable discrete array µn , n = 1, 2, . . . , which has no finite limit point. The theory briefly outlined above provides us with a powerful method for solving integral equations of the second kind. First of all, it is clear that for sufficiently small µ the first case of two different possibilities in Fredholm’s alternative always holds, and so equations (1.111) are uniquely solvable for all f , g ∈ C (S ). To show this, let us start with proof by contradiction. If there is a nontrivial solution u j to Eq. (1.110)1 , then uj = µ Guj , which implies ||uj || = |µ| ||Guj || ≤ |µ| ||G|| ||uj || ∼ |µ| ≥ 1/||G||. The last inequality is wrong, since for a compact operator, we have G : ||G|| ≤ A, and hence ||µ|| ≥ 1/A, which contradicts the assumption that µ is small enough. The last conclusion in particular means that for small µ the inverse operators (I – µG) –1 and (I – µG∗ )–1 exist and, besides, are continuous (bounded). Moreover, P∞ they may be explicitly calculated as Neumann series in iterated kernels: (I – µG) –1 = j=0 µj Gj (where Gj = G . . . G with the j -times repeated product of G), and the parameter µ is small enough. Another important practical conclusion from Fredholm’s alternative is that it predetermines, in the case of invertibility, continuous dependence on the right-hand sides f , g , since in this case u = (I – µG)–1 f , v = (I – µG∗ )–1 g with both continuous operators, which guarantees stability of a solution with respect to small perturbations in f and g . This property is very important in the context of what is discussed in Chapter 8, devoted to the theory of ill-posed problems, which covers also operator equations of the first kind. A more precise estimate for a maximum value of the parameter µ, for which the operator I – µG is surely invertible, follows from the well-known theorem: if the norm of the operator K : X → X is ||K || < 1, then the operator (I – K ) is invertible. From this statement we can conclude that Eq. (1.108) or (1.109) is surely solvable and there exists a continuous inverse operator (I – µG)–1 if |µ| < 1/||G||. Of course, this does not mean that for large µ Eq. (1.108) is unsolvable. The Fredholm theory for equations of the second kind gives you also a powerful practical tool for solving these equations numerically. Let us assume that we replace the operator G by some approximate operator G1 that admits a simple and direct numerical treatment, so


Page 22

that it is much easier to solve the equation (I – µG 1 ) u1 = f instead of Eq. (1.109). Then if (I – µG)–1 is bounded and ||G – G1 || ≤ ε/( µ ||(I – µG)–1 || ),

(1.113)

with 0 < ε < 1 being a small quantity, the operator (I – µG 1 )–1 is bounded too, and ||u – u1 || ≤ ε||u|| ||(I – µG1 )–1 ||/(I – µG)–1 ||, hence ||u – u1 || → 0 as ε → +0. In order to prove this, let us consider the operator (I – µG 1 ) = (I – µG) + µ(G – G1 ). Note that (I – µG)–1 (I – µG1 ) = (I – µG)–1 [(I – µG) + µ(G – G1 )] = I + µ(I – µG)–1 (G – G1 ). Then, according to the last remark of the previous paragraph and due to condition (1.113), it turns out that the operator µ(I – µG)–1 (G – G1 ) has a finite norm that is less than unit (more precisely, less than ε < 1), and hence the operator (I – µG) –1 (I – µG1 ) is invertible. This implies the invertibility of (I – µG1 ). Now we have u – u1 = u – (I – µG1 )–1 f = u – (I – µG1 )–1 (I – µG)u

= u – (I – µG1 )–1 [(I – µG1 ) + µ(G1 – G)]u

(1.114)

–1

= –µ(I – µG1 ) (G1 – G)u, and consequently

||u – u1 || ≤ ε||u|| ||(I – µG1 )–1 ||/(I – µG)–1 ||,

(1.115)

as was to be proved. The last result allows you to construct efficient algorithms for solving equation (1.108). This is well described in the literature (see, for example, Banerjee and Butterfield, 1981; Hackbusch, 1995), and we recommend that you use the most natural and direct algorithm, the so-called collocation method, which involves splitting the total domain S into N small subregions and then applying a simple quadrature formula with a constant integrand over each small subregion Sj of measure ∆Sj , j = 1, 2, . . . , N : Gu =

Z

G(x, ξ ) u(ξ ) dξ ≈

S

N X

G(x, ξj ) ∆Sj uj ≈ G1 u,

uj = u(ξj ),

j=1

ξj ∈ Sj . (1.116)

It is quite obvious that with N 1 the operator G1 represents a good approximation to G in the space X = C (S ). If you change the integral operator G in Eq. (1.108) by its approximation (1.116) and N let the variable x run through the same set of nodes {x i }N i=1 = {ξj }j=1 ; xi = ξi ∈ Si , i = 1, 2, . . . , N , then you arrive at a linear algebraic system of dimension N × N with respect to the unknowns ui = u(xi ), i = 1, 2, . . . , N : ui – µ

N X

G(xi , ξj ) ∆Sj uj = f (xi ),

i = 1, 2, . . . , N .

(1.117)

j=1

Helpful remarks 1◦ . An excellent survey of existing methods for the investigation of various types of integral equations can be found in (Polyanin and Manzhirov, 1998); and in Hackbusch (1995). Description of some efficient classical numerical methods for constructing solutions of integral equations is presented, for example, in Hackbusch (1995). 2◦ . It is rather unexpected that this elegant Fredholm theory is absolutely inapplicable for equations of the first kind. Furthermore, as we will see below, equations of the first kind possess qualitative properties opposite to equations of the second kind.


Page 23

1.6. Fredholm Integral Equations of the First Kind Let us start with the attempt to extract from the Fredholm alternative something helpful for equations of the first kind. We could see in the previous section that if G is a linear continuous compact operator, then the operator (µG – I ) is surely invertible if |µ| is small enough, which is equivalent to invertibility of (G – λI ) for sufficiently large |λ| (λ = 1/µ). Now, when studying an operator equation of the first kind Gu = f ,

(1.118)

we should estimate how small λ can be, provided the operator (G – λI ) –1 exists and is bounded. It is clear that: if λ here can be arbitrarily small, then it is very probable that the operator G is invertible. Recall that the irregular values of the parameter λ pertaining to the second case in the Fredholm alternative, when the homogeneous equation has a nontrivial solution u j ≠ 0, Guj = λuj , form the spectrum {λn } of the operator G or its eigenvalues. Evidently, for these λ = λn the operator (G – λI )–1 does not exist, since in this case there is a nontrivial solution to the corresponding homogeneous equation. Let us consider for simplicity the case of self-adjoint operator: G∗ = G. Then the set of eigenvalues can be completely described in the case of positive definite operator. DEFINITION. Let X = H be a Hilbert space. Then a linear continuous operator G : H → H is called positive definite if (Gu, u) ≥ γ 2 (u, u) ∼ (Gu, u) ≥ γ 2 ||u||2 for all u ∈ H , with some positive constant γ > 0. It is simply proved (see Mikhlin, 1964; Kantorovich and Akilov, 1980) that all eigenvalues of a self-adjoint and positive definite operator G are positive and displaced on the real axis between γ and the norm of the operator M = ||G||: γ ≤ λ n ≤ M . Perhaps, the most impressive result of the theory of linear operators is stated by the following theorem (Kantorovich and Akilov, 1980). THEOREM 1. A linear continuous (not necessarily compact) operator G : X → Y acting from a normed space X to a normed space Y is invertible if and only if the homogeneous equation Gu = 0 has only the trivial solution u = 0. Corollary. Any self-adjoint positive definite linear continuous operator in Hilbert space is invertible. Indeed, if there is a nontrivial solution u ≠ 0 to the homogeneous equation Gu = 0, this means that the point λ0 = 0 belongs to the spectral set of the operator, which is impossible when G is self-adjoint and positive definite. Unfortunately, all these results become useless when we consider compact operators, because of the following theorem. THEOREM 2. If G : X → Y (X and Y are arbitrary normed spaces) is a bounded compact linear operation, then G–1 cannot exist. Indeed, if G–1 is continuous, then I = G–1 G : X → X , as a composition of the compact (G) and the continuous (G–1 ) operators, maps any bounded set U ⊂ X onto a compact set V ⊂ X , which for the identity operator I means U = V . Thus, in this case any bounded set in X would be simultaneously compact, which is impossible if the space X is not finite-dimensional. This result is certainly disappointing, since all integral operators with regular and even weakly singular kernels are compact (see the previous section). So, the operator of the Fredholm integral equation (1.118) cannot be invertible. However, this does not mean that Eq. (1.118) cannot have solutions for special classes of right-hand sides f from a set F of a space Y : F ⊂ Y .


Page 24

Besides, it follows from Theorem 2 and the corollary of Theorem 1 that a positive compact operator cannot be positive definite. Generally, there is no constructive result that would guarantee that Eq. (1.118) possesses a solution in one or another case. However, an interesting theory for Fredholm integral operators with convolution kernels has been developed. It is genetically connected with the general theory of weak (or generalized) solutions of operator equations. Let us consider an integral equation of the first kind, which we assume, for simplicity, to hold on a finite interval of the real axis. By a trivial change of variable the equation can be made symmetric with respect to the origin: Z a Gu = f ∼ G(x – ξ ) u(ξ ) dξ = f (x), |x| < a. (1.119) –a

We assume that the kernel G(x) is given for all arguments, |x| < ∞. It is very important, for further consideration, to know qualitative properties of the Fourier image L(s ) of the kernel: Z ∞ Z ∞ 1 isx L(s ) e–isx ds . (1.120) L(s ) = G(x) e dx, G(x) = 2π –∞ –∞ We assume that the Fourier image L(s ) is continuous, positive, even and bounded, 0 < L(s ) ≤ L0 . Note that the evenness of L(s ) guarantees that the kernel G(x – ξ ) is real and symmetric, and the operator G is self-conjugate in the real Hilbert space H = L 2 (–a, a). Indeed, Z ∞ Z∞ is(x–ξ) G(x – ξ ) = L(s ) e ds = 2 L(s ) cos[s (x – ξ )] ds = –∞ (1.121) 0 Z ∞ Z ∞ =2 L(s ) cos[s (ξ – x)] ds = L(s ) eis(ξ–x) ds = G(ξ – x). 0

–∞

Now, there are various possibilities: 1 . The lower bound of L(s ) is positive: 0 < l0 ≤ L(s ) ≤ L0 . Let us prove that in this case the operator G is positive definite in L2 (–a, a), and hence, according to the corollary of Theorem 1, is invertible in this Hilbert space. Indeed, Z ∞ Z aZ a 1 L(s ) |U (s )|2 ds (Gu, u) = G(x – ξ ) u(ξ ) u(x) dx dξ = 2 π –∞ –a a Z ∞ Z a (1.122) l0 2 2 |U (s )| ds = l0 u (x) dx = l0 (u, u), ≥ 2π –∞ –a ◦

where we have used the Parseval identity (1.7) and the property (1.4). Note that to use these properties, we assume that the function u(x) is extended to the entire axis x ∈ (–∞, ∞), outside from the interval x ∈ [–a, a], by a zero value: u(x) = 0, |x| > a. Thus, in this case, G is positive definite and Eq. (1.119) has a unique solution for arbitrary f (x) ∈ L2 (–a, a). It should be noted that in this case, Eq. (1.118)–(1.119) is not a standard Fredholm equation of the first kind. Actually, if this was a Fredholm-type equation with a compact operator G, then G would not be invertible. In this sense, the Fourier transform of the kernel, L(s ), considered here determines an operator that is much like a second-kind Fredholm operator from the previous section rather than the first-kind one. In order to explain this in more detail, let us consider a rather widespread case where e (s ), L(s ) = l0 + L

e (s ) = O (s –α ), α > 1, s → ∞. L

(1.123)


Page 25

Then the kernel becomes as follows (δ (x) is the Dirac delta, see Eq. (1.62)): Z ∞ Z ∞ 1 A l0 e (s )e–isx ds = l0 δ (x)+G e (x), |L e (s )| ≤ e–isx ds + L , (1.124) G(x) = 2π –∞ 2π –∞ (|s | + 1)α

e (s ) is bounded for finite s , with the asymptotic behavior given by Eq. (1.123). since L Now it follows from Eq. (1.124) that Z ∞ Z ∞ A ds A 1 e e |L(s )| ds ≤ = , (1.125) |G(x)| ≤ α 2π –∞ 2π 0 (s + 1) π (1 – α)

e (x) is continuous and bounded, generating a compact operator in L 2 (–a, a). so the kernel G Now for the full operator we have (cf. Eq. (1.62)) Z a Z a Z a e (x – ξ )u(ξ ) dξ (Gu)(x) = G(x – ξ )u(ξ ) dξ = l0 δ (x – ξ )u(ξ ) dξ + G (1.126) –a –a –a e = l0 u(x) + (Gu)(x), e . We thus have arrived at a standard Fredholm integral equation with compact operator G of the second-kind, and the result obtained here offers an alternative treatment for secondkind Fredholm equations. Specifically, if one considers an integral equation with a bounded e (x – ξ ), continuous convolution kernel G e) u = f, (I + G

(1.127)

such that the Fourier transform of the full kernel is positive, e (s ) > 0, 1+L

and

e (s ) = O (|s |–α ), L

α > 1,

s → ∞,

(1.128)

then equation (1.126) is uniquely solvable in L2 (–a, a). 2◦ . The lower bound of L(s ) is equal to zero. The most typical case is when L(s ) vanishes at infinity (L(s ) → 0, s → ∞) as some power: L(s ) = O (|s |–α ), s → ∞, α > 0. It follows from general properties of the Fourier transform (see Sections 1.1 and 1.4) that G(x – ξ ) is not worse than a weakly singular kernel. Hence, G is a compact operator and so cannot be invertible. But ideas of the standard theory of weak (generalized) solutions of operator equations in energetic spaces may be applied to study this case. Let X = H = L2 (–a, a) and let us study equation (1.119) in the Hilbert space. We take the scalar product of Eq. (1.119) with a function u(x) ∈ L 2 (–a, a) to obtain (Gu, u) = (f , u).

(1.129)

DEFINITION. Any function u ∈ L2 (–a, a) that satisfies identity (1.129) is called a weak , or generalized solution of equation (1.119). Let us write out the expressions in Eq. (1.129) in detail, by using the Parseval identity and the convolution theorem for the Fourier transform (cf. Eq. (1.122)) (the bar denotes a complex conjugate): Z ∞ Z ∞ F (s ) U (s ) ds , (1.130) L(s ) |U (s )|2 ds = –∞

–∞


Page 26

where F (s ) is the Fourier transform of the right-hand side: f (x) =⇒ F (s ). Now we introduce the new (energetic) Hilbert space with the scalar product Z ∞ (u, v )e = L(s ) U (s ) V (s ) ds . (1.131) –∞

By standard methods (see Mikhlin, 1964) it can directly be proved that this expression possesses all required properties of a scalar product in a Hilbert functional space. We thus have constructed a new functional linear space and the only missed property is completeness. To complete this space, let us supplement the space by all its limit points. The thus constructed space He becomes a standard Hilbert space, which is a complete subspace of L2 (–a, a). The variational identity (1.129)–(1.130) can be rewritten now as (u, u)e = (f , u).

(1.132)

Let us study the solvability of our operator equation written in the form (1.132), by using the following estimate: Z ∞ Z ∞ p F¯ (s ) ¯ ds L(s ) U (s ) √ |(f , u)| = F (s ) U (s ) ds = L(s ) –∞ –∞ (1.133) 1/2 1/2 Z ∞ Z ∞ |F (s )|2 2 ds . L(s ) |U (s )| ds = B ||u||e , B = ≤ B –∞ L(s ) –∞

It is now evident that the linear functional (f , u) on the right-hand side is bounded (continuous) if the constant B is finite. It certainly depends on the asymptotic behavior of F (s ) and L(s ) at infinity, because these functions provide the regularity of the integrand in the constant B at any finite point, including s = 0, if we assume that the constant L(0) is finite and positive: 0 < L(0) < ∞. (1.134)

The most typical behavior of the symbolic function L(s ) at infinity is that of a power-law function: L(s ) = O (|s |–α ), α > 0, (1.135) as follows from results of Section 1.4 (see formula (1.73) in the case S (x) ≡ x, λ = s ). We will distinguish between three different cases. 2.1. 0 < α < 1. Here the constant B is finite, B < ∞ at least if f (x) ∈ C 1 [–a, a], since in this case F (s ) = O (1/s ), see Eq. (1.73), and so |F (s )|2 = O (|s |α–2 ), L(s )

(1.136)

which is integrable over the line –∞ < s < ∞. 2.2. 1 ≤ α < 3. Here a necessary condition for the constant B to be finite is that the asymptotics of F (s ), s → ∞, must be like o(1/s ), which requires that two leading asymptotic terms in Eq. (1.73) vanish. A sufficient condition for this is satisfied if at least f (x) ∈ C2 [–a, a], and f (–a) = f (a) = 0. Indeed, under such conditions F (s ) = O (s –2 ), and hence |F (s )|2 = O (|s |α–4 ), (1.137) L(s ) which together with α < 3 provides that the constant B in Eq. (1.133) is finite. 2.3. α ≥ 3. This case is of no practical significance, so we do not pay any attention to it, although it admits quite a similar analysis as in 2.1 and 2.2.


Page 27

In all these cases, where B < ∞, the functional (f , u) is continuous in H e . Consequently, according to the classical Riesz theorem (Mikhlin, 1964; Riesz and Sz.-Nagy, 1972), there exists a unique element u0 ∈ He such that (f , u) = (u0 , u)e , and this allows us to rewrite the energetic relation (1.132) in the equivalent form (u, u)e = (u0 , u)e .

(1.138)

Hence, it is clear that in all these cases an integral equation of the first kind has a unique weak (or generalized, or energetic) solution u = u 0 ∈ He .

Helpful remarks 1◦ . Of course, the existence of a weak solution does not imply the existence of a classical one. However, it is obvious that if there is a classical solution to Eq. (1.119), it is at the same time a weak solution. 2◦ . The uniqueness of the solution is guaranteed by the positiveness of the symbolic function L(s ), both in the classical and weak sense, since if there are two solutions, Gu1 = f , Gu2 = f , then it follows from G(u1 – u2 ) = 0 that (see Eq. (1.122)) Z ∞ 1 L(s )|U1 (s ) – U2 (s )|2 ds = 0, (G(u1 – u2 ), (u1 – u2 )) = 2π –∞ (1.139) hence U1 (s ) ≡ U2 (s ) ∼ u1 (x) ≡ u2 (x). 3◦ . In a sense, Fredholm equations of the second and first kinds possess opposite properties. The former admits a clear description by the Fredholm theory, which in the case of regular kernel gives explicit-form conditions of unique solvability of the equation. Thus, the better the kernel the better qualitative properties of the equation. By contrast, equations of the first kind with regular kernels are, as a rule, unsolvable. Therefore we are faced here with an astonishing phenomenon—the worse the kernel the better qualitative properties of the equation. 4◦ . The most interesting and important for applications in diffraction theory is the case when 1 D 1+O , s → ∞. (1.140) L(s ) = |s | s

Unfortunately, the theory discussed here states a (weak) solvability of Eq. (1.119) only for those right-hand sides that vanish at the ends of the interval: f (±a) = 0, f (x) ∈ C 1 [–a, a]. However, in this case formulas (1.82)–(1.85) explicitly describe the behavior of the kernel as follows: G(x) ∈ C [–2a, 2a] \ {0}; (1.141) G(x) ∼ –D ln |x| [1 + O (x)] , x → 0,

(1.142)

i.e., we arrive at equations with logarithmic-type kernels. The theory of integral equations with such kernels is closely connected with Cauchy-type integral equations and admits an absolutely different and much more advanced study based on the Riemann boundary value problem and the principal value of the Cauchy-type integrals. This theory is presented in the next section. 5◦ . Note that the described poor theory of equations of the first kind does not give you any appropriate algorithm how to treat these equations numerically.


Page 28

6◦ . It is very interesting to understand in more detail what the real structure of the kernel in case 2.1 is, when a generalized solution exists being unique for arbitrary smooth right-hand side. If we assume that 1 D , s → ∞, (1.143) L(s ) = α 1 + O |s | s

then formulas (1.82)–(1.83) determine the behavior of the kernel, G(x) ∼

DΓ(1 – α) [1 + O (x)] , |x|1–α

x → 0,

(1.144)

which is a classical Fredholm operator with weak singularity in the kernel.

1.7. Singular Integral Equations with a Cauchy-Type Singularity in the Kernel Integral operators of the type (Gu)(x) =

Z

a –a

u(ξ ) dξ , x–ξ

x ∈ (–a, a),

(1.145)

do not fall into any class among those considered in the previous two sections. If we calculate the Fourier image of the kernel (which is certainly a convolution-type kernel), then we obtain Z ∞ Z ∞ isx sin sx e dx = 2 dx = π sign(s ), (1.146) L(s ) = x x 0 –∞

which does not decrease at infinity, i.e., as s → ∞. Another problem is related to the fact that the integral (1.146) does not exist in any classical sense. If the density u(x) is at least from the Lipschitz class, u(x) ∈ Lip[–a, a], i.e., |u(x1 ) – u(x2 )| ≤ A |x1 – x2 |, ∀x1 , x2 ∈ [–a, a],

with a certain constant A > 0, then Z a Z a dξ u(ξ ) – u(x) dξ + u(x) , (Gu)(x) = x – ξ x –ξ –a –a

(1.147)

(1.148)

and we can reduce the integral (1.145) to the case with constant density, because the first integrand in Eq. (1.148) is bounded and hence gives a finite integral. It should be noted that all the results below in this section are valid even in a more general case when u(x) is from the Hölder class H δ with some positive exponent δ : |u(x1 ) – u(x2 )| ≤ A|x1 – x2 |δ (δ > 0). Let us try to treat the last integral in the generalized sense (see Eq. (1.54)): Z a Z a |x – ξ |α dξ dξ = lim α→+0 –a x–ξ –a x – ξ Z x Z a α–1 α–1 (x – ξ ) dξ – (ξ – x) dξ = lim α→+0

–a

x

(x + a)α – (a – x)α = lim α→+0 α = lim [(x + a)α ln(x + a) – (a – x)α ln(a – x)] =

(1.149)

α→+0

= ln

a+x , a–x

x ∈ (–a, a),


Page 29

where we have used the L’Hospital rule. Occasionally, it happens that the final result coincides with that trivially obtained when neglecting for a moment that the integrand is not integrable: Z a a a +x a+x dξ = – ln |ξ – x| = ln if x ∈ (–a, a). (1.150) = ln a–x a–x ξ = –a –a x – ξ

In order to clear up completely the treatment of this sort of integrals, Cauchy proposed to consider the so-called principal value. DEFINITION. If the limit lim

ε→ +0

Z

x–ε

–a

u(ξ ) dξ + x–ξ

Z

a x+ε

u(ξ ) dξ x–ξ

(1.151)

is finite, then it is called the principal value of the integral (1.145). The definition for Cauchy-type integrals over arbitrary contour is similar. The key step consists in deleting a small symmetric ε-neighborhood in Eq. (1.151). It is directly seen that if u(x) ∈ Lip[–a, a], then the principal value (1.151) exists and coincides with expressions (1.149)–(1.150). Cauchy-type integrals, which will also be called here singular integrals, possess very interesting unique properties. The most interesting properties of singular integrals are: 1) their analytic properties in the complex plane; 2) explicit inversion of the characteristic singular operator; 3) relation between the Fredholm theory and the theory of singular integral equations. First of all, let us consider an arbitrary closed contour Γ ⊂ D of finite length in some domain D of the complex plain z = {Re z , Im z }. Then the theory of analytic functions states that Z u(ζ ) dζ 1 u(z ), z inside Γ, = (1.152) 2πi Γ ζ – z 0, z outside Γ, provided that the function u(z ) is analytic in D and the contour Γ is traced counterclockwise. This result directly follows from the residue theorem (see formula (1.11)), since ζ = z is a simple pole of the integrand. But what happens at the contour, z ∈ Γ? The Sokhotsky–Plemelj formula gives the answer to this question even if u(ζ ) is not analytic but only u(ζ ) ∈ Lip(Γ): Z Z u(ζ ) dζ u(z ) 1 u(ζ ) dζ 1 =± + , z ∈ Γ, (1.153) lim z→z0 ∈Γ 2πi Γ ζ – z 2 2πi Γ ζ – z0

where the plus sign corresponds to the interior limit and the minus sign to the exterior limit, and the last integral is treated as a singular one. Note that in the case where u(ζ ) is analytic, the integral is equal to zero. The proof is very simple: Z Z Z u(ζ ) – u(z0 ) dζ u(ζ ) dζ = dζ + u(z0 ) = 0. (1.154) ζ – z ζ – z ζ – z0 0 0 Γ Γ Γ

The integrand of the first integral here is analytic and hence is equal to zero owing to the Cauchy integral theorem, but the last one is calculated directly to be trivial. We thus come to the statement that if u(z ) is analytic in a domain containing a closed contour Γ, then  Z u(z ), z inside Γ, u(ζ ) dζ  1 1 = 2 u(z ), z on Γ, (1.155)  2πi Γ ζ – z 0, z outside Γ. © 2005 by CRC Press LLC

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Therefore, we can finally conclude that the analytic properties of a Cauchy-type integral are such that it takes a “jump” every time the point z crosses the integration contour Γ. Moreover, if z passes from outside to inside of Γ, then this jump is equal to the residue at the point z ; but if z , in its passage from outside to inside of Γ, stays just on the contour, then the jump is equal to half of the residue (taken with appropriate sign). The same properties are valid for a singular integral along an arbitrary nonclosed contour of finite length Z zB u(ζ ) dζ , (1.156) ζ –z zA

i.e., it has discontinuities every time when z crosses the contour (z A , zB ), and the “full jump” when passing across the contour is the residue of the integrand (with appropriate sign), while it is half of the residue if the point z “stops” on the contour. Let us however come back to the characteristic singular integral equation holding over the interval (–a, a) Z a u(ξ ) dξ = f (x), |x| < a, f (x) ∈ Lip[–a, a], (1.157) –a x – ξ

which is the main subject of our investigation. It is proved (see, for example, Gakhov, 1966) that the general integrable solution to this equation has the following form: Z a√ 2 2 a – ξ f (ξ ) 1 C– dξ , |x| < a, (1.158) u(x) = √ x–ξ π 2 a2 – x 2 –a

where C is an arbitrary constant. The inversion formula (1.158) is a conclusion from a very thorough theory of the Riemann boundary value problem for analytic functions. Therefore, in the case of characteristic Cauchy-type kernel the considered singular equation admits an explicit-form exact solution by quadrature. But what can be said about the solvability of the full equation Z a 1 + K (x, ξ ) u(ξ ) dξ = f (x), |x| < a, (1.159) (Gu)(x) = x–ξ –a if the kernel K (x, ξ ) is regular? All results below are valid when at least K (x, ξ ) ∈ C1 [(–a, a) × (–a, a)] and f (x) ∈ Lip[–a, a]. The theory of singular equations (1.159) — and also more general equations of the second kind, when the unknown function u(x) is present also outside of the integral—was developed by Noether, and it is now known in the literature as the Noether theory (see Gakhov, 1966). As applied to an equation written in the special form (1.159) it can be represented by the theorem that follows below, similar in a sense to the Fredholm theorems. Let us consider the conjugate equation Z a 1 + K ∗ (x, ξ ) v (ξ ) dξ = g (x), |x| < a. (1.160) (G∗ u)(x) = x–ξ –a

The Noether theory requires that the conjugate of Eq. (1.160) be considered in a conjugate class of functions as follows. If we seek a solution of the main equation (1.159) in the class h of functions unbounded at both endpoints x = ±a, then the conjugate equation (1.160) should be considered in a conjugate class h 0 of functions bounded at one of two edges x = a or x = –a. Let us also introduce an index of the problem χ; in the considered first-kind case (1.159) it is simply the number of the intervals where the equation holds, i.e., χ = 1 here. Then the Noether theory states the following (compare with the Fredholm theory).


Page 31

THEOREM 1 (FIRST NOETHER THEOREM). The subspaces of solutions to the homogeneous equations Gu = 0, G∗ v = 0 (1.161) have finite dimensions n and n0 , respectively, where n – n0 = χ, i.e., n – n0 = 1 in our problem. THEOREM 2 (SECOND NOETHER THEOREM). A nonhomogeneous equation Gu = f

(1.162)

has a solution of the class h if and only if (f , vj ) = 0 for all solutions vj of the conjugate equation G∗ v = 0 in the class h0 . Let us verify the correctness of these results by the example of the characteristic equation (1.157). As we can see, n = 1 in this example, as it follows from Eq. (1.158). At the same time, the conjugate operator G∗ here coincides with the original one G, with the same solution (1.158). It is obvious that the latter has only the trivial solution of the class h 0 for the trivial right-hand side. Thus, we have n = 1, n0 = 0 =⇒ n – n0 = 1. A union of so powerful mathematical instruments as generalized treatment of integrals and operation with singular integrals can be very fruitful for many applications in pure mathematics and mathematical physics. For example, let us give an explicit solution of the Wiener–Hopf equation (1.28) in the case of the above considered kernel with the Fourier √ 2 2 image L(s ) = 1/ s + d . As we could see from (1.32), (1.39), in our case this Wiener–Hopf equation in Fourier images has the form h√ i h√ i Φ (s ) √+ = d + is F+ (s ) + d + is F– (s ). (1.163) – – d – is

Further, the decomposition of the function i h√ d + isz , F+ (s ) = N+ (s ) + N– (s ), –

(1.164)

according to the Sokhotsky–Plemelj formulas (Gakhov, 1966; Muskhelishvili, 1965), can be written as follows: √ 1 (1.165) N+ (s ) = (I + S ) d + is F+ (s ), 2 where I is the identity operator and S is a singular integral operator with a Cauchy kernel. Equation (1.163) then becomes √ Φ (s ) √+ – N+ (s ) = N– (s ) + d + is F– (s ) ≡ 0, d – is

(1.166)

and consequently Φ+ (s ) =

√

d – is N+ (s ) =

h√ i √ 1√ 1√ d – is d + is F+ (s ) + d – is S d + is F+ (s ) . (1.167) 2 2

In order to perform the Fourier inversion in the last identity, the following tabulated operational relations may be used (Bateman and Erdelyi, 1954): –d |x| –d |x| √ √ e e 1 1 , , (1.168) d – is ⇐= – √ d + is ⇐= – √ 3/2 2 π |x| 2 π |x|3/2 – +


Page 32

where the plus sign means that the support of the corresponding function is the positive semi-axis, and the minus sign means the same, with the negative semi-axis. Of course, integrals arising in these relations are understood in the generalized sense (see Section 1.3). Now, returning to origins gives –d |x| –d |x| e e 1 ∗ ∗ f (x) ϕ(x) = 8π |x|3/2 + |x|3/2 – –d |x| (1.169) –d |x| e e ∗ sign(x) ∗ f (x) . + |x|3/2 + |x|3/2 – Here we have used the following relation: sign(x) g (x) =⇒ (SG)(s )

if

g (x) =⇒ G(s ),

(1.170)

which directly follows from the convolution theorem and the formula (1.146) for the Fourier transform of the Cauchy-type singular kernel studied in this section. Recall that the sign ∗ stands for the convolution of two functions. Identity (1.169) can be written out more explicitly: Z 0 –d(x–y) Z ∞ d(y–t) e dy e f (t) 1 dt ϕ(x) = 3/2 8π –∞ (x – y ) (t – y )3/2 0 Z ∞ d(y–t) Z x –d(x–y) e f (t) e dy dt + 3/2 3/2 ( x – y ) ( t – y ) y 0 (1.171) Z ∞ d(y–t) Z 0 –d(x–y) e f (t) e dy dt – 3/2 (t – y )3/2 0 –∞ (x – y ) Z ∞ d(y–t) Z x –d(x–y) e f (t) e dy dt , + (x – y )3/2 y (t – y )3/2 0 or

1 ϕ(x) = 4π

Z

0

x

e–d(x–y) dy (x – y )3/2

Z

y

∞

e d(y–t) f (t) dt, (t – y )3/2

x > 0.

(1.172)

Both integrals here are treated in the generalized sense. Let us make sure that in the case f (x) = e–βx this result coincides with (1.44). To this end, we ought to calculate the internal integral (Gradshteyn and Ryzhik, 1994), which is again treated as a generalized value (see Section 1.3) Z ∞ –(d+β)t Z ∞ d(y–t) –βt e 1 p e e –βy dt = e dt = Γ – d + β e–βy , (1.173) (t – y )3/2 t3/2 2 0 y

and the generalized value of the next integral can be calculated with the help of integration by parts Z x –(d–β)y Z x –d(x–y) –βy e e e –βx dy = e dy 3/2 ( x – y ) y 3/2 0 0 Z x 2 –(d–β)x –(d–β)y dy –βx (1.174) – 2(d – β ) e =e –√ e √ x y 0 p p 2 = – √ e–dx – 2 π (d – β ) e–βx Erf (d – β ) x , x © 2005 by CRC Press LLC

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so the solution we seek is p ϕ(x) = d + β

hp i e–dx p –βx √ + d – β e Erf (d – β )x , πx

(1.175)

which coincides with (1.44). The described theory of Cauchy singular integrals gives also an efficient tool for solving convolution equations with kernels containing logarithmic singularity, which could not be constructed completely by methods of first-kind Fredholm operators (see remark 4 ◦ in the previous section). First of all, it is clear that the characteristic equation Z a ln |x – ξ | u(ξ ) dξ = f (x), |x| ≤ a, (1.176) –a

is reduced, after differentiation with respect to x, to Eq. (1.157), which admits explicit inversion. The unknown constant C arising during these transformations can be uniquely determined by substituting an expression of the type (1.158) (with f 0(x) substituted for f (x)) into the initial equation (1.176) with a certain value of x, for example, at x = 0. This procedure was first performed by Carleman (Carleman, 1922), who gave the following representation for the (unique) integrable solution: Z a √ 2 2 0 Z a 1 f (ξ ) dξ a – ξ f (ξ ) 1 √ √ dξ + , |x| < a, (1.177) u(x) = ξ–x ln(a/2) –a a2 – ξ 2 π 2 a2 – x2 –a in the case a ≠ 2, f 0 (x) ∈ Lip[–a, a]. In the case a = 2 the solution has a slightly different representation. Then, if we study a full equation (of the first or second kind) with a logarithmic-type kernel, considerations here follow the strategy applied for full singular equations with the Cauchy-type kernels. You can apply explicit inversion of the characteristic logarithmic kernel, reducing the problem to a Fredholm equation of the second kind, whose theory is covered by the Fredholm theorems (see Section 1.5).

Helpful remarks 1◦ . Recall remark 3◦ in the previous section. Here we continue to trace the wonderful feature of integral equations of the first kind—worse kernels yield wider classes of solutions. Singular kernels in this sense are even more advanced than kernels with weak singularities (see Section 1.6), since they yield a wide one-parameter class of solutions with arbitrary constant C rather than a unique solution. Thus, singular integral equations are closer to first-order ordinary differential equations than to classical Fredholm-type integral equations. 2◦ . It is very interesting to note that the Noether theory is quite similar to that of Fredholm. The only difference lies in the fact that dimensions of the nuclear subspaces are different for the main and the conjugate operators. 3◦ . The Noether theory remains valid for an arbitrary singular integral equation of the second kind containing a Cauchy-type operator and a regular part of the kernel. The only difference from the described first-kind case (which is of key importance for further context) lies in a more specific definition of the index χ. So, there is essentially no difference between the first- and the second-kind operators in the theory of singular integral equations in contrast to Fredholm-type integral equations. The solvability problem for the first- and the second-kind operators is described by the same Noether theory.


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1.8. Hyper-Singular Integrals and Integral Equations DEFINITION. A convolution integral of the form Z a u(ξ ) dξ , (Gu)(x) = 2 –a (x – ξ )

|x| < a,

(1.178)

is called a hyper-singular integral . First of all, we should clarify in which sense hyper-singular integrals as given by Eq. (1.178) may be treated, since they exist neither as improper integrals of the first kind nor as Cauchy-type singular integrals. At least three different definitions of hyper-singular integrals are known (see Samko, 2000): 1. The integral is a derivative of the Cauchy principal value: Z a Z a d u (ξ ) u(ξ ) dξ =– dξ . (1.179) 2 dx –a x – ξ –a (x – ξ ) 2. The integral is treated as a Hadamard principal value (see Belotserkovsky and Lifanov, 1993): Z x–ε Z a Z a u (ξ ) dξ 2u(x) u(ξ ) dξ = lim + . (1.180) – 2 ε→ +0 ε (x – ξ )2 –a x+ε –a (x – ξ ) 3. The integral is a residue value, in the sense of generalized functions (see Section 1.3), which is an analytic continuation of the integral Z a α |x – ξ| u (ξ ) dξ , (1.181) –a

from a domain where it exists in the classical sense to the value α = –2. When u(x) ≡ 1, the three different approaches give the same result: Z a Z a a–x d dξ d 2a dξ = – = ln =– 2 , (1.182) 2 dx –a x – ξ dx a+x a – x2 –a (x – ξ ) Z a 1 1 1 2 2a 1 dξ + + + , (1.183) – =– 2 = lim – 2 ε→ +0 a + x ε ε x – a ε a – x2 –a (x – ξ ) Z a Z a (a + x)α+1 (a – x)α+1 α + , (1.184) |x – ξ| dξ = (ξ – x)α dξ = α+1 α+1 –a x when Re(α) > –1. The analytic continuation of (1.184) from the half-plane Re(α) > –1 to the value α = –2 gives again Z a (a + x)α+1 (a – x)α+1 2a dξ lim + =– 2 . (1.185) 2 = α→ –2 α + 1 α + 1 a – x2 –a (x – ξ ) All three definitions are also equivalent to each other if the density u(x) ∈ Lip[–a, a], that is u0 (x) ∈ Lip1 [–a, a]. Indeed, let Re(α) > –1 then, according to the definition of the


Page 35

Cauchy principal value of singular integrals, we have Z x–ε Z a Z a u(ξ ) dξ d u(ξ ) dξ d = – lim + – ε→ +0 dx dx –a x – ξ x–ξ Z x–ε –a Z a x+ε u(ξ ) dξ 2ϕ(x) – , = lim + ε→ +0 (x – ξ )2 ε –a x+ε Z x–ε Z a Z a |x – ξ |α u(ξ ) dξ = lim + |x – ξ |α u(ξ ) dξ (1.186) ε→+0 –a –a x+ε Z x–ε Z a α α = lim (x – ξ ) u(ξ ) dξ + (ξ – x) u(ξ ) dξ ε→ +0 –a x+ε Z x–ε Z a (x – ξ )α+1 (ξ – x)α+1 d lim – u(ξ ) dξ + u(ξ ) dξ , = dx ε→ +0 α+1 –a x+ε α + 1

where we have used the rule of differentiation of integrals containing outer variable both in the limit of integration and in the integrand. Now, by applying the analytic continuation to the last relation, one can see (with the help of the standard (ε, δ ) formalism) that the right-hand side results in (1.179). Therefore, equivalence of definitions 1, 2 and 3 (in case all these integrals are finite) is evident. Let us prove that if u(x) ∈ Lip 1 [–a, a], then a finite value of the limit in expression (1.180) exists, and hence for x ∈ (a, b) the integral (1.178) is finite in any sense. Indeed, the expression in square brackets in Eq. (1.180) is Z x–ε Z a 2a a–x u(ξ ) – u(x) – u0 (x)(ξ – x) , (1.187) dξ – u(x) 2 + u0 (x) ln + 2 2 (x – ξ ) a –x a+x x+ε –a

which has a finite limit at ε → +0. Let us proceed however to the investigation of the hyper-singular equation with the characteristic kernel Z a u(ξ ) dξ = f 0 (x), |x| < a, f (x) ∈ Lip1 [–a, a], (1.188) 2 –a (x – ξ )

and construct its bounded solution with the help of relation (1.179) and the inversion of the characteristic singular integral operator (see Section 1.7). Indeed, by definition (1.179), equation (1.188) is equivalent to Z a Z a u(ξ ) dξ u(ξ ) dξ d 0 = –f (x) ∼ = –f (x) + C1 , |x| < a, (1.189) dx –a x – ξ –a x – ξ

where C1 is an arbitrary constant. Now, the inversion formula for the Cauchy characteristic integral operator (see the previous section) determines the function u(x) in a form containing another arbitrary constant C2 , in addition to C1 . This pair of constants should be chosen so as to provide finite values of the solution at both endpoints x = ±a. This can directly be made, resulting in the following THEOREM. If f (x) ∈ Lip1 [–a, a], then the bounded solution of Eq. (188) is unique and given by the following expression: √ Z f (ξ ) dξ a2 – x2 a √ 2 2 , |x| < a. (1.190) u(x) = 2 π a – ξ (x – ξ ) –a You can also directly verify that the above holds true for Z a f (ξ ) dξ 1 √ . C1 = 2 π –a a2 – ξ 2

(1.191)


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It is interesting to note that the bounded solution of Eq. (1.188) vanishes as x → ±a. It is obvious that the full hyper-singular equation Z a 1 + K (x, ξ ) u(ξ ) dξ = f (x), |x| < a, (1.192) (Gu)(x) = (x – ξ )2 –a

where the kernel K (x, ξ ) is regular, can be studied by explicit inversion of its characteristic part. Investigation of this problem is similar to that for full Cauchy-type or logarithmic-type equations. In the final part of this section we study analytic properties of hyper-singular integrals Z u(ζ ) dζ , (1.193) 2 Γ (ζ – z ) as functions of a complex-valued argument z with density u(z ) analytic in some domain D containing a simple closed contour Γ of finite length. Here we would like to give a comparison with analytic properties of the Cauchy-type singular integral (1.155). Let us start with the case of a constant density and prove that the integral Z dζ ≡0 (1.194) 2 Γ (ζ – z ) is identically equal to zero. When z lies inside the contour, this result is evident because the residue of the integrand (i.e., the factor in front of the first term in the Laurent series expansion in negative powers of (ζ – z )) is equal to zero. When z lies outside Γ this is evident due to analyticity of the integrand inside Γ. If z ∈ Γ, it follows immediately from the first two of three above techniques that this is trivial indeed. Now, for arbitrary analytic density we have Z Z Z u(ζ ) – u(z ) v (ζ ) dζ u(ζ ) – u(z ) u(ζ ) dζ = dζ = , v (ζ ) = , (1.195) 2 2 (ζ – z ) ζ –z Γ Γ ζ –z Γ (ζ – z )

where v (ζ ) is now analytic, and we can use appropriate results from the Cauchy-type integral theory. It is obvious that v (ζ )| ζ=z = v 0 (z ), so formula (1.155) yields  Z 0, z inside Γ, u(ζ ) dζ  1 0 1 = 2 u (z ), z on Γ, (1.196)  2πi Γ ζ – z 0, z outside Γ.

Helpful remarks

Let us verify formula (1.196) by the example of the integral Z ∞ ibs e ds , b > 0, –∞ < t < ∞. I (t) = 2 ( –∞ s – t)

(1.197)

Since the function eibs with b > 0 exponentially decays in the upper half-plane Im(s ) > 0, we can close the integration contour by a semi-circle of infinite radius. Then the value of this integral, according to Eq. (1.196), is I (t) = –πb e ibt . At the same time, direct calculations show that Z ∞ ibs Z ∞ ib(s+t) Z ∞ ibs e ds d e ds d e ds d ibt = = e I (t) = dt –∞ s – t dt –∞ s dt s –∞ (1.198) Z ∞ sin(bs ) ibt ibt ibt ds = (ib e )(πi) = –πb e , = (ib e ) 2i s 0 which is exactly the same result.


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By the way, extracting separately the real and imaginary parts of this integral, you can evaluate the following real-valued integrals: Z ∞ Z ∞ sin(bs ) cos(bs ) ds = – π | b | cos( bt ), ds = –π |b| sin(bt), (1.199) 2 2 –∞ (s – t) –∞ (s – t) valid for arbitrary real b.

1.9. Governing Equations of Hydroaeroacoustics, Electromagnetic Theory, and Dynamic Elasticity Linear Hydroaeroacoustics The governing equations of linear acoustics are essentially linearized equations of general nonlinear fluid and gas dynamics with small perturbation velocities. Let a fluid or a gas have under the at-rest conditions a constant density ρ 0 , and a uniform pressure p0 , and zero velocity v0 = 0. Then the total pressure is p = p0 + p0 , the total density ρ = ρ0 + ρ0 , and the net velocity vector v = v0 , where all quantities marked with a prime are small, and we should perform linearization with respect to these small parameters. Then Euler’s equations of motion ρ

∂v + ρ(v ⋅ ∇)v + grad p = 0, ∂t

(1.200)

in the case of small perturbations (linearized equations), have the following form: ρ0

∂ v0 + grad p0 = 0. ∂t

(1.201)

Let us assume that the fluid is barotropic. Then the constitutive equation p = p(ρ) ∼ ρ = ρ(p) in the linearized form is expressed as ∂p (ρ – ρ0 ) ∼ p0 = c2 ρ0 + const, (1.202) p = p0 + ∂ρ ρ=ρ0 where the quantity c,

∂p , c = ∂ρ ρ=ρ0 2

(1.203)

is called the wave speed in the medium. It is easily seen that the dimension of the constant physical parameter c is m/s, which justifies its name. Besides, it is clear for physical reasons that for real media the dependence of density upon pressure is monotonically increasing, so ∂p/∂ρ > 0, which yields c2 > 0. Therefore, the wave speed is always real and positive. The continuity equation ∂ρ + div(ρv) = 0, (1.204) ∂t in the linearized form is ∂ρ0 + ρ0 div(v0 ) = 0. (1.205) ∂t Let us derive from Eqs. (1.201), (1.202), (1.205) the wave equation (the primes are omitted from now on). To this end, let us differentiate Eq. (1.205) with respect to time, and


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then substitute the expression ρ0 ∂ v/∂t, taken from Eq. (1.201), to the obtained equation (∆ is the Laplacian operator): ∂2ρ = div(grad p) ∂t2

∂2ρ = ∆p. ∂t2

∼

(1.206)

Then from the last equation and Eq. (1.202), we obtain the wave equation ∂2p = c2 ∆p. ∂t2

(1.207)

In the case when the wave process is harmonic in time, p(x, y , z , t) = Re{e –iωte p(x, y , z )}, where ω is the angular frequency and the new function e p is independent of time, we obtain the Helmholtz equation (the signs of real part and the tildes are further omitted) ∆p + k 2 p = 0,

k=

ω , c

(1.208)

where k is called the wave number. Note that in the forthcoming consideration the timedependent factor e–iωt will be omitted every time we study a harmonic problem. In this harmonic case the velocity vector can be expressed in terms of pressure from Eq. (1.201) as follows: 1 grad p . (1.209) v= iωρ0 Let {xi }, i = 1, 2, 3, be a fixed system of rectangular Cartesian coordinates. In the harmonic regime of oscillations with an angular frequency ω , when the complex amplitude of pressure (or, simply the pressure function) satisfies Eq. (9), we introduce the term plane wave as a solution to the Helmholtz equation (1.208) of the following form: p = f (n ⋅ r) = f (n1 x1 + n2 x2 + n3 x3 ), ∆p =

(n21 + n22 + n23 )f 00

r = {x1 , x2 , x3 },

= f , hence f + k 2 f = 0 ∼ f = A1 eik n⋅r + A2 e–ik n⋅r , 00

00

(1.210)

where n is an arbitrary fixed unit vector in the space, so that the full structure of the solution corresponding to the plane wave is p = A1 ei(kn⋅r–ωt) + A2 e–i(kn⋅r+ωt) ,

(1.211)

where the argument of an oscillating exponential function is called the phase of the corresponding wave. DEFINITION. For any moment t, a set in the three-dimensional (3D) space that corresponds to a constant value of the phase is called a wave front set. The velocity of propagation of this wave front in time is called the phase velocity . Let us show that the first term in Eq. (1.211) represents a wave going away to infinity, and the second term a wave arriving from infinity. Indeed, the wave front of the two terms is a set, where k n ⋅ r ∓ ωt = const ∼ n ⋅ r = const ± c t, (1.212)

since ω/k = c. It is known from analytic geometry that if a unit direction vector is n = {cos α, cos β , cos γ }, then the equation x1 cos α + x2 cos β + x3 cos γ – q = 0 defines a plane with normal n and distance from the origin equal to q . In our case n = {n 1 , n2 , n3 }, and the distance from the origin is equal to q (t) = const ± c t. Thus, the wave front represents


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a plane, which justifies the name of the wave, and the direction n determines a unit normal to this plane front. Moreover, the velocity of propagation of this front is q˙ (t) = ±c,

(1.213)

which defines the velocity of the wave front propagation. It is no accident that it coincides with the wave speed in the medium. Now it becomes clear that the upper sign is related to the wave that travels away from the origin to infinity, and the lower sign, to that arriving from infinity. A special discussion in Chapter 3 will be devoted to the important question what is physically and mathematically the intrinsic difference between these two types of waves. Here we only mention that throughout the text, diffracted (scattered, reflected) waves must satisfy the so-called radiation condition, whose strict formulation will be given below. In the simplest intuitive sense, the radiation condition reads: among the number of possible solutions in diffraction problems, only diffracted waves going away to infinity are physically realizable, and the waves arriving from infinity are fictitious. Thus, out of two plane waves in Eq. (1.211), only the first one satisfies the radiation condition. The wavelength λ is determined as the minimum distance in the direction n for which the phases of oscillations (i.e., the arguments of the exponential functions in (1.211)) coincide at the points defined by the radius vector r and the radius vector r + λn. This implies λk = 2π , or c 2π 2πc = = , where ω = 2πf . (1.214) λ= k ω f Here ω is the angular frequency, which is measured in rad/s, and f is called the cyclic frequency and is measured in Hz = 1/s. In diffraction problems where an incident wave falls onto an obstacle, complete mathematical formulation implies some boundary conditions on the boundary surface (3D case) or boundary line (2D case) of the obstacle. Typically, there are two types of boundary conditions: 1) acoustically hard boundary and 2) acoustically soft boundary. In the first case the medium particles cannot penetrate through the boundary, so the normal component of the velocity vanishes, vn = 0, which with the help of Eq. (1.209) means ∂p = 0. (1.215) ∂n S In the second case it is assumed that the total pressure on the boundary is equal to zero: p|S = 0.

(1.216)

Electromagnetic wave theory Here the governing equations for an isotropic medium are given by the system of Maxwell’s equations  µ ∂H   rot E = – , div E = 0, c ∂t (1.217)   rot H = σ E + ε ∂ E , div H = 0, c c ∂t where E is the electric field strength, H is the magnetic field strength, ε, µ, σ are some positive constant physical parameters, and c is the wave speed. If we apply the operator rot to the first equation (1.217), and eliminate the magnetic strength, then we arrive at the following equation with respect to E: ∆E =

µε ∂ 2 E µσ ∂ E + 2 , c2 ∂t2 c ∂t

(1.218)


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where we have taken into account that rot rot E = grad div E – ∆E and the additional condition div E = 0. If the wave process is harmonic in time, with the usual time-dependent factor e –iωt , then Eq. (1.218) becomes the Helmholtz partial differential equation r σ ω 2 (Re k > 0, Im k ≥ 0), (1.219) µε + i ∆E + k E = 0, k = c ω

with some (generally) complex-valued wave √ number k , which in the case of vacuum medium (σ = 0) becomes real-valued: k = (ω/c) µε. The magnetic strength H is expressed in terms of E as follows: c rot E. (1.220) H= iωµ The boundary conditions for Eq. (1.219) usually correspond to the surface of a perfect conductor S , where the tangential component of E vanishes: n × E|S = 0.

(1.221)

Linear dynamic elasticity If a linear isotropic elastic medium occupies some domain in the 3D space, then at every point of the medium there are three components of the displacement vector u = {u 1 , u2 , u3 } and six components of the stress tensor. The latter can be represented as a 3 × 3 symmetric matrix {σij }, i, j = 1, 2, 3, σij = σji whose components σij determine the j th stress component arising on a small area element crossing the given point and possessing external normal parallel to xi . The constitutive equations of the linear isotropic elastic medium are 3 X ∂uk ∂ui ∂uj + + λ δij , (1.222) σij = µ ∂xj ∂xi ∂x k k=1

and the equations of motion are expressed as

∂2u , (1.223) ∂t2 where µ and λ are the Lamé elastic moduli, and ρ is the mass density of the material. It is proved that, unlike hydroaeroacoustics and electromagnetism, where physically only one wave speed c can exist, two independent elastic wave velocities, µ∆u + (λ + µ) grad div u = ρ

c2p =

µ , ρ

c2s =

λ + 2µ , ρ

(1.224)

can be introduced in dynamic elasticity. These are the so-called longitudinal and transverse wave speeds, respectively. Then, according to classical results (see, for example, Achenbach, 1973), the following Lamé potential representation gives a general solution to system (1.223): u = grad ϕ + rot ψ

(div ψ = 0),

(1.225)

where ϕ is a longitudinal scalar potential and ψ is a transverse vector potential, which satisfy the following wave equations, respectively: ∆ϕ =

1 ∂2ϕ , c2p ∂t2

∆ψ =

1 ∂2ψ . c2s ∂t2

(1.226)


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Of course, in the case of a harmonic process, with a factor e –iωt , the wave equations (1.226) become the Helmholtz equations ω ω (1.227) ∆ϕ + kp2 ϕ = 0, ∆ψ + ks2 ψ = 0, k p = , ks = , cp cs with two different wave numbers kp and ks . The most typical boundary conditions in diffraction problems for elastic waves correspond to the case of boundary surfaces free of load. This means trivial components of the stress over the boundary surface S . Since the stress vector on the elementary area is expressed in terms of components of the stress tensor, this implies 3 X Pi |S = σij nj = 0, i = 1, 2, 3. (1.228) j=1

The last condition can be reformulated in terms of the Lamé potentials, by using relations (1.222) and (1.225).

Helpful remarks 1◦ . It is interesting to note that the principal governing equation of all three theories of absolutely different physical nature are covered by identical wave or (in a monochromatic process) Helmholtz equations. The difference in the governing equations is that in acoustics the wave equation is scalar, and the total theory can be formulated in terms of a scalar pressure function. In electromagnetic theory the governing equations are vector, but all the components of the vector fields are described by the same wave equation with a single wave speed. The governing equations of dynamic elasticity are tensor, and generally they are described by a pair of wave equations with two different wave speeds. 2◦ . The mathematical uniformity of wave processes in so physically different media explains why they are usually studied as sections of a unified theory. This even results in a mutual influence of terminologies, heuristic ideas, and other things. Thus, in hydroacoustics and elastic wave theory there is a very widespread idea about light and shadow zones in geometrical diffraction theory, though there are no light rays or any other matters that could produce light or shadow. Then, harmonic processes in scalar and elastic theories are often called monochromatic (single-colored in Greek), a terminology certainly generated by the electromagnetic theory, since an electromagnetic wave of a single fixed frequency leads to a certain color in the rainbow. Such examples of genetic uniformity can be cited in various aspects and in many problems. 3◦ . It suffices for many applications to study the so-called (2D) in-plane problem, in Cartesian coordinates (x, y ). In scalar acoustics its solution is determined as a solution of the wave equation which is sought (if exists) in the form p = p(x, y , t) in transient case and p = p(x, y ) in harmonic case. Then the corresponding wave and Helmholtz equations become 2 ∂ p ∂2p ∂2p ∂2p ∂2p 2 = c + , + + k 2 p = 0 (k = ω/c). (1.229) ∂t2 ∂x2 ∂y 2 ∂x2 ∂y 2 In electromagnetism the plane problem can be separated into two independent, E polarized and H -polarized, wave fields. We will write out respective governing equations only for the harmonic case. The former type of polarization corresponds to ∂2E ∂2E + + k 2 E = 0, E = {0, 0, E (x, y )}, ∂x2 ∂y 2 (1.230) c ∂E c ∂E , Hy = – . H = {Hx (x, y ), Hy (x, y ), 0}, Hx = iωµ ∂y iωµ ∂x


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The latter type of polarization is described by the following governing equations: ∂2H ∂2H + + k 2 H = 0, ∂x2 ∂y 2 c ∂H c ∂H , Ey = . E = {Ex (x, y ), Ey (x, y ), 0}, Ex = – iωε ∂y iωε ∂x

H = {0, 0, H (x, y )},

(1.231)

In dynamic elasticity the in-plane problem implies that the displacement vector has only two first components of three, with both of them depending on (x, y ) only. It can directly be proved that in this case the vector Lamé potential ψ can be taken scalar with the only third nontrivial component depending on (x, y ), which automatically makes the condition in brackets in Eq. (1.225) valid. The complete system of governing equations here is given by u = {u(x, y ), v (x, y ), 0}, ϕ = ϕ(x, y ),

ψ = ψ (x, y ),

∂ϕ ∂ψ ∂ϕ ∂ψ + , v= – , ∂x ∂y ∂y ∂x (1.232) ∂2ψ ∂2ψ ∂2ϕ ∂2ϕ 2 2 + + kp ϕ = 0, + + ks ψ = 0. ∂x2 ∂y 2 ∂x2 ∂y 2

u=

The components of the stress tensor can be found from the Hooke law (1.222). 4◦ . In dynamic theory of elasticity there is the so-called anti-plane, or SH-wave process; this is the case when the problem can be reduced to a single scalar equation with only one (transverse) wave number ks . This is characterized by the main property that the wave is polarized in the z -direction but propagates in the (x, y ) plane. Mathematically, this is covered by the equations u = {0, 0, w(x, y )},

∂2w ∂2w + + ks2 w = 0, ∂x2 ∂y 2

σxz = µ

∂w , ∂x

σyz = µ

∂w , (1.233) ∂y

with all other components of the stress tensor being trivial. 5◦ . Due to the outlined unity of wave processes in the three mentioned classes of problems, we will operate, as a rule, with scalar acoustic wave fields in the forthcoming consideration of various theoretical aspects. Only sometimes, we will treat some elastic wave problems in order to demonstrate that the developed methods are efficient in the most complex tensor (elastic) case, too.


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Chapter 2

Integral Equations of Diffraction Theory for Obstacles in Unbounded Medium 2.1. Properties of the Potentials of Single and Double Layers As noted in Section 1.9, the main subject of our consideration is related to harmonic wave processes in scalar media, so we will mainly deal with the scalar Helmholtz partial differential equation ω (2.1) ∆p + k 2 p = 0, k= , c where ω is the angular frequency, c is the wave speed, and k is the wave number. This equation is to be solved in some (3D) domain D bounded by a simple surface S . If the problem is two-dimensional (2D), then we study Eq. (2.1) in some plane domain D with a simple boundary line l. As mentioned in Section 1.9, the standard boundary conditions may be one of the following types: (1) acoustically hard boundary, ∂p (2.2) vn |S = 0 ∼ = 0, ∂n S or (2) acoustically soft boundary,

p|S = 0.

(2.3)

DEFINITION. A fundamental solution (or a Green’s function) is a solution Φ(r ) to the Helmholtz equation that depends only on the distance r between the origin and the current point. In order to construct a Green’s function, we are obliged to treat the 2D and 3D cases separately. 3D case. The Laplace operator in the spherical coordinate system (r , ϕ, θ ) has the form 1 ∂ ∂ 1 ∂2 1 ∂ 2 ∂ r + 2 sin θ + 2 2 . (2.4) ∆= 2 r ∂r ∂r r sin θ ∂θ ∂θ r sin θ ∂ϕ2 Consequently, in the case where the only nontrivial derivatives are those applied with respect to the r -coordinate, a Green’s function can be obtained from the equation 1 d 2 d 2 r + k Φ(r ) = 0, (2.5) r 2 dr dr which is evidently equivalent to d2 (rΦ) + k 2 (rΦ) = 0. dr 2

(2.6)


Page 45

Equation (2.6) is an ordinary differential equation with constant coefficients for rΦ(r ). It admits exact analytical solution by constructing a characteristic polynomial (Smirnov, 1964), which gives rΦ = A e±ikr with a certain constant A, so that the general solution is given by e–ikr eikr + A2 . (2.7) Φ(r ) = A1 r r 2D case. Here the Helmholtz equation (2.1) becomes again an ordinary differential equation but with nonconstant coefficients, d2 Φ 1 dΦ + + k 2 Φ = 0, dr 2 r dr

(2.8)

which is a Bessel equation (Abramowitz and Stegun, 1965). Its general solution is expressed as Φ(r ) = A1 J0 (kr ) + A2 Y0 (kr ), (2.9) where J0 and Y0 are the Bessel functions of the first and second kinds (also called the Bessel and Neumann functions, respectively) of order 0. In this chapter we will consider only diffraction by finite obstacles occupying a domain D and placed in an unbounded acoustic medium, i.e., only exterior problems. For a correct selection of a unique solution, we should adopt a correct condition at infinity, which again is the radiation condition (see Section 1.9). Let us recall that the full structure of the obtained solution (2.7), as a function of the space coordinate r and time, is as follows: Φ(r , t) = A1

e–i(kr+ωt) ei(kr–ωt) + A2 . r r

(2.10)

It is clear, by analogy with what was written in Section 1.9, that the wave front corresponding to these two waves is defined by kr ∓ ωt = const for any chosen moment t, i.e., r = ±(ω/k ) t + const = ±ct + const. We thus can see that a Green’s function represents a spherical wave, since for any fixed t the obtained wave front equation represents a spherical surface. The phase velocity is r˙ = ±c, i.e., it again coincides with the wave speed in the medium. Only the first term in Eq. (2.1) satisfies the radiation condition in the considered 3D case, which gives a Green’s function in the form Φ(r ) =

eikr , 4πr

(2.11)

where the harmonic (in time) factor exp(–iωt) is omitted, and the constant A is taken equal to 1/(4π ) for convenience, which will become clear very soon. It is very important to stress that the Green’s function (2.11) is analytic and satisfies the Helmholtz equation only outside a small neighborhood of the origin. In order to make a correct choice between the two functions in the 2D case, let us quote the far-field asymptotics of Bessel functions (Abramowitz and Stegun, 1965): r π 2 cos kr – + O (r –3/2 ), J0 (kr ) ∼ πkr 4 r → ∞. (2.12) r π 2 sin kr – + O (r –3/2 ), Y0 (kr ) ∼ πkr 4 © 2005 by CRC Press LLC

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It is now evident that the only linear combination of these functions that generates a wave propagating from the origin to infinity, in accordance with the radiation condition, is Φ(r ) = A [J0 (kr ) + iY0 (kr )] = AH0(1) (kr ) r h π i 2 + O (r –3/2 ), exp i kr – ∼ πkr 4

r → +∞.

(2.13)

The expression in the first square brackets here is called the Hankel function of the first kind of order 0 (see Abramowitz and Stegun, 1965). Only for convenience, we specify the constant A and take the 2D Green’s function in the form Φ(r ) =

i (1) H (kr ). 4 0

(2.14)

Note that the Green’s function of the 2D theory (2.14) has again a singularity at the origin. DEFINITION. The integral (Gs u)(x) =

Z

Φ(|x – y |) u(y ) dsy ,

(2.15)

∂D

taken over the boundary ∂D of the domain D , is called a single-layer potential . Here |x – y | is the distance between the points x = (x1 , x2 , x3 ) and y = (y1 , y2 , y3 ) in the 3D space R3 , or between the points x = (x1 , x2 ) and y = (y1 , y2 ) in R2 . The subscript y indicates that the integration is applied with respect to y . It is evident from Eqs. (2.11), (2.14) and the properties of the Hankel function for small arguments (Abramowitz and Stegun, 1965), H0(1) (kr ) ∼

2i [ ln(kr ) + γ – ln 2 ] + 1 + O (kr )2 ln(kr ) , π

(kr ) → +0,

(2.16)

where γ is the Euler constant, that the integral operator G s in (2.15) has a weakly singular kernel (see Section 1.5). It follows that Gs is a compact operator in the Banach space of continuous functions u(y ) ∈ C (∂D ). For further consideration we will also assume that the boundary ∂D is piecewise smooth. THEOREM 1. If the density u(y ) ∈ C (∂D ), then the single-layer potential (2.15) represents a continuous function in R3 (R2 ). Proof. We give a proof only for the 3D case, and the 2D case can be studied by analogy. Since the kernel of the integral operator (2.15) is continuous for all x, y ∈ R 3 excluding a small neighborhood of the origin (when |x – y | → 0), it is evident that the integral (2.15) is certainly a continuous function of x far away from the boundary S = ∂D . Now, let x ∈ S ; then we will show that the difference |(Gs u)(x0 ) – (Gs u)(x)| can be made arbitrarily small if the difference |x0 – x| is small enough. Let ε > 0 be an arbitrary small positive parameter and δ = aε 3 (a > 0). Let us denote a part of the surface S that represents a small neighborhood of the point x with radius r ε = bε (b > 0) by Sε . The constants a and b do not depend on ε. The rest of S is S \ S ε . Now suppose |x0 – x| ≤ δ and estimate the following difference: Z Z 0 |(Gs u)(x ) – (Gs u)(x)| ≤ M + |Φ(|x0 – y |) – Φ(|x – y |)| dsy , Sε S\Sε (2.17) M = max |u(y )|. y∈S


Page 47

The first integral is Z Z 0 |Φ(|x – y |) – Φ(|x – y |)| dsy ≤ |Φ(|x0 – y |)| + |Φ(|x – y |)| dsy Sε Sε Z 1 1 + dsy (2.18) = |x – y | |x0 – y | Sε Z rεZ 2π Z ρ dρ dψ dsy = = 2πrε = 2πbε, ≤ 2 | x – y | ρ 0 0 Sε where ψ is the polar angle of the point y in the polar coordinate system arranged in a tangential plane with center x. Here we have taken into account that if the point x 0 approaches x along the normal direction, then |x 0 – y | ≥ |x – y |. The second integral in (2.17) can be estimated as follows: Z |Φ(|x0 – y |) – Φ(|x – y |)| dsy S\Sε

ik|x–y| ik|x0 –y| 0 e eik|x –y| (|x0 – y | – |x – y |) –e + = dsy |x – y | |x – y | |x0 – y | S\Sε ik|x–y| ik|x0 –y| Z Z e –e ||x – y | – |x0 – y || dsy + dsy . ≤ |x – y | |x – y | |x0 – y | S\Sε S\Sε Z

(2.19)

Further, since |x – y | – |x0 – y | ik|x–y| ik|x0 –y| ≤ ||x – y | – |x0 – y || , – e = 2 sin e 2 equation (2.19) is reduced to Z Z 0 |Φ(|x – y |) – Φ(|x – y |)| dsy ≤ S\Sε

S\Sε

||x – y | – |x0 – y || |x – y |

1 1+ 0 |x – y |

(2.20)

dsy . (2.21)

The last step is to use an inequality evident from the geometry of triangles: |¯z 1 – z¯ 2 | ≥ ||¯z1 | – |¯z2 ||, which in our case may be applied in the form ||x – y | – |x 0 – y || ≤ |x – x0 |, hence Z Z 1 1 0 0 1+ 0 dsy |Φ(|x – y |) – Φ(|x – y |)| dsy ≤ |x – x | |x – y | S\Sε S\Sε |x – y | Z dsy 0 (2.22) ≤ 2 |x – x | 0 S\Sε |x – y | |x – y | 2aS ∗ ε 2δ ≤ 2 S∗ = rε b2 (S ∗ is the square of the surface S ), because for y ∈ S \ S ε we have |x – y | ≥ rε , |x0 – y | ≥ rε (recall that x0 lies on the normal to the point x). By collecting together Eqs. (2.17), (2.18), (2.22), and setting a = 1/(64π 2 M 3 S ∗ ) and b = 1/(4πM ), we arrive at the final estimate: for arbitrary small ε > 0 there exists δ = ε3/(64π 2 M 3 S ∗ ) such that |(Gs u)(x0 ) – (Gs u)(x)| ≤

ε ε + =ε 2 2

(2.23)


Page 48

if |x0 – x| ≤ δ . This is indeed the definition of continuity of any function, which was to be proved. The last integral was estimated over a plane small disk of the radius r ε rather than the corresponding nonplane small area Sε , which is the same, provided that the small quantities of the order of O (ε2 ) are neglected. Note that the presented proof is valid if x lies on a smooth part of the surface S . If x hits an edge or any other sharp ledge of the piecewise boundary the proof is quite similar, with some minor modifications. DEFINITION. Let ny (y ∈ ∂D ) be an outward unit normal. Then the integral Z ∂Φ(|x – y |) u(y ) dsy (2.24) (Gd u)(x) = ∂ny ∂D

is called the double-layer potential . For the same reason as for the single-layer potential this integral is evidently continuous outside of a neighborhood of the boundary ∂D . The following theorem states that it is discontinuous when passing across the boundary ∂D . THEOREM 2. If u(y ) ∈ C (∂D ), then at x → y0 ∈ ∂D the double-layer potential (2.24) can be continuously extended both from outside of the boundary ∂D and from inside, with their limit values being, respectively, Z u(y0 ) ∂Φ(|y0 – y |) . (2.25) u(y ) dsy ± (Gd u)± (y0 ) = ∂ny 2 ∂D

Proof. To be more specific, we will demonstrate the proof for the case of a 2D problem with a boundary contour l = ∂D , and the limit x → y0 ∈ l from outside of the contour. The other cases can be proved likewise. Let us represent the integral (2.24) as follows: Z ∂Φ(|x – y |) ∂Φ0 (|x – y |) – u(y ) dly (Gd u)(x) = ∂ny ∂ny l Z Z (2.26) ∂Φ0 (|x – y |) ∂Φ0 (|x – y |) dly , + [u(y ) – u(x)] dly + u(x) ∂ny ∂ny l l

where Φ0 = Φ|k=0 = –(1/2π ) ln |x–y | is the low-frequency Green’s function, so that (Φ – Φ 0 ) ∈ C1 (R2 ) (cf. Eq. (2.16)). It can easily be shown that the first two integrals here are continuous in R2 , and major effort should be directed at the third one. Let us take the two functions u1 (y ) ≡ 1,

u2 (y ) = Φ0 (|x – y |).

(2.27)

Obviously, Φ0 (|x – y |) as a function of the argument y satisfies the Laplace equation ∆y Φ0 (|x – y |) = 0 throughout R2 , except a small neighborhood of the point x. Therefore, if x lies in the exterior of l, then both functions u1 (y ) and u2 (y ) satisfy the Laplace equation inside l. Then, by using Green’s formula, we can write out successively: ZZ 0= u1 (y )∆y u2 (y ) – u2 (y )∆y u1 (y ) dy1 dy2 Z Z D ∂u1 (y ) ∂Φ0 (|x – y |) ∂u2 (y ) – u2 (y ) dly = dly = u1 (y ) (2.28) ∂ny ∂ny ∂ny l l Z ∂Φ0 (|x – y |) dly = 0, x ∈ R2 \ D , ∼ ∂n y l


Page 49

hence the outer limit here is: lim

x→y0 ∈l

Z

l

∂Φ0 (|x – y |) dly = 0. ∂ny

(2.29)

We thus can evaluate this limit value of the full expression (2.26) as follows: Z Z ∂Φ(|y0 – y |) ∂Φ(|y0 – y |) ∂Φ0 (|y0 – y |) – u(y ) dly + u(y ) dly lim (Gd u)(x) = x→y0 ∈l ∂ny ∂ny ∂ny l Z Zl ∂Φ(|y0 – y |) ∂Φ(|y0 – y |) u(y ) dly – u(y0 ) dly . (2.30) = ∂ny ∂ny l l

In order to calculate the last integral, let us operate again with Green’s formula for the e contained same two functions u1 (y ) ≡ 1 and Φ0 (|y0 – y |) and apply it over the domain D between the contour l and a semi-circle Γ = Γ1 ∪ Γ2 of a large radius R whose diameter Γ1 is tangent to l at the point y0 , so that y0 is the center of the semi-circle (see Fig. 2.1).

l

G2 ~ D

G2

G1 y0

G1

Figure 2.1. Contour Γ = Γ1 ∪ Γ2 is tangent to the boundary line l

It can easily be proved that, despite a singularity of the function Φ 0 (|y0 – y |), considered as a function of the argument y , near the point y0 , the application of Green’s formula is correct for two reasons. Firstly, let us show that the integral ZZ I= ∆y Φ0 (|y0 – y |) dy1 dy2 (2.31) Dε

over a small neighborhood of this singular point of radius ε > 0 is small as ε → 0, and then e is finite. Indeed, if the contour l has a certain radius of the integral over the domain D curvature ρ at the point y0 , then simple observations, with the help of some classical results from differential geometry, show that the part of this ε-neighborhood belonging to the e has the square ε3/(6ρ2 ) + O (ε5 ). But the singularity of the quantity ∆y Φ0 (|y0 – y |) domain D is of the order of O (1/ε2 ), so the integral over this small domain Dε is of the order of O (ε), which is evidently small. Secondly, let us prove that ∂Φ0 (|y0 – y |)/∂ny = O (1) in a small neighborhood of y0 , both over Γ1 and l. Indeed, ∂Φ0 (|y0 – y |) = ny , grady Φ0 , y = (y1 , y2 ), y0 = (y01 , y02 ), ∂ny y 1 – y 01 y 2 – y 02 1 (2.32) , , hence grady Φ0 (|y0 – y |) = – 2π |y0 – y | |y0 – y | |y0 – y | (ny , y – y0 ) , y , y0 ∈ l , Γ1 . ny , grady Φ0 = – 2π |y0 – y |2


Page 50

It is clear that if y , y0 ∈ Γ1 , then (ny , y – y0) = 0, because these two vectors are orthogonal to each other. If y , y0 ∈ l, then |y – y0 | 2 + O |y – y0 | , (ny , y – y0 ) = |y – y0 | cos(ny , y – y0 ) = –|y – y0 | 2ρ

(2.33)

where ρ is the curvature radius of the contour l at the point y 0 . Consequently, 1 ∂Φ0 (|y0 – y |) =– + O (|y – y0 |), ∂ny 4πρ

y , y0 ∈ l ,

(2.34)

i.e., both integrals along the boundary lines l and Γ are finite, since their integrands are continuous and bounded. e gives Now, the application of Green’s formula over the domain D Z

Z

∂Φ0 (|y0 – y |) dly = 0 ∼ – ∂ny Γ l Z Z ∂Φ0 (|y0 – y |) dly . = + ∂ny Γ1 Γ2

Z

l

∂Φ0 (|y0 – y |) dly ∂ny

(2.35)

The first integral over Γ1 here is trivial, since ny ⊥(y0 – y ). The second integral is Z

Γ2

1 ∂Φ0 (|y0 – y |) dly = – ∂ny 2π

Z

π 0

∂ 1 (ln R) R dϕ = – . ∂R 2

(2.36)

It finally follows from Eqs. (2.30), (2.36) that lim (Gd u) (x) =

x→y0 ∈l

Z

l

u(y0 ) ∂Φ(|y0 – y |) u(y ) dly + . ∂ny 2

(2.37)

At the concluding part of the present section we cite without proof two results related to the normal derivatives of the single and double layer potentials (see, for example, Colton and Kress, 1983). THEOREM 3. If u(y ) ∈ C (∂D ), then at x → y0 ∈ ∂D the limit values of the normal derivative of single-layer potential are determined as follows:

∂ (Gs u)(x) ∂nx

±

x=y0

=

Z

∂D

∂Φ(|y0 – y |) u(y0 ) u(y ) dsy ∓ . ∂ny0 2

(2.38)

THEOREM 4. If u(y ) ∈ C (∂D ), then at x → y0 ∈ ∂D the normal derivative of the double-layer potential represents a function, which is continuous in R 3 (R2 ):

∂ (Gs u)(x) ∂nx

+

x=y0

=

∂ (Gs u)(x) ∂nx

–

.

(2.39)

x=y0


Page 51

Helpful remarks 1◦ . Interestingly, Theorem 2 asserts that the double-layer potential (2.24) is continuous up to the boundary line l, both from outside and inside. The key feature of the two limit boundary values (Gd u)± is that they are different, and the jump at the boundary line is equal to the value of the density at the boundary point: u(y 0 ). In this sense, properties of the double-layer potential are congeneric with those of the Cauchy-type singular integral. 2◦ . Theorem 2 was proved above for the case when y0 ∈ l lies on the smooth portion of the piecewise smooth boundary line. If y0 comes to a break of smoothness, of the inner angle α, then the term u(y0 )/2 in Theorem 2 must be changed by αu(y0 )/(2π ). Theorem 1 also remains valid in this case. 3◦ . The kernel of the integral operator (2.25) is piecewise continuous for piecewise smooth contour, since due to Eq. (2.34) we can observe that ∂Φ0 (|y0 – y |) 1 ∂Φ(|y0 – y |) ∼ ∼ – + O (|y – y0 |), ∂ny ∂ny 4πρ

y → y0

(y , y0 ∈ l). (2.40)

4◦ . A more detailed and more formal treatment of the modern potential theory, with its application to diffraction problems, can be found in Colton and Kress (1983). 5◦ . Physicists prefer an alternative definition of a Green’s function, as a solution to the equation ∆y Φ + k 2 Φ = –δ (y – x), (2.41) where the presence of Dirac’s delta function means that a point source is applied at the point x. However, it can be directly tested, by using properties of the Fourier transform (1.3) and the delta function (1.62), that such an approach leads to the same representations (2.11) and (2.14) with r = |x – y |.

2.2. Basic Integral Equations of the Diffraction Theory Let us study the exterior boundary value problem for the Helmholtz equation in the scalar case: ω (2.42) ∆u + k 2 u = 0, k= , c which holds for acoustic pressure u(y ) outside a simple closed piecewise smooth boundary of finite measure (see Fig. 2.2).

n

le x

l n D

Figure 2.2. Boundary contour l and a small ε-neighborhood of the observation point x


Page 52

THEOREM. Let x ∈ Rn \ D (n = 2, 3). If u(y ) is a solution of Eq. (2.42) satisfying the radiation condition, then the following Kirchhoff–Helmholtz integral formula is valid: Z ∂u(y ) ∂Φ(|x – y |) – Φ(|x – y |) dsy . (2.43) u(x) = u(y ) ∂ny ∂ny ∂D

Here ny is an outer normal to the domain D . Proof. To be more concrete, we give again a proof in the 2D case, because the 3D case is studied in very much the same way. e contained inside a disk DR of infinitely large radius R, Let us consider a domain D with the region D and an ε-neighborhood Dε of the point x removed (see Fig. 2.2). Then both functions u(y ) and Φ(|x – y |), the latter treated as a function of the argument y , satisfy e . Therefore, Green’s formula applied to this pair of functions the Helmholtz equation in D leads to the following succession of relations: ZZ 0= u(y )∆y Φ(|x – y |) – Φ(|x – y |)∆y u(y ) dy1 dy2 = e Z D (2.44) ∂u(y ) ∂Φ(|x – y |) e – Φ(|x – y |) dly , ∂ D = l ∪ lR ∪ lε , = u(y ) ∂ny ∂ny e ∂D e. where ny points outwards of D First of all, we will prove that, since both functions u(y ) and Φ(|x – y |) satisfy the radiation condition, the integral along a far-zone circle l R vanishes as R → ∞. Indeed, all functions satisfying the radiation condition in unbounded acoustic medium have far-field asymptotics of the following form:

eikR eikR Φ(|x – y |) = Φ∗ (x, ϕ) √ , R → ∞, (2.45) u(y ) ∼ u0 (ϕ) √ , R R where x = (R cos ϕ, R sin ϕ) are polar coordinates of the point x. This fact can be strictly proved for an arbitrary function satisfying the radiation condition, but we demonstrate it here only for the Green’s function. In the Cartesian coordinate system, let x = (x1 , x2 ) and y = (y1 , y2 ). Then p p |x – y | = (y1 – x1 )2 + (y2 – x2 )2 = |y |2 – 2(y , x) + |x|2 1 1 (y , x) (2.46) +O = R – |x| cos(α – ϕ) + O , = |y | – |y | |y | |y | x = (|x| cos α, |x| sin α), |y | → ∞.

Now let us recall the asymptotic behavior of the Green’s function Φ(|x – y |) = (i/4) × – y |) at infinity (see Eq. (2.13)): r 1 2 (1) i(k|x–y|–π/4) e +O , |y | = R → ∞. (2.47) H0 (k |x – y |) ∼ πkR R3/2

H0(1) (k |x

So, with the help of the far-field asymptotics (2.46), we have r 2 ikR i[k|x| cos(α–ϕ)+π/4] (1) e e H0 (k |x – y |) ∼ πkR 1 eikR ∗ , |y | = R → ∞, = Φ (x, ϕ) √ + O R3/2 R

(2.48)

for fixed x.


Page 53

Let us in the meantime come back to the integral along l R , where ∂/∂ny = ∂/∂R if y ∈ lR . It can trivially be seen from Eq. (2.45) that ∂u(y ) 1 ∂Φ(|x – y |) – Φ(|x – y |) =O , R → ∞, hence u(y ) ∂ny ∂ny R3/2 (2.49) Z ∂u(y ) 1 1 ∂Φ(|x – y |) – Φ(|x – y |) dly = 2πR O =O . u(y ) ∂ny ∂ny R2 R lR The next step is to estimate a similar integral over l ε (see Fig. 2.2). We have for ε → +0 (i.e., y → x): Z ∂u(y ) ∂Φ(|x – y |) – Φ(|x – y |) dly u(y ) ∂ny ∂ny lε Z Z (2.50) ∂u(y ) ∂Φ(|x – y |) dly – Φ(|x – y |) dly . = u(x) ∂ny ∂ny lε lε The last integrand here has a weak singularity only, which permits the estimate (see the small-argument asymptotics of the Green’s function (2.16)) Z Φ(|x – y |) ∂u(y ) dly ≤ max ∂u(y ) 1 ln ε 2πε → 0, ε → +0, (2.51) lε ∂ny ∂ny 2π lε since the function u(y ) is analytic outside l. The first integral on the right-hand side of Eq. (2.50) with ε → +0 behaves as Z 2π Z 1 ∂ (ln r ) ∂Φ(|x – y |) dly = 2πr dψ = 1, lim ε→+0 l ∂n 2 π ∂r y 0 ε

where we have introduced the polar coordinate system (r , ψ ) with origin at x. Finally, Eq. (2.44) yields Z ∂u(y ) ∂Φ(|x – y |) – Φ(|x – y |) dly u(y ) ∂ny ∂ny l Z ∂u(y ) ∂Φ(|x – y |) – Φ(|x – y |) dly = –u(x). =– u(y ) ∂ny ∂ny lε

(2.52)

(2.53)

If we change simultaneously the direction of the normal n y and the sign of the integral in (2.53), we obtain exactly what was to be proved. The Kirchhoff–Helmholtz integral formula (2.43) implies that an arbitrary scalar wave field, satisfying the radiation condition, can be represented outside the boundary ∂D as a combination of single and double layer potentials with some densities distributed on the boundary ∂D . Thus, if we knew the values of both u and ∂u/∂n on the boundary, then we would be able to directly determine the unknown wave function at an arbitrary point x from Eq. (2.43). Unfortunately, these two quantities are never known simultaneously. Actually, if we study a boundary value problem of the Dirichlet type, then the unknown function u(y ) itself is given on the boundary. If we consider a boundary value problem of the Neumann type, then the value of the normal derivative ∂u/∂n is known on the boundary. In this sense, the Kirchhoff–Helmholtz formula does not give any solution to the boundary value problem. However, the result of the proved theorem, in conjunction with the properties of the potentials of the single and double layers, permits immediately the derivation of the basic boundary integral equation (BIE) of diffraction theory within the framework of the so-called


Page 54

direct BIE method. Let us move the point x to the boundary, x → y 0 ∈ ∂D , from outside. Then, according to Eq. (2.43) and the boundary values of the potentials of the single and double layers from the previous section, we have Z ∂u(y ) u(y0 ) ∂Φ(|y0 – y |) – Φ(|y0 – y |) dsy + u(y0 ) = u(y ) ∂n ∂n 2 y y ∂D (2.54) Z ∂u(y ) ∂Φ(|y0 – y |) – Φ(|y0 – y |) dsy , y0 ∈ ∂D . ∼ u(y0 ) = 2 u(y ) ∂ny ∂ny ∂D Below we give a formulation of a problem that covers the principal subject of the present chapter.

Formulation of diffraction problem Let a given incident monochromatic acoustic wave with angular frequency ω and (complexvalued) amplitude of acoustic pressure pinc (x) fall on an obstacle occupying a domain D (see Fig. 2.3). When the wave encounters this obstacle, the wave structure begins to

x2 ny

y

S

p inc x1 D x3 Figure 2.3. Incident acoustic wave and obstacle D with the boundary surface S change. Any change in the incident wave due to its interaction with the obstacle is called diffraction. Synonyms to diffraction are scattering and (in the short-wave regime) reflection. A diffracted wave is marked in the present book with the superscript sc. Thus, due to the presence of an obstacle and the arising of a scattered wave, the structure of the full wave field, p = pinc + psc , is different from that of the incident wave. All three introduced wave pressures satisfy the Helmholtz equation (2.42), but only p sc of them satisfies the radiation condition. The incident wave comes from the far-zone and hence can be assumed to arrive from infinity. Therefore, the radiation condition is broken for full pressure also, which is the sum of the pressures of the incident and the scattered waves. Hence, the derived Kirchhoff–Helmholtz integral representation (2.43) is valid only for the scattered wave, Z ∂psc (y ) ∂Φ(|x – y |) sc sc – Φ(|x – y |) dsy , x ∈ Rn \D (n = 2, 3), (2.55) p (x) = p (y ) ∂n ∂n y y ∂D and so is the (BIE) boundary equation (2.54), Z ∂psc (y ) ∂Φ(|y0 – y |) sc sc – Φ(|y0 – y |) dsy , p (y0 ) = 2 p (y ) ∂ny ∂ny ∂D

y0 ∈ ∂D .

(2.56)


Page 55

Now the two different natural types of boundary conditions are physically related to the cases of acoustically hard and acoustically soft boundary of the obstacle. In the first case (recall that vn = (iρ0 ω )–1 (∂p/∂n) if the wave process is harmonic in time) ∂psc ∂pinc ∂p = 0 ∼ = – , (2.57) vn |∂D = 0 ∼ ∂n ∂D ∂n ∂D ∂n ∂D

and in the second case

p |∂D = 0

∼

psc |∂D = – pinc |∂D .

(2.58)

In both cases we arrive at a BIE by substituting the respective boundary condition (2.57) or (2.58) into Eq. (2.56). The only inconvenience will be that in both cases the right-hand side is represented by quadrature. The following slight modification of the Kirchhoff–Helmholtz treatment allows us to avoid this inconvenience. To this end, let us come back once again to our standard operation with Green’s formula applied to a pair of functions u1 (y ) = pinc (y ) and u2 (y ) = Φ(|x – y |), with both of them being analytic and satisfying the Helmholtz equation inside the domain D if x lies outside D . Then the application of Green’s formula just over D proves that Z inc 0= p (y )∆y Φ(|x – y |) – Φ(|x – y |)∆y pinc (y ) dVy D

∂pinc (y ) ∂Φ(|x – y |) – Φ(|x – y |) dsy , =– p (y ) ∂ny ∂ny ∂D x ∈ Rn \ D (n = 2, 3). Z

inc

(2.59)

By adding (2.59) and (2.56) together, we obtain a more advanced representation for the scattered wave field in terms of the total wave pressure: Z ∂p(y ) ∂Φ(|x – y |) – Φ(|x – y |) dsy , x ∈ Rn \ D (n = 2, 3). (2.60) psc (x) = p(y ) ∂n ∂n y y ∂D Now, using the boundary limit values of single and double layer potentials and by letting x → y0 ∈ ∂D , we arrive at an alternative boundary representation, instead of Eq. (2.56): Z ∂p(y ) p(y0 ) ∂Φ(|y0 – y |) sc – Φ(|y0 – y |) dsy + , y0 ∈ ∂D . (2.61) p (y0 ) = p(y ) ∂ny ∂ny 2 ∂D It becomes clear from Eqs. (2.57), (2.58), (2.61) that the diffraction problem for an acoustically hard obstacle (Neumann-type boundary value problem) is reduced to a Fredholm integral equation of the second kind Z ∂Φ(|y0 – y |) p(y ) dsy = 2pinc (y0 ), y0 ∈ ∂D , (2.62) p(y0 ) – 2 ∂ny ∂D since psc = p – pinc . For the same reason, the acoustically soft case (Dirichlet-type boundary value problem) can be reduced to a Fredholm equation of the first kind Z ∂p(y ) dsy = pinc (y0 ), y0 ∈ ∂D . (2.63) Φ(|y0 – y |) ∂n y ∂D


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Helpful remarks 1◦ . Let us briefly outline qualitative properties of equations (2.62) and (2.63). As remarked above, the first of them is a Fredholm integral equation of the second kind. As stated in Eq. (2.34), for a piecewise smooth boundary the kernel is piecewise continuous, so the integral operator in Eq. (2.62) is compact in the space C (∂D ) and equation (2.62) can be studied within the framework of the Fredholm theory (Section 1.5). This implies the application of both theoretical results and well-developed numerical methods. The integral operator (2.63) is of the first kind, and it is obvious that it possesses a weak singularity. Hence, it is again of the Fredholm type, and equation (2.63) is a Fredholm integral equation of the first kind with operator compact in C (∂D ). 2◦ . Equation (2.62) is valid if the point y0 belongs to the smooth part of the boundary. If y0 hits a sharp angle of the value α, then the coefficient 2 in front of the integral operator should be changed by one expressed through α, in accordance with the described properties of the double-layer potential (see the previous section). 3◦ . Since only the second-kind Fredholm integral equations are described by so harmonious Fredholm theory, much effort was directed to the formulation of the Dirichlet boundary value problem in terms of the second-kind Fredholm BIE. Fortunately, this can be achieved by applying the so-called indirect BIE method, which is as follows. The function Z ∂Φ(|x – y |) ψ (y ) dsy , x ∈ Rn \ D (n = 2, 3), (2.64) psc (x) = ∂n y ∂D satisfies the Helmholtz equation and the radiation condition with any continuous density ψ . If in the case of acoustically soft obstacle the point x approaches the boundary ∂D , x → y0 ∈ ∂D , then by using the properties of the double-layer potential and the boundary condition (2.58), we arrive at the following equation: Z ∂Φ(|y0 – y |) ψ (y ) dsy = –2pinc (y0 ), y0 ∈ ∂D , (2.65) ψ (y0 ) + 2 ∂n y ∂D which is an equation of the second kind.

2.3. Properties of Integral Operators of Diffraction Theory: General Case and Low Frequencies Qualitative properties of the integral equations obtained are closely connected with general properties of the exterior Dirichlet and Neumann boundary value problems established by classicists (see, for example, Weyl, 1952; Kupradze, 1950; Atkinson, 1949), which we quote without proof for the sake of brevity. THEOREM 1. Both the Dirichlet and Neumann exterior boundary value problems for the Helmholtz equation have a unique solution in the class of functions satisfying the radiation condition. THEOREM 2. The interior Dirichlet and Neumann boundary value problems for the Helmholtz equation have a unique solution for all values of the wave number k , except a countable set of positive values of k (related to eigen, or resonance frequencies). From this classical result we can extract some important properties of the considered integral operators.


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THEOREM 3. The second-kind integral equation (2.65) of the indirect BIE method for the Dirichlet exterior boundary value problem has a unique solution for any k , except those corresponding to eigenvalues of the interior Neumann problem. Proof. According to the Fredholm theory for integral equations of the second kind, equation (2.65) is surely uniquely solvable if the homogeneous equation (2.65) has only the trivial solution. So, let the parameter k be such that there is a nontrivial solution ψ of the homogeneous equation (2.65), i.e., Z ∂Φ(|y0 – y |) ψ (y0 ) + ψ (y ) dsy = 0, y0 ∈ ∂D . (2.66) 2 ∂ny ∂D Let us introduce the double-layer potential with this density Z ∂Φ(|x – y |) ψ (y ) dsy . u(x) = ∂ny ∂D

(2.67)

Then, due to the limit boundary values of the double-layer potential, we have (the subscript “+” refers to the outer limit) Z ∂Φ(|y0 – y |) ψ (y0 ) + ψ (y ) dsy = 0. (2.68) u+ (y0 ) = 2 ∂ny ∂D According to Theorem 1, u(x) ≡ 0 outside D , as a solution of the exterior Dirichlet problem with the trivial boundary condition. This immediately implies ∂u + (y0 )/∂n = 0, y0 ∈ ∂D . Further, the continuity of the normal derivative of the double-layer potential (Theorem 4, Section 2.1) proves that ∂u– (y0 )/∂n = 0, y0 ∈ ∂D . We can thus finally see that a nontrivial solution of the exterior Dirichlet problem generates a nontrivial (i.e., eigen) solution of the interior Neumann problem. A similar technique can be applied to prove a similar theorem for the exterior Neumann problem. THEOREM 4. The second-kind integral equation (2.62) of direct BIE for the Neumann exterior boundary value problem has a unique solution for any k , except eigenvalues of the interior Dirichlet problem. It is somewhat unexpected that a similar result is valid for the first-kind BIE (2.63). A proof to the following theorem can be found, for example, in Colton and Kress (1983). THEOREM 5. The Fredholm integral equation of the first kind (2.63) with a kernel containing a weak singularity is uniquely solvable for all values of the parameter k , except those corresponding to eigenvalues of the interior Dirichlet problem.

Low-frequency diffraction problem From the general variational principles it follows that the countable set of eigenvalues for both types of interior boundary value problems has a limit point at infinity (see Courant and Hilbert, 1953; and our discussion in Chapter 4). Therefore, there is always a minimum eigenfrequency. Then it directly follows from Theorems 3–5 that the considered integral equations (2.62), (2.63), (2.65) have a unique solution if the frequency is less than the first eigenvalue of the corresponding interior problem. Thus, for sufficiently low frequencies we may directly treat these equations, both analytically and numerically.


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Example. Low-frequency diffraction by a hard round disk. We consider this example as a 2D diffraction problem for acoustically hard obstacle. Let a be a radius of the disk and let the incident wave be plane propagating in direction of the x1 axis. Then we have 1 i i (1) H0 (kr ) ∼ – [ ln(kr ) + γ – ln 2 ] + + O k 2 ln k , k → 0; hence 4 2π 4 (2.69) ∂Φ ∂r 1 ∂r cos(r , ny ) ∂Φ inc ikx1 = =– =– , r = |y0 – y |, p (x) = e . ∂ny ∂r ∂ny 2πr ∂ny 2πr

Φ=

Representation in the polar coordinate system implies y = {a cos θ , a sin θ }, y0 = {a cos ψ , a sin ψ }, r = y – y0 = {a (cos θ – cos ψ ), a(sin θ – sin ψ )}, ny = {cos θ , sin θ }, cos θ (cos θ – cos ψ ) + sin θ (sin θ – sin ψ ) (r ⋅ ny ) =a , cos(r , ny ) = r r

(2.70)

hence 1 1 – cos(θ – ψ ) 1 1 – cos(θ – ψ ) 1 ∂Φ =– =– =– . (2.71) 2 2 ∂ny 2πa (cos θ – cos ψ ) + (sin θ – sin ψ ) 2πa 2[1 – cos(θ – ψ )] 4πa This yields the following integral equation of low-frequency diffraction for a hard round disk, which follows from Eq. (2.62): Z 2π 1 p(θ ) dθ = 2pinc (ψ ) (dly = a dθ ) p(ψ ) + 2π 0 (2.72) Z 2π P ika cos ψ = 2e , P = p(θ ) dθ , ∼ p(ψ ) + 2π 0 where P is some constant. Now, by integrating both sides of the last equation over the interval [0, 2π ], we obtain (J0 is the Bessel function) Z 2π Z 2π p(ψ ) dψ + P = 2 eika cos ψ dψ ∼ P = 2π J0 (ak ), (2.73) 0

0

where we have used the value of the tabulated integral arisen in Eq. (2.73). Now, combining Eqs. (2.72) and (2.73) together, we arrive at the following solution of the basic integral equation in the case of low frequencies: p(ψ ) = 2 eika cos ψ – J0 (ak ) ≈ 1 + 2ika cos ψ ,

k → 0.

(2.74)

Note that the highest order of error for this solution was caused by approximation of the kernel in Eq. (2.69), so the real error of expression (2.74) that gives the solution of BIE is O (k 2 ln k ). A plane incident wave has been taken in the above example as the most standard type of incident waves used in theoretical study and in practice. Perhaps, it is so widespread in diffraction theory due to the fact that any wave can locally be treated as plane. Another reason is that all other simple types of waves (cylindrical, spherical, etc.) can be, as a rule, represented as a superposition of plane waves (see, for example, Brekhovskikh, 1980).


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Arbitrary 3D obstacle: Acoustically hard boundary Let us derive a general low-frequency form of an integral equation in the 3D case up to the linear (with respect to the small parameter k ) small terms. To be more specific, we will consider a plane wave incident on an obstacle D with acoustically hard boundary, i.e., we will treat Eq. (2.62). If the direction of propagation of the plane coincides, for instance, with the axis x3 , then we have pinc (x) = exp(iky0 3 ) in Eq. (2.62), or pinc (x) = 1+ iky0 3 + O (k 2 ) for small k . Let us write out the first asymptotic terms of the normal derivative in the kernel of Eq. (2.62): ikr ∂ e ∂ 1 + ikr ∂Φ(|y0 – y |) = cos(ny , r ) = cos(ny , r ) + O (k 2 ) ∂ny ∂r 4πr ∂ny 4πr ∂Φ0 (|y0 – y |) 1 (2.75) cos(ny , r ) + O (k 2 ) = + O (k 2 ), where =– 4πr 2 ∂ny ∂Φ0 (|y0 – y |) 1 1 , =– cos(ny , r ), r = |y0 – y |, Φ0 (|y0 – y |) = 4πr ∂ny 4πr 2 so, for small frequencies, with an O (k 2 ) error, Eq. (2.62) becomes Z ∂Φ0 (|y0 – y |) p(y ) dsy = 2(1 + iky0 3 ), p(y0 ) – 2 ∂ny S

y0 ∈ S .

(2.76)

It is very interesting that within the framework of this approximation the kernel is independent of the parameter k . Moreover, it coincides with the limit (k → 0) lowfrequency value. Hence, the first two asymptotic terms of a low-frequency solution to the considered integral equation may be obtained by solving Eq. (2.76) just for the two respective terms of the right-hand side, so that p(y ) = p 0 (y ) + ikp1 (y ) + O (k 2 ). Let us prove that for arbitrary smooth shape of the obstacle boundary surface S the leading asymptotic term, which is determined as a solution of the equation Z ∂Φ0 (|y0 – y |) p0 (y ) dsy = 2, y0 ∈ S , (2.77) p0 (y0 ) – 2 ∂ny S

is a certain constant p0 (y ) ≡ p0 . In order to prove this statement, let us evaluate how the integral operator in Eq. (2.77) acts on a constant. To this end, let us take, following our standard technique, the two functions u1 (y ) ≡ p0 and u2 (y ) = Φ0 (|x – y |), with both of them being analytic inside D when x is located outside D . Then Green’s formula, applied over a domain D , yields (cf. Eq. (2.28)) Z 0= u1 (y )∆y u2 (y ) – u2 (y )∆y u1 (y ) dy1 dy2 ZD (2.78) ∂Φ0 (|x – y |) 3 p0 dsy , x ∈ R \ D . = ∂ny S Now, by applying the outer limit, x → y0 ∈ S , to both sides of Eq. (13), we have Z Z ∂Φ0 (|x – y |) ∂Φ0 (|x – y |) p0 p0 + p0 dsy = 0 ∼ p0 dsy = – , 2 ∂ny ∂ny 2 S S

(2.79)

so that Eq. (2.77) becomes

2 p0 = 2

∼

p0 (y ) ≡ p0 = 1.

(2.80)


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The obtained result is quite natural from the intuitive (heuristic) point of view. Indeed, the low-frequency limit is equivalent to the case of a very small obstacle with a fixed frequency. But if the obstacle is infinitely small, then the incident wave does not feel it, and so the total pressure remains without any perturbation, i.e., p 0 |S = pinc |k=0 = 1. Unfortunately, the next asymptotic term, related to the second term on the right-hand side of Eq. (2.76), cannot be constructed for arbitrary shape, but it admits explicit treatment for some canonical shapes. The case of a spherical obstacle is considered in the next section.

Helpful remarks 1◦ . Theorems 3–5 state that the considered integral equations are uniquely solvable for all k excluding a countable set of eigenvalues of the corresponding interior boundary value problem. The results stated in the final part of Section 1.5 guarantee that, at least for second-kind equations, you may easily construct a stable numerical solution for arbitrary boundary of noncanonical shape in this case. A great deal of papers were devoted to proposing a theory of boundary integral equations that would be correctly solvable for all k , since exterior problems are always uniquely solvable, in contrast to their integral equations (see, for example, Jones, 1974; Kleinman and Roach, 1982). In a formal theory these results are perhaps very important; however they are less important from a practical aspect. If you create your own computer code in any algorithmic language, without regard for the value of the parameter k , then your algorithm written for the exterior problem must formally crash for the described set of eigenfrequencies, However, in practice this never happens so. It is not so easy to compel the algorithm to crash with a particular value of k . In practice, you will not feel that the algorithm behaves distinctively from regular cases until you set an “irregular” value of k with an accuracy of 10 –6 –10–7 . So, if you take any k with three or four significant digits, you may use your code for arbitrary k without any problem. 2◦ . Analogous observations take place when you operate with a piecewise smooth boundary surface (3D case) or contour (2D case). All derived equations of the second kind, as they are stated, are valid whenever y0 belongs to the smooth portion of the boundary. If y0 is at a corner or cusp point, the factor in front of the unknown function outside the integral must be changed, and you should take this fact into account when writing your computer code. But we advise that you forget forever about this feature of the problem and choose instead such a grid that its nodes do not hit any sharp edge. Then you will not face any trouble in practice. The precision of your computations will be even higher than if you take a different coefficient only for one (“sharp”) node.

2.4. Full Low-Frequency Solution for Spherical Obstacle Acoustically hard obstacle The second asymptotic term of the solution to Eq. (2.76) ought to be found from the following equation: p1 (y0 ) – 2

Z

S

∂Φ0 (|y0 – y |) p1 (y ) dsy = 2y0 3 , ∂ny

y0 ∈ S .

(2.81)

Let us introduce a spherical coordinate system for the points y and y 0 on the surface of


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a sphere of radius a: y = a(sin θ cos ϕ, sin θ sin ϕ, cos θ ),

y0 = a(sin θ0 cos ϕ0 , sin θ0 sin ϕ0 , cos θ0 ),

r = |y0 – y | = 2a [1 – cos θ cos θ0 – sin θ sin θ0 cos(ϕ – ϕ0 )], a cos(ny , r ) = [1 – cos θ cos θ0 – sin θ sin θ0 cos(ϕ – ϕ0 )], r cos(ny , r ) 1 ∂Φ0 (|y0 – y |) =– =– . 2 ∂ny 4πr 8π a r 2

2

2

(2.82)

The problem in hand is axially symmetric. So we seek a solution that does not depend upon the angle ϕ. Then equation (2.81) in the spherical coordinate system becomes Z π 1 K0 (θ0 , θ ) p1 (θ ) dθ = 2a cos θ0 , 0 ≤ θ0 ≤ π , p1 (θ0 ) + √ 4π 2 0 Z 2π dϕ √ K0 (θ0 , θ ) = sin θ (2.83) 1 – cos θ cos θ0 – sin θ sin θ0 cos ϕ 0 √ √ sin θ sin θ0 2 2 sin θ K , = |sin[(θ + θ0 )/2]| |sin[(θ + θ0 )/2]| where K is the full elliptic integral of the first kind (Gradshteyn and Ryzhik, 1994). One can hardly believe that this equation admits exact analytical solution. Many mathematicians spend much effort to prove that an exact solution of this equation can be found in the form p1 (θ ) = B a cos θ = By3 , (2.84) where B is some constant. The discussed question demonstrates an excellent example of what is discussed in the Preface: it happens so frequently that heuristic ideas help to achieve a breakthrough in some problems of pure or applied mathematics that could not be resolved directly by strict formal methods. It was Lord Rayleigh who first discovered that a small sphere behaves like a dipole when irradiated by a plane wave (a good presentation can be found, for instance, in Morse and Feshbach, 1953; H o¨ nl et al., 1961). So, let us prove that the solution of Eq. (2.83), or, what is the same, of Eq. (2.81), has the form (2.84). For this purpose, we apply (over a domain D ) once again our standard approach using Green’s formula, with two functions u 2 (y ) = y3 and u2 (y ) = Φ0 (|x – y |) (the point x is fixed outside D ). Then Green’s formula gives Z 0= u1 (y )∆y u2 (y ) – u2 (y )∆y u1 (y ) dy1 dy2 ZD (2.85) ∂y3 ∂Φ0 (|x – y |) 3 – Φ0 (|x – y |) dsy , x ∈ R \ D . = y3 ∂ny ∂ny S Now, by letting x → y0 ∈ S in Eq. (2.85) and by using the boundary properties of the single and double layer potentials, we obtain Z ∂Φ0 (|y0 – y |) ∂y3 y0 3 + y3 – Φ0 (|y0 – y |) dsy = 0. (2.86) 2 ∂ny ∂ny S It is obvious that on a spherical surface of radius a, the following relations hold (cf.


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Eq. (2.82)):

y3 ∂y3 = cos θ = , ∂ny a cos(ny , r ) 1 ∂Φ0 (|y0 – y |) =– =– , (2.87) 2 ∂ny 4πr 8π a r ∂Φ0 (|y0 – y |) 1 = –2a , Φ0 (|y0 – y |) = 4π r ∂ny and consequently the integral in (2.86) becomes Z Z ∂Φ0 (|y0 – y |) ∂y3 ∂Φ0 (|y0 – y |) y0 3 = y3 – Φ0 (|y0 – y |) dsy = 3 y3 dsy . (2.88) – 2 ∂ny ∂ny ∂ny S S y3 = a cos θ ,

Hence, the integral operator in Eq. (2.81), in the case of a small spherical obstacle, acts on the function y3 as follows: Z y0 ∂Φ0 (|y0 – y |) y3 dsy = – 3 . (2.89) ∂n 6 y S It now becomes clear that an exact solution to Eq. (2.81), and also to that written in an equivalent form (2.83), can be obtained by substituting the potentially appropriate structure (2.84) to Eq. (2.81), which together with Eq. (2.89) leads to the relation

3 3 3 4 Bay3 = 2ay3 ∼ B = ∼ p1 = y3 = a cos θ , (2.90) 3 2 2 2 which proves the hypothesis about the structure of the function p 1 (θ ) written in the form (2.84). Finally, taking into account that p0 (θ ) ≡ 1 and also what was written in the paragraph right after Eq. (2.76), we arrive at the full two-term low-frequency asymptotics in the case of a spherical obstacle in the following form: p(θ ) = 1 +

3 ika cos θ . 2

(2.91)

Acoustically soft obstacle In this case the problem may be reduced to Eq. (2.63). The low-frequency expansion for the kernel is here represented as Φ(|y0 – y |) =

1 ik ik eikr = + + O (k 2 ) = Φ0 (|y0 – y |) + + O (k 2 ) 4πr 4πr 4π 4π

so, with an O (k 2 ) error, equation (2.63) is equivalent to Z ik g (y ) dsy = 1 + iky0 3 , y0 ∈ S , Φ0 (|y0 – y |) + 4π S

(r = |y0 – y |), (2.92)

g (y ) =

∂p(y ) . ∂ny

(2.93)

The second term in the kernel generates a certain constant; hence the solution of Eq. (2.93) admits the following decomposition: Z ikI , y0 ∈ S , Φ0 (|y0 – y |) g (y ) dsy = 1 + iky0 3 – 4π S Z Z (2.94) I= g (y ) dsy = g0 (y ) dsy + O (k ), g (y ) = g0 (y ) + ikg1 (y ), S

S


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where the functions g0 and g1 are independent of the wave number k and give, respectively, the leading (zeroth) and the first asymptotic terms of the solution. The first of them is a solution of Eq. (2.94) with the unit right-hand side, and the second with the right-hand side in the form y0 3 – I/(4π ). Let us show that a constant is still a solution to the first equation, and this is also true in the case of acoustically soft obstacle. To this end, let us study how operator (2.94) acts on the constant g0 (y ) ≡ g0 . The answer to this question can be given immediately in the case of spherical shape (see Eq. (2.87) and Eq. (2.79)): Z

Φ0 (|y0 – y |) g0 dsy = –2a S

Z

S

g ∂Φ0 (|y0 – y |) 0 g0 dsy = (–2a) – = ag0 , ∂ny 2

(2.95)

hence g0 (y ) ≡ g0 = 1/a. Further, if all terms of the order of O (k 2 ), including that arising for ikI in Eq. (2.94), are neglected, the first-term function g1 (y ) is found from the equation Z

I + y0 3 , where Φ0 (|y0 – y |) g1 (y ) dsy = – 4π S Z g0 dsy = 4πa2 g0 = 4πa, y0 ∈ S . I=

(2.96)

S

Of course, a solution of this equation can be constructed as the sum g 1 (y ) = h1 (y ) + h2 (y ), where the first of these new functions is responsible for the constant right-hand side in (2.96): h1 (y ) ≡ –I/(4πa) = –1, and the second one is to be defined from the equation Z

Φ0 (|y0 – y |) h2 (y ) dsy = y0 3 , S

y0 ∈ S .

(2.97)

Since Φ0 (|y0 – y |) = –2a ∂Φ0 (|y0 – y |)/∂ny for a spherical obstacle (see Eq. (2.87)) and due to relation (2.89), we conclude that h2 (y ) = 3y3 /a. At last, by combining together the expressions for the functions g 0 , h1 , h2 , we arrive at the following low-frequency asymptotic solution: 1 3 y3 1 + ika(3 cos θ – 1) ∂p(y ) = g (y ) = + ik –1 + = . ∂ny a a a

(2.98)

Helpful remarks 1◦ . Interestingly, the method discussed here allows us to construct a solution to the very complex integral equation (2.83) for two specific right-hand sides, which is not so easy to obtain by any alternative method. Very often general principles of mathematical physics (like Green’s theorem) allow one to derive explicit-form solutions of some complex differential and integral equations. 2◦ . Some other canonical shapes admit explicit-form solutions of the diffraction problem in the low-frequency range. Among others, we mention here diffraction by a half-plane, by ellipses, by a wedge with infinite faces, etc. A good survey of the known analytical solutions can be found in Bowman et al. (1987).


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2.5. Application: Scattering Diagram for Obstacles of Canonical Shape The basic integral representation of the solution to the diffraction problem given by the Kirchhoff–Helmholtz formula (2.60) permits the derivation of the so-called scattering diagram, or scattering pattern. The latter is defined as a function that represents the dependence of the real scattered amplitude upon the angle of observation. Let us consider again, as in formula (2.46), the far-field representation for the Green’s function, now in the 3D case and for large |x|: p r = |x – y | = (x1 – y1 )2 + (x2 – y2 )2 + (x3 – y3 )2 p 1 (x, y ) +O , = |x|2 – 2(x, y ) + |y |2 = |x| – |x| |x|

|x| → ∞.

(2.99)

Then in a far zone we have, in the asymptotic sense,

eik|x| –ik(x,y)/|x| eikr ≈ e , 4πr 4π |x| eik|x| –ik(x,y)/|x| ik ikr ∂Φ(|x – y |) e cos(ny , y – x) ≈ – e ik cos(ny , x), ≈ ∂ny 4πr 4π |x| Φ(|x – y |) =

(2.100)

so that the scattered amplitude given by Eq. (2.60) takes in the far zone the following form: psc (x) ∼ –

eik|x| 4π |x|

Z

p(y ) ik cos(ny , x) +

S

∂p(y ) –ik(x,y)/|x| e dsy , ∂ny

|x| → ∞.

(2.101)

Recall that for acoustically hard obstacle the second term in the square brackets vanishes, and so does the first term for acoustically soft body. It should also be noted that expression (2.101) confirms that any scattered wave field in a far zone behaves like a spherical wave. An analogous formula can be derived in the 2D case: (1 + i) eik|x| p (x) ∼ – √ 4 πk |x| sc

Z l

∂p(y ) –ik(x,y)/|x| p(y ) ik cos(ny , x) + e dly , |x| → ∞. (2.102) ∂ny

Both expressions are valid for arbitrary boundary shape and arbitrary value of the frequency parameter (wave number) k . However they admit explicit analytical treatment only for canonical shapes and low frequencies.

Low-frequency scattering from a hard round disk Recall that in this case ∂p(y )/∂ny in Eq. (2.102) vanishes and p(y ) is given in the polar coordinate system by Eq. (2.74). If the far point x is represented in the same polar system as x = (x1 = R cos α, x2 = R sin α), R = |x|, then by using the evident relations (ny , x) R(cos ψ cos α + sin ψ sin α) = = cos(ψ – α), |x| R (2.103) (x, y ) Ra(cos α cos ψ + sin α) sin ψ = = a cos(ψ – α), y = (a cos ψ , a sin ψ ) |x| R cos(ny , x) =


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and by omitting an inessential factor, finding the real-valued scattered amplitude can be reduced to the calculation of the following integral: Z k –ik(x,y)/|x| sc p(y ) cos(ny , x) e dly |p (α)| ∼ √ k |x| l Z ak 2π –iak cos(ψ–α) (2.104) (1 + 2iak cos ψ ) cos(ψ – α) e dψ = √ k |x| 0 (ak )2 |1 – 2 cos α| , ∼ √ k |x| where only the leading asymptotic term (with respect to k ) is given here. The last line in Eq. (2.104) shows that as the frequency decreases the scattered amplitude decreases like O (k 3/2 ). The dependence upon the observation angle α is demonstrated as a diagram in Fig. 2.4.

90° 1.5

r = |1 - 2 cos a|

1.0 0.5

a = 0°

180° -3.0

-2.5

-2.0

-1.5

-1.0

-0.5

0

0.5

1

-0.5 -1.0 -1.5

270° Figure 2.4. Scattering diagram for hard disk as a function of polar angle α

Low-frequency diagram for acoustically hard sphere Here we return again to the spherical coordinate system (see Eq. (2.82)) and substitute appropriate expressions into formula (2.101). Then, by setting the angles of observation (α, β ) for the far-field point x so that x = (R sin α cos β , R sin α sin β , R sin α),

R = |x|,

(2.105)

taking into account the evident relations ∂p(y ) 3 ika cos θ , = 0, 2 ∂ny (ny , x) (y , x) = = sin α sin θ cos(ϕ – β ) + cos α cos θ , cos(ny , x) = |x| |y | |x| (x, y ) = a [sin α sin θ cos(ϕ – β ) + cos α cos θ ] , |x|

p(y ) = p(θ ) = 1 +

(2.106)


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and omitting some inessential factor, we arrive at the leading asymptotic term in the form Z k –ik(x,y)/|x| sc p(y ) cos(ny , x) e dsy |p (α)| ∼ |x| S Z Z 3 ak 2π π 1 + iak cos θ [sin α sin θ cos(ϕ – β ) + cos α cos θ ] (2.107) = |x| 0 2 0 –iak[sin α sin θ cos(ϕ–β)+cos α cos θ] ×e sin θ dϕ dθ .

It is easily proved that the last integral does not depend on the polar angle β , and calculating the integral with respect to ϕ, we obtain the following dependence on the polar angle α: 3 (ak )2 sc 1 – cos α (2.108) |p (α)| ∼ |x| 2

which is a small quantity of the order of O (k 2 ) as the frequency decreases. It should be noted that the diagrams in Figs. 2.4, 2.5 are selected so that α = 0 refers to direct scattering and α = 180◦ to back scattering.

90° 1.0

r = |1 - 32 cos a| 0.5

a = 0°

180° -2.5

-2.0

-1.5

-1.0

-0.5

0

0.5

-0.5 -1.0

270° Figure 2.5. Scattering diagram for hard sphere as a function of polar angle α

Scattering by acoustically soft sphere In this case we should take into account that p(y ) = 0, ∂p(y )/∂n y = [1 + ika(3 cos θ – 1)]/a (see Eq. (2.98)) in Eq. (2.101). So we arrive at the following expression for the leading asymptotic term: Z 1 ∂p(y ) –ik(x,y)/|x| sc e dsy |p (α)| ∼ |x| S ∂ny Z Z 1 2π π –iak(sin α sin θ cos ϕ + cos α cos θ) [1 + ika(3 cos θ – 1)] e sin θ dϕ dθ (2.109) = |x| 0 0 a a |1 – ika| + O (k 2 ) = + O (k 2 ). ∼ |x| |x|


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Helpful remarks 1◦ . It is very important for applications that for acoustically hard obstacles there are always some angles of scattering where (in the asymptotic approximation) the object does not radiate any energy. This means that while staying in a far zone at these angles of observation, one cannot detect the presence of the object. For the 2D problem these angles are α = ± arccos(1/2) = ±60◦ , and in the 3D case α = ± arccos(2/3) ≈ ±48◦ ; see Figs. 2.4 and 2.5. By contrast, a small soft sphere radiates equal energy in all directions. 2◦ . A low-frequency amplitude of scattering for an acoustically hard obstacle seems to be credible, since in both the 2D and 3D cases it vanishes as the frequency decreases. However, in the case of an acoustically soft obstacle the scattered amplitude does not vanish for infinitely low k , which looks paradoxical. But the clue to this paradox is rather simple. The constructed scattered amplitude is obtained under two assumptions: 1) small k and 2) large distance |x|, which may under certain conditions contradict each other, so the result obtained is not uniform with respect to these two asymptotic parameters. Actually, in a far zone the only dimensionless parameter is k |x|, where the first factor is small and the second one is large, so we cannot a priori predict whether their product is small or large. Mathematically, this is equivalent to the alternative notation of Eq. (2.109) in the form |psc (α)| ∼

ak + O (k 2 ), |kx|

(2.110)

and it becomes now clear that the scattered amplitude decreases with decreasing dimensionless parameter ak . Physically, this paradox is connected with the fact that a smaller frequency implies a longer wave. As regards the far-field approximation, it should be remembered that this implies also that the distance |x| is much greater than the (long) wavelength. In the final part of the present discussion we would like to notice once again that many examples with applications to scattering by various canonical shapes can be found in Bowman et al. (1987); and in Felsen and Marcuvitz (1973).

2.6. Asymptotic Character of the Kirchhoff Physical Diffraction Theory As we already mentioned, the BIE method in the generic case of arbitrary obstacle’s shape can be solved by one or another numerical method. In regular cases (for instance, Fredholm equations of the second kind) a standard direct numerical treatment (see Section 1.5) is quite applicable. The application of numerical methods in irregular problems will be discussed in Chapter 9. We saw above that canonical geometries admit analytical analysis at low frequencies. Another extreme case where the problem can be studied analytically refers to high-frequency (or, short-wavelength) diffraction by convex obstacles. The foundations of the theory were created by Kirchhoff, and his theory is called the physical diffraction theory. We first give heuristic foundations of the theory and then prove its asymptotic character. There are a number of different formulations of the Kirchhoff theory (see H o¨ nl et al., 1961). The key points can be understood from Fig. 2.6. If there is a plane acoustic wave pinc (x) = eik(m⋅x) incident to a convex obstacle, then the boundary of the obstacle is naturally divided into “light” (l + ) and “shadow” (l – ) zones, which for simplicity are shown for a 2D problem in Fig. 2.6. The light zone l + is defined as one containing the boundary points with n ⋅ m < 0 and the shadow zone l – contains the boundary points with n ⋅ m > 0, where n is the outward normal to the obstacle contour and


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C

m y

l+ A

y

| y - y 0|

O y0 B

l

-

Figure 2.6. Incidence of a high-frequency acoustic wave onto a convex obstacle m is the unit vector defining the incident wave direction. The structure of the solution is absolutely different in the light and in the shadow. Asymptotically, as k → ∞, the total pressure p(x) = pinc (x) + psc (x) in the shadow zone is zero: p(x) = 0, x ∈ l – . The solution on l+ is more complex. The Kirchhoff idea was founded on the heuristic assumption that at high frequencies the surface becomes locally plane in the neighborhood of any point x0 ∈ l+ , and so p(x0 ) can be asymptotically obtained from the solution to the problem on the reflection of a plane incident wave pinc (x) from a plane tangent to the boundary contour at the point x0 . To be more specific, let us assume that the contour l is acoustically hard: (∂p/∂n)|l = 0, where p = pinc + psc is the total acoustic wave field. Then the solution to the problem on the reflection from an infinite plane reflector is easily constructed (see, for example, Brekhovskikh, 1980). It can be verified that the total wave field, ik sc ik(m⋅x) + eik[m–2(m⋅n)n]⋅x (2.111) p∞ (x) = pinc (x) + psc ∞ (x) = e (m ⋅ x) + p∞ (x) is p∞ (x) = e

if local Cartesian coordinate axes coincide with the (τ , n) directions (τ is the unit tangential vector and n is the unit outward normal), and the origin is placed at the point x 0 . We can see from Eq. (2.111) that p∞ (x0 ) = p∞ (0) = 2 = 2pinc (0) = 2pinc (x0 ),

(2.112)

and consequently the Kirchhoff approximation implies p(x0 ) ≈ 2pinc (x0 ),

k → ∞.

(2.113)

The Kirchhoff–Helmholtz integral representation (2.43) with the Kirchhoff approximation (2.113) now contains both functions p|l and (∂p/∂n)|l to be known, and the diffraction problem is now reduced to the calculation of some quadratures. For a long time there were numerous heated debates whether this theory is asymptotic at k → ∞ or not. Recently it was proved that the leading asymptotic term of the exact solution coincides with Kirchhoff’s prediction (see, for example, Taylor, 1981), but the proofs required too abstract mathematical instruments. We give here a simple and graphical proof. It is based upon the basic boundary integral equation (see Section 2.2). The only restriction is the established fact that the BIEs studied in this chapter are uniquely solvable for all k > 0, except a countable set of values {kn } (kn → ∞, n → ∞) corresponding to the eigenvalues of the corresponding interior boundary value problem. So, strictly speaking, the asymptotic property of Kirchhoff’s theory stated by the following theorem is valid only for regular high-frequency values of k . A more refined analysis should be carried out in order to prove that Kirchhoff’s solution is asymptotically valid also for irregular values of k (k → ∞), i.e., Kirchhoff’s approximation represents the leading asymptotic term of the solution for all large k . Assuming that p(x0 ) = o(1), k → ∞, x0 ∈ l– , we prove the following


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THEOREM. Let the boundary contour l be smooth, convex, and acoustically hard, and x0 ∈ l+ . Then p(x0 ) = 2pinc (x0 ) + o(1), k → ∞ (k ≠ kn ). (2.114)

Proof. If k ≠ kn , then the boundary integral equation in the case of the Neumann boundary condition (∂p/∂n) = 0, x → l, and with p(x 0 ) ≡ 0(1), k → ∞, is asymptotically equivalent to Z ∂Φ(|y0 – y |) dly = 2pinc (y0 ) + o(1), y0 → l + , (2.115) p(y0 ) – 2 p(y ) ∂ny l+

being uniquely solvable. Here Φ(r ) = (i/4)H0(1) (kr ), r = |y – y0 |, is the Green’s function for full 2D space. Moreover, for the considered k , the integral operator of the left-hand side, I – G, has a bonded continuous operator (I – G)–1 in the Banach space C (l). Let p(y ) = 2pinc (y ) + ϕ(y ), y ∈ l + , where ϕ(y ) is a new unknown function. Then equation (2.115) takes the form Z Z ∂Φ(|y0 – y |) ∂Φ(|y0 – y |) dly = 4 pinc dly + o(1), y0 ∈ l. (2.116) ϕ(y0 ) – 2 ϕ(y ) ∂n ∂ny + + y l l

Let us prove that Z ∂Φ(|y0 – y |) dly = O (1), lim pinc (y ) k→∞ l+ ∂ny

y0 ∈ l+ = AB ,

k ≠ kn .

(2.117)

Indeed, Z Z Z ∂Φ(|y0 – y |) ∂Φ(|y0 – y |) ∂Φ(|y0 – y |) inc inc dly = p (y ) dly + pinc (y ) dly , p (y ) ∂ny ∂ny ∂ny l+ \l◦ l◦ l+ ε ε (2.118) where lε◦ is a small ε-neighborhood of the point x0 . Since the integrand is continuous (see Section 2.1), the second integral can be made arbitrarily small. Further, i i (1) H0 (k |y0 – y |) = H0(1) (kr ) 4 4 ik (1) ∂r ∂Φ = – H1 (kr ) ∂ny 4 ∂ny r 1 2 ei(kr–π/4) ik √ cos(r , ny ) 1 + , =– 4 πkr kr kr

Φ(|y0 – y |) = ∼

(2.119) r = |y0 – y |,

and this representation can be used when estimating the first integral in (2.118). Hence, Z ∂Φ(|y0 – y |) dly I= pinc (y ) ∂ny l+ √ Z (2.120) 1 dly k ik(m⋅y+|y–y0 |) –3πi/4 √ , +O √ e cos(y – y0 , ny ) √ = e |y0 – y | 2 2π l+ \l◦ε k since the integrand is integrable and so the contribution of the l ε◦ -integral is small. To estimate (2.120) we have to study the behavior of the phase function (see the stationary phase method in Section 1.4) S (y ) = (m ⋅ y ) + |y – y0 |,

(2.121)


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√ since any stationary point makes a contribution of the order of O (1/ k ), which renders I nonsmall. Otherwise, the l + -integral is at least O (1/k ), so I is small. R RC RA RB Let us divide the integral (2.120) into three parts (see Fig. 2.6): l+ = 0 + C + 0 , where the point O = y0 is the origin of the chosen Cartesian coordinate system, A and B are points that separate the light zone l + from the shadow zone l – , and C is the summit of the contour in direction towards the incident wave. It is obvious that the phase S (y ) = (m ⋅ y )+|y | monotonically increases over the arcs CA and OB , since both terms increase. Therefore, there is no stationary point. By contrast, when y ∈ OC , the first term decreases and the second term increases so that, strictly speaking, there is a chance that a stationary point y ∗ may arise on the arc OC . However, the phase S here is S (y ) = (m ⋅ y ) + |y | = –|y | cos ψ + |y | = |y |(1 – cos ψ ), (2.122) where ψ is the angle between y and the direction towards the incident wave: ψ = yd , –m. It is clear that ψ (y ) is a monotonically increasing function if y ∈ OC , and 0 ≤ ψ ≤ π . Since cos ψ over this variation range of ψ is a decreasing function of ψ , so cos ψ decreases monotonically, but (1 – cos ψ ) increases monotonically. Since S (y ) is the product of two positive and monotonically increasing functions, |y | and [1 – cos ψ (y )], it also increases monotonically, and hence there is no stationary point√ on the arc OC . This proves that the integral I in Eq. (2.120) admits the estimate I = O (1/ k ), k → ∞, uniformly over y0 ∈ l+ , i.e., equation (2.116) may be rewritten in the form Z ∂Φ(|y0 – y |) dly = f (y0 ), y0 ∈ l + , (2.123) ϕ(y0 ) – 2 ϕ(y ) ∂ny l+ 1 + o(1) = o(1). (2.124) f (y0 ) = 4I (y0 ) + o(1) = O √ k The final part of the proof is based on the invertibility of the integral operator on the left-hand side of Eq. (2.123) and on the fact that the right-hand side is asymptotically small, uniformly over y0 ∈ l+ . Obtaining precise asymptotic estimates is not so simple here and a detailed residue analysis can be found in (Taylor, 1981). Example. High-frequency scattering pattern of a round disk. Let a plane incident acoustic wave pinc = eikx1 be incident upon an acoustically round disk (2D problem) of radius a. As follows from Eq. (2.104), if we consider the far-field scattering with the observation point x = (R cos α, R sin α), we obtain s Z k –ik(x⋅y)/|x| sc p(y ) cos (ny , x)e dly . (2.125) |p | ∼ |x| l

If we substitute into this integral the boundary value of the total acoustic pressure predicted by the Kirchhoff theory, then we notice that 3 π < ψ < π, p(y ) = p(ψ ) = 2pinc = 2eika cos ψ , 2 2 (2.126) π y = (a cos ψ , a sin ψ ). p(ψ ) = 0, |ψ | < , 2 Let us also recall (see Section 2.5) that cos (ny , x) = cos(ψ – α), (x ⋅ y )/|x| = a cos(ψ – α), so Eq. (2.125) is equivalent to s Z 3π/2 k cos(ψ – α) eika cos ψ e–ika cos(ψ–α) dψ |psc (α)| ∼ 2a |x| π/2 (2.127) s Z k 3π/2 –2ika sin(α/2) sin(ψ–α/2) cos(ψ – α) e dψ . ∼a |x| π/2


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We calculate the integral (2.127) by the stationary phase method. If ka 1, then the phase S (ψ ) = –2a sin(α/2) sin(ψ – α/2) has the only stationary point ψ = (π + α)/2 on the interval π/2 < ψ < 3π/2 for any 0 < α ≤ π . Moreover, S (ψ ∗ ) = –2a sin(α/2), cos(ψ ∗ – α) = sin(α/2), S 00 (ψ ∗ ) = 2a sin(α/2), so the stationary phase method (Section 1.4) leads to the following asymptotic expression for |psc (α)|: s |sin(α/2)| k A(α), (0 < α ≤ π ), A(α) ∼ , k → ∞. (2.128) |psc (α)| ∼ a |x| ak Due to the evident evenness, this formula can be symmetrically extended to the interval –π ≤ α < 0. The only range where the representation (2.128) is not valid is related to the case when the argument of the exponential function in (2.127) is not large for large ka. This takes place when ka sin(α/2) = O (1), i.e., α = O (1/ka), k → ∞. However, this special case admits a different estimate of the integral (2.127): s Z k 3π/2 –2ika sin(α/2) sin ψ sc cos ψ e dψ |p (α)| ∼ a |x| π/2 s Z k 3π/2 –2ika sin(α/2) sin ψ (2.129) e d(sin ψ ) =a |x| π/2 s k sin[2ka sin(α/2)] . =a |x| 2ka sin(α/2) Hence it follows that A(α) =

sin[2ka sin(α/2)] . 2ka sin(α/2)

(2.130)

As you can see, A(α) = 1 for α = 0, then outside of a small neighborhood of this value α = 0 the scattering pattern abates and becomes a rapidly oscillating function with an amplitude of the order of O (1/k ). This sharp peak in the direction of the incident wave is called the shadow formed leaf.

Helpful remarks 1◦ . Evidently, the basic ideas of Kirchhoff’s theory can be simply extended to a nonplane incident wave, since any wave at high frequency can be locally considered to be plane. 2◦ . You may ask two very important and interesting questions: 1) Why does the proved theorem require convexity of the boundary? 2) Why the same arguments cannot be applied for x0 ∈ l– ? In fact, both the condition of convexity and that of belonging of the point x 0 to the right side of the boundary are essential. Indeed, we can insure that the angle ψ (y ) monotonically increases along the arc x0 C only for convex shapes. On the other hand, if you undertake the same analysis as in the theorem, for the case x 0 ∈ l– , you may see that ψ (y ) is monotonically decreasing, and hence – cos ψ (0 < ψ < 180 ◦ ) is surely monotonically decreasing. Therefore, the phase S (y ) = |y | – |y | cos ψ is the sum of a monotonically increasing (|y |) and monotonically decreasing functions. There is a good chance for a stationary point, where S 0 = 0, to appear. This idea is clearly seen by the example of a circle of radius a. Here S (y ) = S (θ ) = |y |(1 – sin(θ/2) = 2a sin(θ/2) [1 – sin(θ/2)], whose stationary point is θ ∗ = 60◦ , as follows from the equation S 0 (θ ∗ ) = 0.


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Chapter 3

Wave Fields in a Layer of Constant Thickness 3.1. Wave Operator in Acoustic Layer: Mode Expansion, Homogeneous and Inhomogeneous Waves The present chapter will be devoted to the analytic properties of wave fields in an acoustic layer of constant thickness h. Since, for the sake of simplicity, we restrict our consideration to a 2D problem, we slightly change notations and choose a Cartesian coordinate system (x, y ) so that its horizontal axis coincides with the lower boundary of the layer y = 0, and the upper boundary surface is y = h. The considered geometry represents a problems where the Fourier transform (Section 1.1) brilliantly demonstrates its high power. Let us construct a Green’s function in the considered acoustic layer. As noted in the last paragraph of Section 2.1, it suffices to solve the nonhomogeneous Helmholtz equation with a point source represented by Dirac’s delta function on the right-hand side and placed at (x0 , y0 ): ∆Φ(x, y ) + k 2 Φ(x, y ) = δ (x – x0 ) δ (y – y0 ). (3.1) If we apply the Fourier transform with respect to the variable x, then by using Eq. (1.3) and the properties of the delta function (1.62), we arrive at the ordinary differential equation d2 e e (s , y ) = eisx0 δ (y – y0 ), Φ(s , y ) – (s 2 – k 2 ) Φ dy 2

(3.2)

where s is a parameter of the Fourier transform and the tilde denotes the Fourier image of the sought Green’s function. The solution of this equation may be expressed as the sum of the general solution to the homogeneous equation (3.2) and a particular solution of the full equation: √ e (s , y ) = A(s ) eγy + B (s ) e–γy + Φ e part (s , y ), (3.3) Φ γ = γ (s ) = s 2 – k 2 ,

where A(s ) and B (s ) are two arbitrary constants, which typically arise when solving ordinary differential equations (here, of the second order). e part , we notice that a particular solution may be chosen as one for the whole R 2 To find Φ space, since it may not satisfy any boundary conditions on the faces of the layer. So, let us apply again the Fourier transform to Eq. (3.2), this time with respect to y . Then we obtain i(sx0 +αy0 ) e ∗ (s , α) = e , Φ part α2 + γ 2

(3.4)

where the superscript ∗ denotes a Fourier image taken with respect to the variable y . Now, the application of the inverse transformation to Eq. (3.4) (see formula (1.20)) gives isx0 Z ∞ iα(y0 –y) e eisx0 e–|y0 –y|γ b part (s , y ) = e dα = (3.5) Φ 2π –∞ α2 + γ 2 2γ © 2005 by CRC Press LLC

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(the modulus |y0 – y | can replace the difference since the denominator is even in α). Hence, the general solution (3.3) reads isx0 –|y0 –y|γ e (s , y ) = A(s ) eγy + B (s ) e–γy + e e . Φ 2γ

(3.6)

Now, the functions A and B should be determined from the boundary conditions on the faces of the layer y = 0 and y = h. These may be of various type, and below we quote the most important cases with solutions. A solution is given by the inverse Fourier transform and has the form Z ∞ 1 L(s , y ) e–is(x–x0 ) ds . (3.7) Φ(x, y ) = 2π –∞

1) The acoustic layer is enclosed between two rigid plates. Then both boundary surfaces are acoustically hard, therefore the boundary conditions and a relevant function L(s , y ) are ∂p cosh[γ (h – |y – y0 |)] + cosh[γ (h – y – y0 )] ∂p = 0, = 0, L1 (s , y ) = . (3.8) ∂y y=0 ∂y y=h 2γ sinh(γh) 2) Both boundary faces are free of applied pressure. Then

p|y=0 = 0,

p|y=h = 0,

L2 (s , y ) =

cosh[γ (h – |y – y0 |)] – cosh[γ (h – y – y0 )] . 2γ sinh(γh)

(3.9)

3) One boundary of the layer is in contact with an acoustically hard plate and the other one is free of applied load. In this case sinh[γ (h – |y – y0 |)] + sinh[γ (h – y – y0 )] ∂p = 0, p|y=h = 0, L3 (s , y ) = . (3.10) ∂y y=0 2γ cosh(γh)

What can be concluded from these expressions? As can be seen, a branching function γ (s ) is present in these structures (cf. Section 1.1). However, the most important conclusion is that all these functions L(s , y ) are meromorphic with respect to s . This means that they are analytic, since their Laurent series about the branching points s = ±k contain only even powers of γ (cf. Eq. (1.18)), except no more than a countable set of poles. Let us note that the poles of the quoted functions (3.8)–(3.10) can be expressed explicitly. They coincide with zeros of the respective hyperbolic sine or cosine in the denominator and are given by (1,2) γ sinh(γh) = 0 ∼ s = ± sm , r r πm 2 πm 2 (1,2) =i – k2, sm = k2 – h h

m = 0, 1, 2, . . . ,

(3.11)

in the first two cases, and

(3) cosh(γh) = 0 ∼ s = ± sm , s s 2 2 π (m + 1/2) π (m + 1/2) (3) 2 =i – k2, sm = k – h h

m = 0, 1, 2, . . . ,

(3.12)

in the third case. It is obvious that for relatively low frequencies (small values of the wave number k ) there are no real poles, and hence all singularities lie outside of the integration contour.


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However, as the frequency increases, more and more real poles hit the real axis Im(s ) = 0. A natural question arises: In which sense should integrals of the type (3.7) be correctly treated if the integration contour crosses poles of the meromorphic integrand? It will be shown in the next section that the integration path should bend a little around the positive poles from below and the negative ones from above. Then we can apply integration with residues at simple poles sn situated in the upper complex half-plane (see the generalized Jordan lemma and the final example in Section 1.1). In order to have the possibility of arranging deformation of the integration contour in the upper half-plane (which is more often applied in the available literature) instead of the lower one, we note that all the functions L(s , y ) in (3.8)–(3.10) are even with respect to s and hence we can write the positive argument is |x – x0 | of the exponential function in Eq. (3.7) instead of the negative one –is |x – x0 |: Z ∞ ∞ 2πi X 1 is |x–x0 | Res L(s , y ) eis |x–x0| , ism . (3.13) L(s , y ) e ds = 2π –∞ 2π m=0

This yields

Φ(1,2) (x, y ) = Φ(3) (x, y ) =

∞ X (1,2) cos[πm|y – y0 |/h] ± cos[πm(y + y0 )/h] √ (–1)m eism |x–x0 | , 2 2 2 (πm) – (kh) m=0 ∞ X sin[π (m + 1/2)|y – y0 |/h] + sin[π (m + 1/2)(y + y0 )/h] p 2 π (m + 1/2)2 – (kh)2 m=0

(3.14)

(3)

× (–1)m eism |x–x0 | , which is called a mode expansion. A structure that possesses mode solutions is called a waveguide. We can thus conclude that a layer of constant thickness represents a typical waveguide. It is clear from the expansions (3.14) and Eqs. (3.11), (3.12) that if the frequency is less than a certain positive quantity, then only waves exponentially decaying with the distance |x – x0 | can exist in the layer (except that with m = 0 in symmetric cases 1 and 2; see Eq. (3.11)). Such decaying waves are called inhomogeneous mode waves. Otherwise, if the wave number is large enough, there is a finite number of waves that can propagate far along the length of the layer. Such nondecaying waves are called homogeneous mode waves. They appear to be some plane waves for any fixed 0 < y < h if we write them in their complete form with the time-dependent factor, Φ(x, y , t) = A(y ) ei(sm |x–x0| –ωt) , (3.15) (1,2) (3) where sm = sm or sm = sm , with some velocity of propagation, which for the mth mode is equal to vg = ω/sm , or "r # "r # πmc 2 –1 π (m + 1/2) c 2 –1 ω ω (3) (1,2) , vg = (3) = c . (3.16) 1– 1– vg = (1,2) = c sm ωh sm ωh

This quantity is called the group velocity of the mth mode. Obviously, the group velocity is always greater than the wave speed in the medium: v g ≥ c. Only in symmetric cases, with m = 0, the group velocity is equal to c: vg(1,2) = c (m = 0). It should also be noted that all constructed mode solutions satisfy the radiation condition, since for x → + ∞ they behave like a wave radiated to the right and for x → –∞ like that radiated to the left: Φ(x, y , t) = B (y ) ei(sm x –ωt) , x → + ∞, (3.17) Φ(x, y , t) = D (y ) e–i(sm x + ωt) , x → –∞, which is directly seen by considering the respective wave front set (cf. Section 1.9).


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Helpful remarks 1◦ . It is conventionally recognized that wave problems in closed structures, as in a layer (possibly with some discontinuities), can generate only meromorphic symbolic functions L(s , y ), and those in open structures (like a half-plane, possibly containing some discontinuities) always lead only to certain functions with branching points. We do not know whether this statement has been rigorously proved anywhere, but we know no example that refutes this statement. 2◦ . The powerful methods of the theory of complex-valued analytic functions (see Section 1.1) can be understood well by the example of the problem studied in the present section. An explicit-form solution was first obtained as a Fourier integral, whose numerical implementation is too hard, since the integrand is a rapidly oscillating function with some singularities (poles) on the real axis. Besides, the interval of integration is infinite. By applying the residue theory combined with the Jordan lemma, we reduce this integral to an equivalent form represented by exponentially convergent series, which makes the problem of calculation of the solution much easier.

3.2. Principles of Selection of Unique Solution in Unbounded Domain It is quite natural that, when solving the wave (or Helmholtz) equation in any unbounded domain, it is necessary to set some boundary condition at infinity in order to select a unique solution. Previously, we already treated the radiation condition in Section 2.2 when studying the exterior diffraction problem in the full unbounded space R n (n = 2, 3), but it has not yet been formulated as a strict mathematical condition.

Sommerfeld’s radiation condition It is conventionally accepted in the literature that a mathematically correct form of the radiation condition, where the time-dependent factor is taken in the form exp(–iωt), is given by the following pair of asymptotic identities: ∂p 1 1 – ikp = o , , R = |x| → ∞ (x ∈ R3 ), (3.18) p(x) = O R ∂R R in the 3D case, and 1 , p(x) = O √ R

∂p – ikp = o ∂R

1 √ R

R = |x| → ∞

(x ∈ R2 ),

(3.19)

in the 2D case. Indeed, if these two conditions hold, then, by using Green’s formula as in Section 2.2, we easily prove that Eq. (2.49) is valid, and consequently the Kirchhoff–Helmholtz formula (2.43) gives a unique solution to the exterior diffraction problem. This conclusion is based on the evident estimates valid if both functions u(y ) and Φ(|x – y |) satisfy conditions (3.19), since in the considered 2D case u(y )

∂u(y ) ∂Φ ∂u ∂Φ(|x – y |) – Φ(|x – y |) = u(y ) –Φ ∂ny ∂ny ∂Ry ∂Ry 1 1 1 – Φ iku + o √ =o , = u(y ) ikΦ + o √ R R R

(3.20)


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and hence Z 1 ∂u(y ) ∂Φ(|x – y |) – Φ(|x – y |) dly ∼ 2πR o → 0 as R → ∞. (3.21) u(y ) ∂ny ∂ny R lR Unfortunately, it cannot be guaranteed that conditions (3.18), (3.19) are universal, in the sense that they provide selection of a unique solution for arbitrary unbounded domain different from the whole space. In particular, it is not clear how one may treat integral representations for a mode solution in a layer (3.7)–(3.10) when the integrand has simple poles on the real axis. An inexperienced researcher may try to treat solution (3.7) in the layer, in the case where there is a finite number of simple poles on the real axis, as a singular Cauchytype principal value (see Section 1.7). In doing so, the explicit expression (3.7) can be again calculated by adding an (upper half-plane) semi-circle of a large radius R, in order to make the integration contour closed. Then by applying the generalized Jordan lemma (Section 1.1) and the boundary values of Cauchy-type singular integrals of analytic integrand (see Eq. (1.155)), we can conclude that the function Φ(x, y ) in Eq. (3.7) is Z ∞ X 1 L(s , y ) eis|x–x0 | ds = i Res[L(s , y ) eis|x–x0 | , ism ] Φ(x, y ) = 2π –∞ (3.22) i X i X Res[L(s , y ) eis|x–x0 | , sm ] + Res[L(s , y ) eis|x–x0 | , –sm ], + 2 m≤N ∗ 2 m≤N ∗ where the first sum is taken over all imaginary poles in the upper complex half-plane, the second sum is related to N ∗ positive poles, and the third one, to N ∗ negative poles. It becomes now evident that the first sum contains inhomogeneous mode waves decaying with distance, the second sum contains homogeneous waves satisfying Sommerfeld’s radiation condition, and the third sum generates homogeneous waves that contradict the radiation condition. Therefore, the integral (3.7) cannot be considered as a principal Cauchy value. A correct approach should be founded on the evident observation that the residues at positive poles are in accordance with the radiation condition, and the residues at negative poles yield mode waves arriving from infinity, thus running into a contradiction to Sommerfeld’s principle. This leads to the clear understanding that, when arranging a large closed semi-circle in the upper complex half-plane, all negative poles must be outside this circle, and the positive ones may lie inside the circle. The natural symmetry in the disposition of poles (even in the integrand!) allows us to come to the following conclusion: in order to satisfy Sommerfeld’s radiation condition, the integration contour should slightly bend around the positive poles from below and the negative ones from above. In order to establish a more physically justified principle for selecting the unique solution, some authors proposed alternative conditions at infinity, and it is very interesting to investigate, to what extent they conform to the radiation condition.

Principle of extremely low absorption (Ignatowsky’s principle) It naturally arises in electromagnetic theory (see Section 1.9). If in Eq. (1.219) the parame√ ter σ is a small positive constant (0 < σ 1), then the wave number k˜ = k + iδ , k = ω µε/c, √ δ = ωσ/(2c), 0 < δ 1, is a complex-valued quantity with positive imaginary part. This causes a finite number of simple poles sn , n = 1, 2, . . . , N ∗ , in Eqs. (3.11), (3.12), that were real-valued for real k to become complex-valued and slightly shifted off the real axis. Moreover, positive poles shift a little up, and negative ones shift a little down, so that all


Page 77

mode waves (3.13) decay with distance due to absorption. The principle of extremely low absorption (first proposed by Ignatowsky, see Tikhonov and Samarsky, 1977) states that a correct solution of the problem in an ideal (i.e., nonabsorbent) medium is the limit of the corresponding solution in the absorbent medium as δ → +0. In ideal hydroacoustic and elastic cases, where there is no parameter σ , this principle may be used too, by introducing an artificially added absorption, which in the final structure of the solution should be set equal to zero. For the problem considered in the previous section this principle leads to the same solution as the radiation condition. Indeed, if with a low absorption the positive poles move up and the negative ones down, then the natural integration contour of the inverse Fourier transform (–∞, ∞) passes below the positive poles and above the negative ones. With the absorption tending to zero, this results in the same relative disposition as in Sommerfeld’s principle.

Energy radiation condition (Mandelshtam’s principle) This principle asserts that the energy flux taken for the period of oscillations and flowing out to the outside through an appropriate cross-section of the domain must be positive. Let us calculate the energy flux through a vertical cross-section of the layer x = x ∗ . To this end, we first derive a general formula for energy power W averaged over the period of oscillations and considered on a small area with normal n : Z 2π/ω ω pvn dt. (3.23) W = 2π 0 Then we have p(x, y , t) = Re[e–iωt p˜ (x, y )] = p˜ re cos ωt + p˜ im sin ωt, 1 ∂vn 1 ∂ p˜ re ∂ p˜ im ∂vn (x, y , t) =– grad p, ∼ =– cos ωt + sin ωt , ∂t ρ0 ∂t ρ0 ∂n ∂n ∂ p˜ re ∂ p˜ im 1 sin ωt + cos ωt , – ∼ vn = ρ0 ω ∂n ∂n

(3.24)

so the substitution of relation (3.24) into Eq. (3.23) and subsequent integration determines the averaged energy in the form ∂ p˜ im ∂ p˜ re 1 p˜ re – p˜ im . (3.25) W = 2ρ0 ω ∂n ∂n On the other hand, if we operate in the harmonic regime with the standard representation expressed in terms of complex amplitudes for appropriate functions, namely p(x, y , t) = e–iωt p˜ (x, y ),

v(x, y , t) = e–iωt v˜ (x, y ),

v˜ n =

1 ∂ p˜ , iωρ0 ∂n

then (an asterisk designates here the complex conjugate) re ∂ p˜ ∂ p˜ im 1 ∗ re im +i Re(˜p v˜ n ) = Re (˜p – ip˜ ) iωρ0 ∂n ∂n im re ∂ p ˜ ∂ p ˜ 1 p˜ re – p˜ im ; = ωρ0 ∂n ∂n

(3.26)

(3.27)


Page 78

hence

1 1 Re(˜p ∗ v˜ n ) = Re(˜p v˜ n∗ ). (3.28) 2 2 Formula (3.28) allows us to calculate the energy flux through the cross-section x = x ∗ > x0 over the period, taken as the contribution of the mth homogeneous mode. To be more specific, we consider cases 1 and 2 only (tildes are omitted below): W =

cos[πm|y – y0 |/h] ± cos[πm(y + y0 )/h] ism (x–x0 ) √ e , –2i (kh)2 – (πm)2 1 ∂p cos[πm|y – y0 |/h] ± cos[πm(y + y0 )/h] √ = sm eism (x–x0 ) , vn = iωρ0 ∂x –2iωρ0 (kh)2 – (πm)2 r πm 2 > 0 (m = 1, . . . , N ∗ ), sm = k 2 – h

p=

(3.29)

and consequently

1 W = Re 2 since B=

Z

0

h

Z

0

h

p∗ vn dy =

Bsm > 0, 8ωρ0 [(kh)2 – (πn)2 ]

cos[π |y – y0 |/h] ± cos[πn(y + y0 )/h]

2

dy > 0.

(3.30)

(3.31)

The energy flux is thus found to be positive. The case of x = x ∗ < x0 can be proved likewise. Therefore, the constructed solution satisfies Mandelshtam’s energy radiation principle. In this sense both Sommerfeld’s and Mandelshtam’s radiation conditions give equivalent formulations of correct boundary conditions at infinity.

Principle of amplitude for extremely large time (Tikhonov–Samarsky principle) This principle asserts that any harmonic in a time problem where p(x, y , t) = e –iωt p˜ (x, y ) and ω (3.32) ∆p˜ + k 2 p˜ = δ (x – x0 ) δ (y – y0 ), k= , c which is a partial differential equation of the elliptic type and so does not require any initial condition, may be posed in a transient regime, 1 ∂2p = δ (x – x0 ) δ (y – y0 ) e–iωt , c2 ∂t2 as a nonstationary problem with the trivial initial conditions ∆p –

(3.33)

∂p(x, y , 0) = 0. (3.34) ∂t This new initial Cauchy problem has a unique solution under certain conditions. Then the asymptotics of this transient solution p(x, y , t) as t → ∞ gives a correct solution to the monochromatic problem in hand. Some papers were devoted to the clarification of the relationship between this principle and Sommerfeld’s radiation condition (see Tikhonov and Samarsky, 1977). However, it is noteworthy that this principle is hardly applicable in practice, since solving the arising transient problem is a more difficult task compared with the one under consideration, because the dimension of the new problem increases by one due to the introduction of a new variable, t. p(x, y , 0) =


Page 79

Helpful remarks 1◦ . If you take a look at the described principles you may come to the following conclusions. Sommerfeld’s and Mandelshtam’s principles give a correct condition at infinity in the case of full acoustic space or acoustic layer of constant thickness. It is not quite clear how they can be applied in more general cases. Principles of Ignatowsky and Tikhnov–Samarsky are more general; they give a concrete algorithm how to construct the desired solution. However, Tikhonov–Samarsky principle is hardly realizable. So, we advise that you operate always with the extremely low absorption principle. 2◦ . Nevertheless, there arises a natural question if there is known any problem where the described principles may lead to different results. In order to clarify this question, let us rewrite, in the case of scalar acoustics, the derived formula for energy flux (3.28) starting from governing equations. Let us consider an x > x 0 . Then it is obvious that the constructed solution with any mode number m gives a nontrivial solution p(x, y ) = pm (y ) eis (x–x0 ) , s = sm > 0, (3.35) to the homogeneous value problem for equation (3.1) since the right-hand side there disappears if x > x0 . Reformulated for the function pm (y ), it gives a nontrivial solution to the homogeneous boundary value problem for the ordinary differential equation 2 ω 2 00 – s pm (y ) = 0, (3.36) pm (y ) + c2 with a certain type of homogeneous boundary conditions: 1) p0m (0) = p0m (h) = 0;

2) pm (0) = pm (h) = 0;

3) p0m (0) = 0, pm (h) = 0. (3.37)

Let us differentiate Eq. (3.35) with respect to s , taking into account that s = s m (ω ) ∼ ω = ω (s ) and pm (y ) = pm (y , s ): 2 ω dω ω 2 00 – s g (y ) = 2 s – 2 pm (y ), (3.38) g (y ) + c2 c ds where we have denoted g (y ) = dpm /ds . The differentiation of the boundary conditions (3.37), 1)–3), gives 1) g 0 (0) = g 0 (h) = 0;

2) g (0) = g (h) = 0;

3) g 0 (0) = 0, g (h) = 0.

(3.39)

Now, let us take the scalar product of the conjugate of Eq. (3.36) with the function g (y ) over the interval 0 ≤ y ≤ h and the analogous scalar product of Eq. (3.38) with the function p∗m (y ) and calculate the difference of the two relations obtained: Z h Z h 00 ω dω ∗ ∗ 00 pm (y ) p∗m (y ) dy . (3.40) g (y ) pm (y ) – pm (y ) g (y ) = 2 s – 2 c ds 0 0

Here we consider the mode with positive s = sm (so that s ∗ = s ). It is easily seen that for any pair of boundary conditions 1)–3) in Eqs. (3.39), integration by parts in the first integral in Eq. (3.40) makes it trivial. This immediately leads to the relation Z Z h ω dω h |pm (y )|2 dy , (3.41) s pm (y ) p∗m (y ) dy = 2 c ds 0 0


Page 80

where, taking into account that vx (x, y ) =

1 ∂p ; iωρ0 ∂x

vx (x, y ) = vm (y ) eis(x–x0 ) ,

∼

vm (y ) =

s pm (y ), (3.42) ωρ0

the identity (3.41) finally gives Z Z h ω dω h ∗ |pm (y )|2 dy ωρ0 vm (y )pm (y ) dy = 2 c ds 0 0 Z h Z 1 dω h 1 ∗ vm pm dy = 2 |pm (y )|2 dy . ∼ W = Re 2 ρc ds 0 0

(3.43)

It becomes clear now that any mode with real group velocity (i.e., a nondecaying wave mode) provides a positive energy flux if and only if dω/ds > 0 ∼ ds/dω > 0 for any positive s (ω ). DEFINITION. The function s = s (ω ) that determines the dependence of the group velocity upon frequency is called a dispersive function. Its graph in the Cartesian coordinates is called a dispersive curve . In the considered problem for the scalar acoustic layer (see Eqs. (3.11), (3.12)), we have ω ds = >0 dω cs

if

s > 0.

(3.44)

Graphically, dispersive lines are shown in Fig. 3.1, and it is obvious that they represent monotonically increasing functions, so that ds/dω > 0 for s > 0. This proves once again that all modes propagating to the right with positive group velocities and so satisfying Sommerfeld’s radiation condition provide also a positive energy flux, i.e., they also satisfy the Mandelshtam’s energy radiation condition. Simultaneously, this proves that the integration contour in Eq. (3.7) must bend around the positive poles from below, so that the positive poles lie above the contour to give mode waves satisfying both radiation principles. Due to the evident symmetry, a similar analysis of modes propagating to the left (i.e., with s < 0) justifies that the integration contour in Eq. (3.7) should bend around the negative poles from above. 20

sh

15 10 5

kh 0

5

10

15

20

Figure 3.1. Dispersive curves of the scalar acoustic problem 3◦ . From this consideration it is clear that the integration contour, being symmetric about the origin, must pass through the origin. Therefore, all above results remain valid if no pole hits the origin. Otherwise, we deal with a so-called resonance wave process; a detailed analysis of various resonance cases, both in acoustic and elastic theory, is presented in


Page 81

(Vorovich and Babeshko, 1979). Resonance frequencies for the three considered scalar problems can be found explicitly: (1,2) sm = 0, (3) sm = 0,

c m = 1, 2, . . . , ωm = πm , h c m = 0, 1, 2, . . . ωm = π (m + 1/2) , h

(3.45)

Coming back to the question formulated in the beginning of paragraph 2 ◦ , in accordance with what is expressed by formula (3.43), we may expect that such a “nontypical” regime with negative energy flux can arise in the case where ds/dω < 0 for positive s . Figure 3.1 shows that for the considered geometry of a scalar acoustic layer of constant thickness this is never realized. We do not know any scalar acoustic or vector electromagnetic problem with such a feature. However, Mindlin (1955) first demonstrated that Sommerfeld’s radiation condition and the energy radiation condition may contradict each other in a problem for an elastic layer. The next section deals with some aspects of wave propagation in an elastic layer.

3.3. Waves in Elastic Layer As follows from equations of Sections 1.9, the 2D (in-plane) dynamic elastic problem is governed by the relations ux =

∂ϕ ∂ψ + , ∂x ∂y

∆ϕ + kp2 ϕ = 0,

uy =

∂ϕ ∂ψ – , ∂y ∂x

∆ψ + ks2 ψ = 0,

∂2ϕ ∂2ψ ∂2ψ σxy = 2 + – , ρc2s ∂x∂y ∂y 2 ∂x2 2 ∂ ϕ ∂2ψ σxx =2 + + ks2 ϕ. ρc2s ∂y 2 ∂x∂y

ω ω ks = , kp = , cp cs 2 σyy ∂ ϕ ∂2ψ =2 + + ks2 ϕ, ρc2s ∂x2 ∂x∂y

(3.46)

Here cp , cs are the longitudinal and transverse wave speeds; {u x (x, y ), uy (x, y )} are the components of the displacement vector; ϕ and ψ are the Lamé potentials; σxx , σyy , σxy are the components of the stress tensor involved in the 2D problem; and ρ is the mass density. The number of various combinations of boundary conditions here is greater than that for the scalar acoustic layer. For simplicity, we will consider only one concrete boundary value problem: y = 0: σxy = 0, uy = 0, (3.47) y = h: σxy = 0, uy = δ (x – x0 ), whose physical meaning is that the layer rests (along the line y = 0) upon a rigid foundation and the upper surface of the layer is subjected to harmonic vibrations caused by an oscillating point indentor. The application of the Fourier transform with respect to the x-variable reduces the Helmholtz equations in (3.46) to ordinary differential equations with constant coefficient, whose solutions in Fourier images are given by q Φ = A(s ) cosh(γy ) + B (s ) sinh(γy ), γ = s 2 – kp2 , (3.48) p Ψ = C (s ) cosh(qy ) + D (s ) sinh(qy ), q = s 2 – ks2 , © 2005 by CRC Press LLC

Page 82

The boundary conditions (3.47) for the Fourier transforms have the form (a prime denotes a derivative with respect to y ) – 2isΦ0 + Ψ00 + s 2 Ψ = 0, Φ0 + isΨ = 0, y = 0,

y = 0,

– 2isΦ0 + Ψ0 + s 2 Ψ = 0, y = h, ix0 s e , y = h. (ks2 – 2s 2 ) Φ – 2is Ψ0 = ρc2s

(3.49)

System (3.49) with Φ and Ψ defined by (3.48) leads to a 4 × 4 linear algebraic system for the coefficients A, B , C , D :  – 2isγB + (2s 2 – ks2 ) C = 0,       γB + isC = 0, (3.50) – 2 isγA sinh(γh) + (2s 2 – ks2 ) D sinh(qh) = 0,     eix0 s   – (2s 2 – ks2 ) A sinh(γh) – 2isq sinh(qh) = , ρc2s

whose solution is B = C = 0,

A=–

eix0 s (2s 2 – ks2 ) sinh(qh) , ρc2s ∆(s )

D=–

eix0 s 2isγ sinh(γh) , ρc2s ∆(s )

∆(s ) = (2s 2 – ks2 )2 sinh(qh) cosh(γh) – 4s 2 γq sinh(γh) cosh(qh).

(3.51) (3.52)

This finally determines all elastic functions in the layer. In particular, the Lam e´ potentials are Z ∞ (2s 2 – ks2 ) sinh(qh) cosh(γy ) is(x–x0 ) 1 e ds , ϕ(x, y ) = – 2πρc2s –∞ ∆(s ) (3.53) Z ∞ 2isγ sinh(γh) sinh(qy ) is(x–x0 ) 1 e ds , ψ (x, y ) = – 2πρc2s –∞ ∆(s ) where the evenness of the integrand allows us again to write exp[is (x – x 0 )] rather than exp[–is (x – x0 )]. In order to clarify what is the relation between the Sommerfeld and energy radiation conditions here in the case of elastic layer, let us first derive a representation for the energy flux averaged over the period T = 2π/ω : Z 2π/ω dux duy ω σxx + σxy dt, (3.54) W =– 2π 0 dt dt

where the sign “minus” is put in front of this formula because v xx and p are directed oppositely in the scalar acoustics. Now, in terms of real quantities, the components of the displacement vector are ui (x, y , t) = Re{e–iωt u˜ i (x, y )} = u˜ rei cos ωt + u˜ im (i = 1, 2), i sin ωt re im re ∂ u˜ im ∂ u˜ x ∂ u˜ y ∂ u˜ x ∂ux ∂uy y + =µ + cos ωt + µ + sin ωt (3.55) σxy = µ ∂y ∂x ∂y ∂x ∂y ∂x duy re im = ω –u˜ rey sin ωt + u˜ im = σ˜ xy cos ωt + σ˜ xy sin ωt, y cos ωt , dt © 2005 by CRC Press LLC

Page 83

and consequently ω 2π

Z

2π/ω

0

By analogy,

im re ∂ u˜ im ∂ u˜ x duy ω ∂ u˜ x ∂ u˜ rey y im u˜ y – u˜ rey σxy dt = µ + + dt 2 ∂y ∂x ∂y ∂x ω re im im re σ˜ xy u˜ y – σ˜ xy u˜ y . = 2 ω 2π

hence

Z

2π/ω 0

σxx

dux ω re im im re dt = σ˜ xx u˜ x – σ˜ xx u˜ x ; dt 2

ω re im im re re im im re σ˜ xx u˜ x – σ˜ xx u˜ x + σ˜ xy u˜ x – σ˜ xy u˜ x . 2 On the other hand, when dealing with the complex amplitudes W =–

ui (x, t) = e–iωt u˜ i (x),

we have

σij (x, t) = e–iωt σ˜ ij (x)

(3.56)

(3.57) (3.58)

(i, j = 1, 2),

(3.59)

∗ ∗ re im im re re im im re Im(σ˜ xy u˜ y + σ˜ xx u˜ x ) = (σ˜ xx u˜ x – σ˜ xx u˜ x ) + (σ˜ xy u˜ y – σ˜ xy u˜ y );

(3.60)

hence, taking into account Eq. (3.54), we obtain W =–

ω ω ∗ ∗ Im(σ˜ xx u˜ x + σ˜ xy u˜ y ) = Im(σ˜ xx u˜ ∗x + σ˜ xy u˜ ∗y ). 2 2

(3.61)

This formula is very convenient for the calculation of the energy flux, to avoid transfer from complex to real-valued quantities. Further, we will derive a different representation for W containing dω/ds , where ω (s ) and s (ω ) are dispersive functions, i.e., solutions of the Rayleigh–Lamb equation ∆(s , ω ) = 0. Let us consider any homogeneous mode obtained as the residue at a simple positive pole in Eq. (3.53), i.e., as a root of the equation ∆ = 0 in the denominator. Then the structure of any mode in terms of the displacement vector in the case x > x 0 is ui (x, y ) = uim (y ) eis(x–x0 ) ,

s = sm ,

σij (x, y ) = σijm eis(x–x0 ) ,

(3.62)

where m designates the mode number. Since for x > x0 all right-hand sides in Eqs. (3.46)–(3.47) are homogeneous, the constructed mode wave (3.62) is a solution to the homogeneous boundary value problems, which is now convenient to be taken as equations of motion (see Section 1.9) expressed through components of the stress tensor, σij,j + ρω 2 ui = 0 (i = 1, 2), with summation over the repeated index j . Then with the help of Eq. (3.62) we have ( 0 isσxxm + σxy + ρω 2 uxm = 0, m (3.63) 0 isσxym + σyy + ρω 2 uym = 0, m with the boundary conditions y = 0:

uym = 0,

y = h, σxy = 0:

σxy = 0:

u0xm + isuym = 0,

u0xm + isuym = 0,

σyy = 0:

c2p is uxm + (c2p – 2c2s )u0ym = 0

(3.64)

(a prime denotes a derivative with respect to y ).


Page 84

The differentiation of system (3.63) with respect to s gives  dω 0   isgxx + gxy ux , + ρω 2 gx = –iσxxm – 2ρω ds m dω   isg + g 0 + ρω 2 g = –iσ uy , xy y xym – 2ρω yy ds m

(3.65)

where we have denoted gxx =

dσxxm , ds

gxy =

dσxym , ds

dσyym , ds

gyy =

gx =

duxm , ds

gy =

duym . ds

(3.66)

Analogously, the differentiation of the boundary conditions (3.64) yields y = 0:

gy = 0,

gx0 + isgy = –iuym ,

y = h:

gx0 + isgy = –iuym ,

c2p isgx + (c2p – 2c2s )gy0 = –ic2p uxm .

Now, let us take the conjugate of system (3.63): ( ∗ ∗0 – isσxx + σxy + ρω 2 u∗xm = 0, m m ∗ ∗0 – isσxy + σyy + ρω 2 u∗ym = 0, m m

(3.67)

(3.68)

and combine the following scalar products: (3.65)1 ⋅ u∗xm + (3.65)2 ⋅ u∗ym – (3.68)1 ⋅ gx – (3.68)2 ⋅ gy ,

(3.69)

where the sign of the scalar product implies the integration of an equation multiplied by an appropriate function. This procedure leads to the identity Z h Z h ∗ 0 ∗0 is (gxx u∗xm + σxx g ) dy + (gxy u∗xm – σxy g ) dy x m m x 0 0 Z h Z h ∗0 ∗ ∗ 0 (3.70) g ) dy + is (gyy u∗ym – σyy (gxy uym + σxym gy ) dy + m y 0

0

= –i

Z

h

0

(σxxm u∗xm

+

∗ u∗ ) dy σxy m ym

dω – 2ρω ds

Z

h

0

(|uxm |2 + |uym |2 ) dy .

Very scrupulous mathematical manipulations with the integrals on the left-hand side in (3.70), with the help of the boundary conditions (3.64), (3.67), show that, as in the scalar acoustic case, this left-hand side is equal to zero. Then it follows from Eqs. (3.70), (3.61) that Z Z h ω h ∗ ∗ 2 dω Im(σ˜ xx u˜ x + σ˜ xy u˜ y ) dy = ρω (|u˜ xm |2 + |uym |2 ) dy . (3.71) W = 2 0 ds 0 Formula (3.71) allows us to clarify whether there is any correlation between Sommerfeld’s and Mandelshtam’s principles. Let s = sm > 0 be a positive root of the Rayleigh–Lamb equation ∆(s ) = 0 (since only squares of s are present in the structure of the function ∆(s ), if s = sm is a root of the equation, then s = –sm is also a root of it). Then the point s = sm is a simple pole in all integrands in the corresponding Fourier-transform representations of the type (3.14) for all physical quantities, so that the pair of functions ϕ(x, y , t) = ϕ˜ (y ) ei[sm (x–x0 )–ωt] ,

ψ (x, y , t) = ψ˜ (y ) ei[sm (x–x0 )–ωt] ,

(3.72)


Page 85

determines a homogeneous mode that propagates, according to Sommerfeld’s radiation condition without decay far to the right. It is now clear from Eq. (3.71) that this mode is in agreement with Mandelshtam’s energy radiation condition if and only if 1 dω (s = sm ) = > 0. ds ds/dω

(3.73)

In order to check whether condition (3.73) holds or not, we need to construct dispersive curves. In the scalar acoustics, as we remember, all dispersive curves are monotonically increasing graphs. For an elastic layer, an example of dispersive curves, as solutions to the Rayleigh–Lamb equation ∆(s , ω ) = 0, c2 = c2s /sp2 , 0 < c2 < 1, is shown in Fig. 3.2. 16

sh

12 8 4

ks h 0

4

8

12

16

Figure 3.2. Dispersive curves of the dynamic elastic problem It can be seen from this graph that there are cases with ds/dω < 0. As follows from the described theory, in such cases the modes propagating to the right bring energy from infinity, and hence contradict the energy radiation condition. Physically, this looks mystical, and this is indeed so, confirming that physical intuition may sometimes be in contradiction to strict mathematical analysis. Anyway, we should agree that the energy radiation condition is more fundamental than Sommerfeld’s principle, and so we have to reject the modes for which ds (s = sm ) < 0, sm > 0. (3.74) dω This conclusion influences the choice of the integration contour in Eq. (3.7). As regards the positive half of the contour, it should bend around the poles s m of the integrand where ds/dω > 0 from below and in the case of ds/dω < 0, from above. It should be noted once again that, apparently, Mindlin (1955) was first to indicate a case where a wave satisfying Sommerfeld’s radiation condition can bring energy from infinity. A fundamental and very impressive analysis of correlations between various radiation conditions can be found in Vorovich and Babeshko (1979); and also in Babeshko (1971).

Helpful remarks The Rayleigh–Lamb equation ∆(s , ω ) = 0, where the function ∆(s , ω ) is defined by Eq. (3.52), determines the behavior of the dispersive lines, as in Fig. 3.2. In contrast with the dispersion equation of the scalar acoustic theory, which admits simple solution (see Section 3.2), the Rayleigh–Lamb equation cannot be solved explicitly. That is why this dispersion equation is studied in the literature in detail. Many papers were devoted to asymptotic estimates for its large roots and many authors dealt with numerical analysis of the Rayleigh–Lamb equation. Some helpful results with further references can be found


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in Mindlin (1955), Vorovich and Babeshko (1979), and other papers. Detailed numerical implementation allowed the authors of those papers to discover cases of the cited irregular behavior of the dispersive curves. Note that the only dimensionless parameter whose variation changes the configuration of the dispersive lines is a relation between the longitudinal and the transverse wave speeds. Here in Fig. 3.2 we demonstrate these curves in the case where cp /cs = 2, which is very realistic for metals.

3.4. Generalized Riemann’s Zeta Function and Summation of Some Oscillating Series As we could see in Section 3.1, in many problems the application of Poisson’s summation formula transforms some given slowly convergent series to series convergent much more rapidly, so that the new series obtained can be calculated more efficiently. Unfortunately, this technique is not applicable, as a rule, for series with oscillating terms, at least in its straightforward application. Here we study the following series that will be encountered in our forthcoming analysis of high-frequency processes in a layer of constant thickness: Cβ (ξ ) =

∞ X cos mξ , β m m=1

Sβ (ξ ) =

∞ X sin mξ . β m m=1

(3.75)

As noted in the Section 1.3, these series converge in the classical sense if Re(β ) > 0. However, practical calculation for relatively small β is a difficult problem. Concrete computations show that in order to guarantee three significant digits, one should take nearly a few hundred thousand terms. Naturally, the question arises about the convergence acceleration. To this end, we attract some results from the theory of Riemann’s zeta function. Recall that ordinary and generalized Riemann’s zeta functions are defined, respectively, by the following series: ∞ ∞ X X 1 1 , ζ (s , v ) = . (3.76) ζ (s ) = s m (m + v )s m=0 m=1 It can be shown that both functions, as functions of the argument s , are analytic in the whole complex plane except the only singular point s = 1, which is a simple pole for both of them (see Bateman and Erdelyi, 1953). In what follows we will use the profound classical result given by the Hurwitz formula (see Bateman and Erdelyi, 1953) ∞ 2 Γ(β ) X cos(mξ – πβ/2) ξ = ζ 1 – β, 2π (2π )β m=1 mβ ∞ ∞ πβ X cos mξ πβ X sin mξ 2Γ(β ) cos + sin = (2π )β 2 m=1 mβ 2 m=1 mβ πβ πβ 2Γ(β ) cos Cβ (ξ ) + sin Sβ (ξ ) , Re ξ > 0. = (2π )β 2 2

(3.77)

Our final goal is to extract from Eq. (3.77) an expansion in powers of ξ . To this end, we will try to expand the second relation in (3.76) into a Taylor series about the point v = 0,


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taking into account that the derivative of ζ (s , v ) with respect to v at v = 0 is the ordinary Riemann’s zeta function with some power s . Namely, we obtain from (3.76) ∞

∞

1 1 X s (s + 1) . . . (s + m – 1) 1 X = + (–1)m ζ (s + m) v m ζ (s , v ) = s + s s v m=1 (m + v ) v m=0 m! ∞

Γ(s + m) 1 X (–1)m ζ (s + m) v m = s + v m=0 Γ(s ) m!

(3.78)

(0! = 1).

Therefore, the left-hand side of (3.77) admits the alternative representation m 1–β X ∞ ξ 2π m Γ(m + 1 – β ) + (–1) ζ (m + 1 – β ) . ζ (1 – β , ξ/2π ) = ξ Γ(1 – β ) m! 2π m=0

(3.79)

Since the left-hand sides of Eqs. (3.77) and (3.79) are identical, their right-hand sides should coincide with each other, too.PThis allows us to construct P∞ two independent power ∞ expansions in (ξ/2π ), separately for m=1 cos(mξ )/mβ and m=1 sin(mξ )/mβ , since the even powers of ξ in (3.79) refer to the former and the odd powers, to the latter. The only obstacle consists in the presence of the irregular term (2π/ξ ) 1–β in (3.79). One part of this term belongs to a series of cosines and the other part to a series with sines. Correct separation of the irregular term is not simple. To do this, we have to use Poisson’s formula (see, for example, Collin, 1960), which is usually useless in the treatment of oscillating series. In our case it is applied in the following way: # " ∞ ∞ ∞ imξ X X X e 1 eimξ H (2πm) , (3.80) = eiξ = eiξ + H (0) + 2 β β m ( m + 1) 2 m=0 m=1 m=1 where

H (2πm) =

Z

∞ 0

cos(2πmu)

eiuξ du, (u + 1)β

m = 0, 1, 2, . . . .

(3.81)

If m is a positive integer here (m ≥ 1), then the last relation yields only a regular power expansion in ξ for H (2πm). This fact can be proved by expanding exp(iuξ ) into a Taylor series, since the integral Z ∞ up du (3.82) cos(2πmu) (u + 1)β 0

has a finite generalized value for arbitrary large positive integer p (cf. Section 1.3), which implies an analytic continuation with respect to p, from the range –1 < Re(p) < 0, where it converges in the classical sense. The irregular term is therefore extracted from the case m = 0 as follows: Z ∞ iu Z ∞ e du eiuξ β–1 du = ξ H (0) = β (u + 1) (u + ξ )β 0 0 (3.83) Z ∞ iu e du β–1 πi(1–β)/2 = Γ (1 – β ) ξ e ( ξ → +0), ∼ ξ β–1 uβ 0 and hence ∞ ∞ X X eimξ β–1 πi(1–β)/2 = Γ (1 – β ) ξ e + am ξ m β m m=0 m=1

(ξ → +0),

(3.84)


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with some coefficients am of the regular power part. By separating the real and imaginary parts in the last relation, we see that ∞ ∞ hπ i X X cos mξ β–1 = Γ(1 – β ) cos (1 – β ) ξ + bm ξ 2m , (3.85) β m 2 m=0 m=1 ∞ ∞ hπ i X X sin mξ β–1 = Γ (1 – β ) sin (1 – β ) ξ + dm ξ 2m+1 . β m 2 m=0 m=1

(3.86)

As stated above, the coefficients of the regular expansions can be defined by comparing the right-hand sides in formulas (3.77) and (3.79), which finally yields ∞ hπ i X (2π )β cos mξ β–1 = Γ (1 – β ) cos (1 – β ) ξ + Cβ (ξ ) = mβ 2 2 Γ(β ) Γ(1 – β ) cos(πβ/2) m=1 (3.87) 2m ∞ X ξ Γ(2m + 1 – β ) ζ (2m + 1 – β ) (ξ , β > 0), × (2m)! 2π m=0 ∞ hπ i X (2π )β sin mξ β–1 = Γ (1 – β ) sin (1 – β ) ξ – mβ 2 2 Γ(β ) Γ(1 – β ) sin(πβ/2) m=1 (3.88) ∞ 2m+1 X ξ Γ(2m + 2 – β ) ζ (2m + 2 – β ) (ξ , β > 0). × (2m + 1)! 2π m=0

Sβ (ξ ) =

In order to estimate how rapidly the terms of the last series decrease with increasing m, let us apply the Sterling formula (see Abramowitz and Stegun, 1965) r Γ(z + α) 2π –z z e z (z → +∞); hence ∼ z α–γ (z → +∞). (3.89) Γ(z ) ∼ z Γ(z + γ )

This estimate shows that the terms of the series (3.87) and (3.88) decrease according to O [m–β (ξ/2π )2m ] with m → ∞, because (2m)! = Γ(2m + 1) and (2m + 1)! = Γ(2m + 2). It should also be noted that it suffices to evaluate them on the interval 0 < ξ ≤ π only, owing to the evident periodicity: Cβ (2π – ξ ) = Cβ (ξ ), Sβ (2π – ξ ) = –Sβ (ξ ). Therefore, the rate of convergence of the series (3.13) is the same as that of a geometric progression {q m } with basis q ≤ 1/2. In practice, it suffices to keep 3 or 4 terms to provide a good accuracy. Let us quote explicit expressions for particular cases, which will appear in the next section: r ξ2 ξ4 ξ6 π – 1.460 + 0.0255 + 0.00444 + 0.00267 , (3.90) C1/2 (ξ ) ≈ 2ξ 2! 4! 6! r ξ3 ξ5 π – 0.208 ξ – 0.00852 – 0.00309 , (3.91) S1/2 (ξ ) ≈ 2ξ 3! 5! p ξ4 ξ6 ξ2 + 0.00852 + 0.00309 , (3.92) C3/2 (ξ ) ≈ – 2πξ + 2.612 + 0.208 2! 4! 6! p ξ5 ξ3 + 0.00444 , (3.93) S3/2 (ξ ) ≈ 2πξ – 1.460 ξ + 0.0255 3! 5! √ ξ2 ξ4 ξ6 2 2π 3/2 ξ + 1.341 + 1.460 – 0.0255 – 0.00444 , (3.94) C5/2 (ξ ) ≈ – 2! 4! 6! √3 ξ3 ξ5 2 2π 3/2 ξ + 2.612 ξ + 0.208 + 0.00852 . (3.95) S5/2 (ξ ) ≈ – 3 3! 5!


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In practice all of them are to be used for 0 < ξ ≤ π and then to be continued periodically. The relative error of these formulas is less than 1%. The accuracy of these formulas can be improved by taking more terms in the expansions. Below we give a comparison of results predicted by Eqs. (3.90)–(3.95) with those obtained by direct computations by formulas (3.75) with a double-precision Fortran code and 5 × 10 5 terms retained in the series: ξ π/8 π/4 3π/8 π/2 5π/8 3π/4 7π/8 C1/2 , (3.75) 0.5418 –0.0380 –0.2874 –0.4275 –0.5136 –0.5665 –0.5954 C1/2 , (3.90) 0.5420 –0.0379 –0.2872 –0.4274 –0.5135 –0.5664 –0.5956 S1/2 , (3.75) 1.916 1.249 0.9067 0.6672 0.4743 0.3058 0.1501 S1/2 , (3.91) 1.918 1.250 0.9073 0.6675 0.4745 0.3060 0.1506 C3/2 , (3.75) 1.058 0.4552 0.0366 –0.2705 –0.4938 –0.6465 –0.7357 C3/2 , (3.92) 1.057 0.4548 0.0363 –0.2708 –0.4939 –0.6466 –0.7359 S3/2 , (3.75) 0.9975 1.077 1.007 0.8645 0.6783 0.4653 0.2363 S3/2 , (3.93) 0.9977 1.077 1.008 0.8651 0.6790 0.4659 0.2366 C5/2 , (3.75) 1.043 0.6283 0.2160 –0.1533 –0.4573 –0.6825 –0.8207 C5/2 , (3.94) 1.042 0.6278 0.2153 –0.1542 –0.4585 –0.6840 –0.8223 S5/2 , (3.75) 0.6167 0.9054 0.9976 0.9486 0.7961 0.5700 0.2966 S5/2 , (3.95) 0.6166 0.9051 0.9972 0.9481 0.7954 0.5691 0.2952

Helpful remarks The above ideas can be applied to the calculation of sums of divergent series. As described in Section 1.3, identities (3.87) and (3.88) can be analytically continued over the parameter β in the complex plane, taking into account those β for which both the left- and the right-hand sides are finite. As an example, we quote here two remarkable divergent series ∞ ∞ X X 1/2 C–1/2 (ξ ) = m cos mξ , S–1/2 (ξ ) = m1/2 sin mξ , (3.96) m=1

m=1

and compare results predicted by the direct summation and the method using the derived formulas (3.87), (3.88) with β = –1/2. These series are expressed as r ξ2 ξ4 ξ6 1 π –3/2 ξ – 0.208 – 0.00852 – 0.00309 – 0.00346 , (3.97) C–1/2 (ξ ) ≈ – 2 2 2! 4! 6! r 3 5 ξ ξ 1 π –3/2 ξ – 0.0255 ξ – 0.00444 – 0.00267 . (3.98) S–1/2 (ξ ) ≈ 2 2 3! 5! In the direct summation we have applied ideas proposed in Section 1.3. More precisely, we set x = 0.999 and take the upper limit of summation 5 × 10 4 in the series ∞ ∞ X X m 1/2 x m cos mξ , xm m1/2 sin mξ , (3.99) m=1

m=1

treated as a generalized value for the series (3.96). Results of the comparison are reflected in the following table: ξ π/8 π/4 3π/8 π/2 5π/8 3π/4 7π/8 C–1/2 , (3.75) –2.745 –1.109 –0.7035 –0.5372 –0.4540 –0.4094 –0.3869 C–1/2 , (3.97) –2.755 –1.111 –0.7041 –0.5376 –0.4543 –0.4096 –0.3870 S–1/2 , (3.75) 2.5461 0.8816 0.4594 0.2755 0.1715 0.1017 0.0476 S–1/2 , (3.98) 2.5364 0.8799 0.4588 0.2752 0.1714 0.1019 0.0485 An impressive table for divergent series, isn’t it?


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3.5. Application: Efficient Calculation of Wave Fields in a Layer of Constant Thickness Let us come back to the problem of efficient calculation of wave fields in a layer of constant thickness. We will deal here only with acoustic case, since the proposed ideas can be applied to elastic case, too. To begin with, we note that for arbitrary fixed k only a finite number of terms, N , in the mode representation (3.14) have a real group velocity s n . For all other terms when n > N , the quantities sn are imaginary and the modulus |sn | grows with n. Therefore, this series (3.14) converges like a geometric progression, in the case of x ≠ x 0 . Many investigations have been devoted to the question of efficient calculation of acoustic waves in a layer of constant thickness in the case of x = x0 , and we give here a short survey of existing methods. In principle, a direct calculation of the integral (3.7), where L(s ) is defined by any of Eqs. (3.8)–(3.10), gives an acceptable tool for calculations, since all singularities of the integrand lie on the finite interval –k < s < k . Outside of this interval the integrand decays exponentially (we exclude the case y = y0 from the consideration, because in the case that x → x0 , y → y0 simultaneously, the constructed Green’s function has a logarithmic singularity). But in practice this approach is never applied for the following two reasons. Firstly, a finite number of real-valued singularities are all simple poles, and the integral can be reduced to a Cauchy-type singular integral by adding together the halves of the residues at appropriate poles. However, direct numerical treatment of singular integrals is a more complex problem compared to those with continuous integrands. Secondly, numerical calculation of any integral (even with continuous integrand) is always a more difficult problem than the calculation of the sum of a series. So, now we are going to transform the integral to some series. On the other hand, the mode expansions (3.14) with x = x 0 cannot be used, as they are written, for direct numerical calculations, since the common term decreases as 1/n. The series converges, due to the presence of an oscillating factor, but it is absolutely unsuitable for direct computations. To overcome this difficulty, some authors apply the so-called Kummer’s transformation in this case (see, for example, Jorgenson and Mittra, 1990; Mathis and Peterson, 1996). If x = x0 , then the mode expansions (3.14) can be represented as a combination of the following series: ∞ X (–1)m cos(p(1,2) m η) , F (η ) = qm sin(p(3) m η) m=0 p p(1,2) qm = p2m – k 2 , m = πm, ±

More precisely,

1 Φ (x0 , y ) = 2 (1)

1 Φ (x0 , y ) = 2 (2)

1 Φ (x0 , y ) = 2 (3)

F

+

F

+

F

–

|y – y0 | h |y – y0 | h |y – y0 | h

p(3) m = π m+

+F

+

–F

+

+F

–

y + y 0

h

y + y 0

h

y + y 0

h

1

2

(3.100) .

, ,

(3.101)

,

and the problem can be reduced to the calculation of the series (3.100). This is based on


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the observation that X ∞ ∞ X (–1)n cos(p(1,2) 1 1 cos(p(1,2) n η) n η) n ± , – + (–1) F (η ) = sin(p(3) qn rn rn sin(p(3) n η) n η n=0 n=0 p p rn = p2n + b2 . qn = p2n – k 2 ,

(3.102)

where F (s ) is the cosine Fourier transform of f (x): Z ∞ F (s ) = cos(sx) f (x) dx.

(3.104)

Now, for any fixed value of the parameter k the common term in the first series decreases as 1/n3 , and the series is quite suitable for efficient calculations, but the second series converges still as slowly as in (3.14). However, this admits the application of the Poisson summation formula (see, for example, Titchmarsh, 1948). We already treated some series by this formula in the previous section, and give here a more complete overview. If f (x) ∈ C [0, ∞), f (x) ∈ L1 [0, ∞) and a is a parameter, then ∞ ∞ X X 2πn a F , (3.103) a f (an) = f (0) + F (0) + 2 2 a n=1 n=0

0

This formula has a finite-form analogue: a

n X n=0

Since

f (an) =

∞

X a F [f (0) + f (a)] + F (0) + 2 2 n=1

2πn a

.

X ∞ ∞ X 1 cos(p(1,2) cos(p(1,2) (–1)n cos(p(1,2) n η) n ) n η) = , rn sin(p(3) r sin(p(3) sin(p(3) n η) n ) n η) n=0 n n=0

(3.105)

(3.106)

the transformation of the products of trigonometric functions to their sums reduces the calculation of the second series (3.102) to the summation of some other series, which we consider separately in the (1, 2) and (3) problems. The top line in Eq. (3.106) leads to a series of the form S (1,2) =

∞ X cos(πnζ ) √ , (πn)2 + b2 n=0

(3.107)

which admits immediate √ application of the Poisson formula (3.103)–(3.104), since here a = π , f (x) = cos(xζ )/ x2 + b2 and Z Z ∞ cos(sx) cos(xζ ) 1 ∞ cos[x(s + ζ )] + cos[x(s – ζ )] √ √ = F (s ) = 2 0 x 2 + b2 x 2 + b2 0 (3.108) K0 (b |s + ζ |) + K0 (b |s – ζ |) , = 2 where K0 is the McDonald function, which exponentially decreases with increasing argument. Therefore, in these problems we arrive at the series S

(1,2)

∞ K0 (b |ζ |) 1 X 1 + + = [K0 (b |2n + ζ |) + K0 (b |2n – ζ |)] , 2b π π n=1

(3.109)

which converges exponentially.


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In problem (3) the bottom line in (3.106) yields a series of the form S

(3)

∞ ∞ X 1 X exp[πi(n + 1/2)ζ ] cos[π (n + 1/2)ζ ] p p = = [π (n + 1/2)]2 + b2 2 n=–∞ [π (n + 1/2)]2 + b2 n=0 ∞ exp(πnζi) exp(πζi/2) X p = 2 [π (n + 1/2)]2 + b2 n=–∞ # " ∞ exp(πnζi) exp(πnζi) exp(πζi/2) X p –√ = 2 (πn)2 + b2 [π (n + 1/2)]2 + b2 n=–∞ # "∞ X cos(πnζ ) 1 √ , – + exp(πζi/2) 2 2 2b (πn) + b n=0

(3.110)

where the common term of the first series on the right-hand side (3.110) decreases again as 1/n3 and the last series is reduced to S (1,2) , already calculated. Thus, for low and moderate k , Kummer’s transformation leads to quite rapid convergence. However, this transformation becomes inefficient with increasing frequency parameter k , due to a specific behavior of the difference in parentheses in Eq. (3.102). That is why we develop a different approach. Let us note that all integrands in Eqs. (3.7)–(3.10), Section 3.1, represent some combinations of the following type: L± (s , ζ ) =

More precisely,

e–ζγ , 2γ 1 ± e–2γh

ζ > 0.

L1 (s , y ) = L– (s , |y – y0 |) + L– (s , 2h – |y – y0 |) + L– (s , y + y0 ) + L– (s , 2h – y – y0 ), L2 (s , y ) = L– (s , |y – y0 |) + L– (s , 2h – |y – y0 |) – L– (s , y + y0 ) – L– (s , 2h – y – y0 ), L3 (s , y ) = L+ (s , |y – y0 |) – L+ (s , 2h – |y – y0 |) + L+ (s , y + y0 ) – L+ (s , 2h – y – y0 ).

Therefore, the problem is reduced to efficient calculation of the integral Z ∞ 1 L± (s , ζ ) e–is(x–x0 ) G(x, ζ ) = 2π –∞ Z ∞ –γ(2nh +ζ) ∞ e 1 X n (∓1) e–is(x–x0 ) ds = 4π n=0 γ –∞ ∞ h p i i X (∓1)n H0(1) k (2nh + ζ )2 + (x – x0 )2 = 4 n=0 p i = H0(1) k (x – x0 )2 + ζ 2 + G1 (x, y ), 4 ∞ h p i i X (∓1)n H0(1) k (2nh + ζ )2 + (x – x0 )2 , 0 ≤ ζ ≤ 2h. G1 (x, y ) = 4 n=1

(3.111)

(3.112)

(3.113)

Recall that now we study summation for x = x0 , where the first N terms (N is sufficiently large) can be calculated directly as N i X (∓1)n H0(1) [k (2nh + ζ )] , 4 n=1

(3.114)


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and for n > N we can use the asymptotic representation of the Hankel function (see Bateman and Erdelyi, 1953) r ∞ Γ 21 + m 2 i(z–π/4) X (1) , z → ∞. (3.115) e H0 (z ) ∼ 1 πz – m (–2zi)m m ! Γ 2 m=0

As a result, the summation problem is reduced to the calculation of a series of the following type: ∞ X e2khni . (3.116) (∓1)n (2nh + ζ )m+1/2 n=1 Let us write out a few first terms in the Taylor expansion of the denominator: 1 (m + 1/2) y (m + 1/2)(m + 3/2) y 1 = – + (n + ξ )m+1/2 nm+1/2 nm+3/2 nm+5/2 1 (m + 1/2) y (m + 1/2)(m + 3/2) y 1 – + – , + (n + ξ )m+1/2 nm+1/2 nm+3/2 nm+5/2

ζ ξ= . 2h

(3.117)

The sum in square brackets here tends to zero at least as n –7/2 with n → ∞, and it suffices to keep only a few terms in the series to obtain a good accuracy, uniformly over the frequency parameter k . Thus, the problem is reduced to a slowly convergent series of the form ∞ X eδni = Cβ (δ ) + iSβ (δ ), nβ n=1

(3.118)

with some values of δ and β ; this series was studied in the previous section.

Helpful remarks It is interesting to note that in the representation (3.113) each term is similar to the pointsource Green’s function for the full unbounded space. This is in complete agreement with the heuristic idea that the problem of determining the Green’s function in a layer differs from the analogous problem in unbounded space by additional boundary conditions ∂p/∂n over two straight lines. Any pair of sources displaced symmetrically with respect to any straight line automatically provides trivial normal derivative on this line. From this point of view, Eq. (3.113) can be treated as representing a sequence of required sources to arrange a symmetrical balance of the total array with respect to these boundary lines.

3.6. Waves in the Stratified Half-Plane Here we will show that a stratified half-space possesses, under certain conditions, the properties of a layer of constant thickness established in Section 3.1. Let us consider a 2D problem for a stratified acoustic half-plane y > 0, |x| < ∞, where the wave process is assumed to be harmonic in time with angular frequency ω . Then, the acoustic pressure p satisfies the Helmholtz equation ∆p(x, y ) +

ω2 p(x, y ) = 0, c2 (y )

(3.119)

where the wave speed c(y ) is a function of the coordinate y varying with the depth of the half-plane.


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The application of the Fourier transform with respect to the variable x reduces the partial differential equation (3.119) to the following ordinary equation of the second order with variable coefficient: r ω (3.120) P 00 (s , y ) – γ˜ 2 (s , y ) P (s , y ) = 0, γ˜ (s , y ) = s 2 – 2 . c (y ) The sign of the derivative is related to the variable y . As usual, the boundary condition may be of various type. To be more specific, we consider a point source placed on the free boundary surface y = 0 at a point x 0 : y = 0:

p = δ (x – x0 ).

(3.121)

In this scalar acoustic problem wave behavior at infinity may be accepted to satisfy Sommerfeld’s radiation condition. Generally, Eq. (3.120) cannot be solved analytically. For some specific functions c(y ) its solution is expressed in terms of certain special functions, and a good survey of these cases admitting exact analytical solution is given by Brekhovskikh (1980). In the present section we construct an explicit analytical solution for high frequencies. In doing so, we assume the wave speed c(y ) to be a monotonic and bounded function of the argument y > 0, at least twice differentiable: c(y ) ∈ C 2 [0, ∞), m ≤ |c(y )| ≤ M . In ocean acoustics, a monotonic behavior of c(y ) is valid for Arctic zone, as well as for moderate latitudes in the winter period of the year. Let us apply the change of variable s = ω s˜ (the tildes are omitted later on), then Eq. (3.120) becomes s 1 (3.122) P 00 (s , y ) – ω 2 γ 2 (s , y ) P (s , y ) = 0, γ (s , y ) = s 2 – 2 . c (y ) The further considerations for the cases of monotonically decreasing and monotonically increasing functions c(y ) are absolutely different. The case of monotonically decreasing wave speed c(y ). Here c(0) = c1 , c(∞) = c2 , c1 > c2 , c2 ≤ c(y ) ≤ c1 .

(3.123)

Since the dependence upon the parameter s in Eq. (3.122) is even, we consider only its positive values 0 ≤ s < ∞. Case 1. Let 1 1 or s > . (3.124) 0≤s< c1 c2 For such values of the parameter s , the function γ (s , y ) is nonzero. Then for high frequencies ω , Eq. (3.122) can be solved by the classical WKB method (see, for instance, Erdelyi, 1956; Brekhovskikh, 1980) to give Z y 1 –ωξ(s,y) ωξ(s,y) [A(s ) e + B (s ) e ], ξ (s , y ) = γ (s , y ) dy , (3.125) P (s , y ) = γ (s , y ) 0 where A(s ) and B (s ) are some unknown functions. It is noteworthy that the representation (3.125) is valid both in the case s > 1/c 2 > 1/c(y ), where γ (sp , y ) > 0, and in the case 0 < s < 1/c1 < 1/c(y ), where γ (s , y ) = –iq (s , y ) and q (s , y ) = [1/c(y )]2 – s 2 > 0. In both cases, the second term in Eq. (3.125) should be neglected, because for inhomogeneous waves (γ > 0) it gives an unbounded solution (since


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ξ (s , y ) → +∞ as y → +∞), and for homogeneous waves (γ = –iq , qp> 0) this gives a solution with a far-field behavior of the type exp(–ibωy ), y → +∞ (b = (1/c2 )2 – s 2 > 0), which represents a wave arriving from infinity rather than a wave going to infinity. The first term in (3.125) certainly satisfies Sommerfeld’s radiation condition. Thus, in the considered range of variation of the parameter s we have B = 0; hence P (s , y ) =

A(s ) –ωξ(s,y) e , γ (s , y )

or

P (s , y ) = P (s , 0)

γ (s , 0) –ωξ(s,y) e . γ (s , y )

(3.126)

Case 2. Now consider the case of 1/c1 < s < 1/c2

∼

c2 < 1/s < c1 .

(3.127)

Then there is always a unique solution of the equation γ = 0 ∼ c(y 0 ) = 1/s , since c(y ) is monotonic and c2 < c(y ) < c1 . The value y = y0 = y0 (s ) is a simple root of the equation γ = 0: γ (s , y0 ) = 0, γ 0 (s , y0 ) ≠ 0. This point y0 (0 < y0 < ∞) is called a turning point of the differential equation (3.122). It should be noted that we assumed the function c(y ) to be monotonic just to guarantee that the root y0 is simple. The general solution of Eq. (3.122) in this case is known to be expressed as follows (see Erdelyi, 1956; Brekhovskikh, 1980): 1 A(s ) Ai [ω 2/3 ξ (s , y )] + B (s ) Bi [ω 2/3 ξ (s , y )] , P (s , y ) = √ 0 ξ (s , y ) (3.128) 2/3 Z y 3 , γ (s , y ) dy ξ (s , y ) = 2 y0 where Ai(x) and Bi(x) are the Airy functions, A(s ) and B (s ) are some unknown functions, and the derivative is applied with respect to y . Let us study the behavior of solution (3.128) at infinity. Note that c 2 < c(y ) < c(y0 ) = 1/s for y > y0 , so in this range s s 1 1 = –iq , q = q (s , y ) = – s 2 > 0; hence as y → +∞, γ = s2 – 2 2 c (y ) c (y ) s 2/3 2/3 3 1 3 iq2 (s )y =– q2 (s )y , q2 (s ) = – s 2 > 0. ξ (s , y ) ∼ 2 2 c22 (3.129) Then we should use the asymptotics of the Airy functions for large negative arguments (see Fedorjuk, 1977): Ai(–y ) ∼

sin(η + π/4) , π 1/2 y 1/4

In the considered case η ∼ ωq2 y ,

hence, as y → +∞,

Bi(–y ) ∼

cos(η + π/4) , π 1/2 y 1/4

y → +∞,

1/3 2 2/3 q2 y –1/3 , ξ (y ) ∼ – 3 0

η=

2 3/2 y . (3.130) 3

y → +∞,

–1/3 –1/6 P (s , y ) ∼ const q2 y 1/6 ω –1/6 q2 y –1/6 × A(s ) sin ωq2 y + π4 + B (s ) cos ωq2 y + π4 = const A(s ) sin ωq2 y + π4 + B (s ) cos ωq2 y + π4 .

(3.131)

(3.132)


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This solution must satisfy the radiation condition, i.e., must be of the form P (s , y ) ∼ const(s ) eidy , d > 0, y → +∞. (3.133) This can be provided if and only if A(s ) = iB (s ). (3.134) Therefore, in this case the general solution of Eq. (3.132) that satisfies the radiation condition at infinity is expressed as B (s ) i Ai[ω 2/3 ξ (s , y )] + Bi[ω 2/3 ξ (s , y )] ; (3.135) P (s , y ) = √ 0 ξ (s , y ) hence s

ξ 0 (s , 0) i Ai[ω 2/3 ξ (s , y )] + Bi[ω 2/3 ξ (s , y )] . (3.136) ξ 0 (s , y ) i Ai[ω 2/3 ξ (s , 0)] + Bi[ω 2/3 ξ (s , 0)] Let us study the structure of the solution for any fixed 0 < y < y 0 : Z γ (s , 0) –ωξ(s,y) is(x–x0 ) 1 e e ds p(x, y ) = π |s|≤1/c1 γ (s , y ) s Z ξ 0 (s , 0) i Ai[ω 2/3 ξ (s , y )] + Bi[ω 2/3 ξ (s , y )] is(x–x0 ) e ds (3.137) + ξ 0 (s , y ) i Ai[ω 2/3 ξ (s , 0)] + Bi[ω 2/3 ξ (s , 0)] 1/c1 0, so the phase in the argument of the exponential function is positive, and we have, as ω → ∞, a exponentially decaying kernel. Then the asymptotics (1.68) with β = 1/2 and α = 1 gives the same estimate O (ω –1/2 ) for I3 . We thus can conclude that homogeneous waves can exist only in the case that the denominator of the second integrand a pole. However,pif 0 < y < y 0 , then 3 R y itself has2/3 > 0, since γ = s 2 – [1/c(y )]2 > 0. c(y ) > c(y0 ) = 1/s and so ξ (s , y ) = 2 y0 γ (s , y ) dy It is clear from this observation that in the considered range we can apply the asymptotic expansion of the Airy functions, present in the integrand of I 2 , for large positive argument: eη 2 e–η (3.138) Ai(y ) ∼ √ 1/4 , Bi(y ) ∼ √ 1/4 , η = y 3/2 , y → +∞, 2 πy 2 πy 3 which converts I2 into the following expression: s 1/4 Z ξ 0 (s , 0) ξ (s , 0) I2 (x, y ) = ξ 0 (s , y ) ξ (s , y ) 1/c1 0 when ε > 0 (see Section 1.1). Hence, the last series is a geometric progression, which yields ∞ Z 1 X ∞ exp[–γ (α)x] + exp[γ (α)(x – 2a)] e–iα|2bn+y| dα s (x, y , a, b) = πi n=–∞ –∞ 1 – exp[–2aγ (α)] γ (α) (4.105) ∞ Z 1 X ∞ cosh[(a – x)γ (α)] e–iα|2bn+y| dα. = πi n=–∞ –∞ sinh[aγ (α)] γ (α)

There is no obstacle here to applying the standard expansion in residues at simple poles, because the integrand is an exponentially decaying meromorphic function with the integration contour removing down to infinity. When doing so, we arrive at the following result: ∞ ∞ X cos(πmx/a) X exp –|2bn + y | qm /a , s (x, y , a, b) = –i δm qm (4.106) n=–∞ m=0 p qm = (πm)2 – (akε )2 , δ0 = 1, δm = 2 (m = 1, 2, . . . ). Since the real part of the quantities qm is positive with ε > 0, the last series is again a geometric progression, with its sum being equal to X ∞ ∞ X (2bn – y ) (2bn + y ) qm + qm exp – exp – a a n=1 n=0 (4.107) exp[–yqm /a] + exp[(y – 2b)qm /a] cosh[(b – y )qm /a] = . = 1 – exp[–2bqm /a] sinh[bqm /a]

Therefore, the final representation (ε → +0) may be obtained from (4.106) and (4.107), as follows: √ ∞ X cosh (πm)2 – (ak )2 (b – y )/a cos(πm x/a) √ . (4.108) s (x, y , a, b) = –i δm √ (πm)2 – (ak )2 sinh (πm)2 – (ak )2 b/a m=0

It should be noted that if 0 ≤ y < 2b, then the last series converges exponentially, i.e., its terms decrease as in some geometric progression with a certain basis q : |q | < 1, when the number m infinitely grows. Extension to more complex polygons. It is not easy to extend the image method to more complex polygonal shapes. Irregular shapes involve a more complicated procedure for finding the positions of the virtual sources. It is important to clarify the intrinsic property of the exact solution for rectangular spaces. If we look at Fig. 4.1 again, it becomes clear that four virtual sources crowd around each corner of the main rectangle, symmetrically with respect to its faces, at equal distances from the corner. They correspond to virtual sources in the exact solution for a point source placed in a quarter-plane. Analogous constructions may be applied to the discovered three triangle geometries, which admit exact analytical solution. In the case of equilateral triangle, the corresponding array of virtual sources is shown in Fig. 4.2. It becomes clear from this figure that the structure of the solution is as follows: √ 12 X 3 3 c, (4.109) p(x1 , y1 ) = s (ξi – x1 , ηi – y1 , a, b), a = c, b = 2 2 i=1

where c is the length of the side.


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y

7-12 c

1-6 x

Figure 4.2. Arrangement of virtual sources for equilateral triangle For isosceles right-angled triangle with the leg a, as it follows from Fig. 4.3, the solution can be expressed in the following way: p(x1 , y1 ) =

8 X

s (ξi – x1 , ηi – y1 , a, a).

(4.110)

i=1

y

a

1-8 x

Figure 4.3. Virtual sources for isosceles right-angled triangle At last, the structure of the solution for right-angled triangle with the acute angles 30 ◦ and 60◦ and the hypotenuse c becomes clear from Fig. 4.4: p(x1 , y1 ) =

24 X i=1

s (ξi – x1 , ηi – y1 , a, b),

√ 3 c, a= 2

b=

3 c. 2

(4.111)


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y 13-24

c 1-12

x

Figure 4.4. Virtual sources for right-angled triangle with the acute angles 30 ◦ and 60◦ Estimate of the eigenfrequencies. Eigenvalues of the wave number k are determined by Eq. (4.108) when the hyperbolic sine in the denominator becomes equal to zero: i hp p (πm)2 – (ak )2 b/a = 0 ∼ (πm)2 – (ak )2 b/a = –πni sinh r (4.112) m2 n2 + 2 ∼ kmn = π a2 b for arbitrary integer values of m and n. For the rectangular domain with sides a and b this result is well known in the literature. For the triangular spaces this involves some new exact estimates: 1) Equilateral triangle with the side-length c: 2π √ 2 m + 3n2 . (4.113) kmn = 3c It can be easily seen that in the case of the considered triangle the first six terms i = 1, 2, . . . , 6 in the sum (4.109) and the other six terms i = 6, 7, . . . , 12 cancel each other if n + m is odd. Indeed, the second crowd of the six virtual sources is obtained from the first six corresponding sources √ if we change (ξi , ηi ) by (ξi + a, ηi + b). Expression Tmn (x, y ) = cos(πm x/a) cosh (πm)2 – (ak )2 (b – y )/a in Eq. (4.108), under constraints (4.112), is (–1)n+m Tmn (x + a, y + b). Therefore, the contributions of the mentioned six pairs of virtual sources cancel each other if n + m is odd, so that there is no singularity in the denominator of expression (4.108), and consequently no eigenvalues. From this consideration we can conclude that formula (4.113) should be applied only for such combinations of m and n, when n + m is even. This is equivalent to 2π p 4π √ 2 2π √ 2 m + 3n2 = (n + 2l)2 + 3n2 = n + n l + l2 kmn = 3c 3c 3c (4.114) 4π √ 2 m + mn + n2 , ∀m, n = 0, 1, 2, . . . . ∼ kmn = 3c


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2) Isosceles right-angled triangle with the length of the leg equal to a: π√ 2 m + n2 . kmn = a

(4.115)

Of course, we should distinguish different eigenfunctions by their form. However, with the wave number defined by Eq. (4.115), the numerator in Eq. (4.108) becomes cos(πmx/a) × cos[πn(a – y )/a] = (–1)n cos(πmx/a) cos(πny/a) (recall that b = a). Since the sign “minus” does not influence the form of the eigenfunction, we can notice that the obtained structure is symmetric with respect to change of variable. So, Eq. (4.115) must be applied only under the following restriction: π√ 2 m + n2 , m ≥ n, n = 0, 1, 2, . . . . (4.116) kmn = a 3) Right-angled triangle with the acute angles equal to 30◦ and 60◦ and the hypotenuse c: 2π √ 2 3m + n2 . (4.117) kmn = 3c For the same reason as in case 1, the real array of eigenvalues here is: 4π √ 2 m + mn + n2 ∀m, n = 0, 1, 2, . . . . kmn = 3c

(4.118)

All three derived formulas (4.114), (4.116), (4.118) coincide with results of Makai (1970) obtained by an absolutely different method.

Helpful remarks 1◦ . It should be noted that the representation (4.108) is also a mode expansion of the solution (compare with results of Section 3.1). Analogously to the case of a layer of constant thickness, a few first terms of the series (the number of which depends on the value of the wave number k ) behave like an oscillating function. Then, beginning from m = [ak/π ] + 1, all terms decay exponentially. 2◦ . It is noteworthy that the double divergent series (4.102) has been transformed during our transformations into a single series which converges like a geometric progression. The constructed solutions give a representation of the Green’s function, which can be efficiently calculated without any problem. 3◦ . Formula (4.108) as representing the sum of divergent series (4.102) is worthy of special discussion. We can test the result obtained, for instance, by a boundary-element technique. In order to create wave field without singularities inside the closed space, we may remove the point-source, which leads to the nontrivial boundary condition: ∂H0(1) (kr ) ∂r ∂p =– = kH1(1) (kr ) , ∂n ∂n ∂n

1/2 r = (x – x0 )2 + (y – y0 )2 ,

(x, y ) ∈ l. (4.119) Then the Kirchhoff–Helmholtz integral formula determines the wave field at arbitrary point (x1 , y1 ) inside the boundary contour l, as follows: Z ∂r ∂r1 ki (1) (1) (1) (1)) H0 (kr1 )H1 (kr ) + u(x, y )H1 (kr1 ) dlxy , p(x1 , y1 ) = H0 (kr0 ) + 4 l ∂n ∂n 1/2 1/2 , r1 = (x – x1 )2 + (y – y1 )2 , r0 = (x1 – x0 )2 + (y1 – y0 )2 (4.120)


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which contains the unknown boundary value of the regular wave function u(x, y ) = p(x, y )– H0(1) (kr ), (x, y ) ∈ l. This function can be defined from the boundary integral equation given by the same Kirchhoff formula: Z ∂r1 u(x1 , y1 ) ki – u(x, y )H1(1) (kr1 ) dlxy 2 4 l ∂n Z (4.121) ∂r1 ki H0 (kr1 )H1(1) (kr ) dlxy , (x1 , y1 ) ∈ l. = 4 l ∂n

Recall that if using the BIE in such a form, we must set the mesh nodes so that no node coincides with any corner of the boundary contour l. The constructed algorithm for this regular Fredholm equation of the second kind was briefly described in Section 1.5. An example of the comparison between the exact explicit solution given by (4.108), (4.109) and direct numerical computations in the case of equilateral triangle is reflected in the table below, where N designates the number of nodes over the boundary line. Note that explicit formula (4.108) predetermines the wave function to be imaginary. The results of the numerical simulation by BEM show a nontrivial but very small real part when compared with the imaginary one. √ The calculated results are related to the case when k = 2, c = 1, x0 = 0.25, y0 = 0.25/ 3, y = 0.4. x

(8), (9)

–0.25 –0.20 –0.15 –0.10 –0.05 0.00 0.05 0.10 0.15 0.20 0.25

3.19 i 3.16 i 3.10 i 3.02 i 2.92 i 2.78 i 2.63 i 2.44 i 2.24 i 2.03 i 1.82 i

BEM (N = 300) BEM (N = 600) –0.0374 + 3.15 i –0.0378 + 3.12 i –0.0380 + 3.06 i –0.0381 + 2.97 i –0.0381 + 2.85 i –0.0379 + 2.71 i –0.0376 + 2.55 i –0.0371 + 2.37 i –0.0365 + 2.18 i –0.0357 + 1.99 i –0.0347 + 1.80 i

–0.0222 + 3.17 i –0.0225 + 3.14 i –0.0226 + 3.08 i –0.0227 + 2.99 i –0.0227 + 2.88 i –0.0227 + 2.73 i –0.0225 + 2.57 i –0.0222 + 2.39 i –0.0219 + 2.20 i –0.0214 + 2.01 i –0.0209 + 1.82 i

The results discussed in the present section are related to the author’s work (Sumbatyan et al., 2000). 4◦ . It can directly be verified that the explicit formulas (4.115), (4.118) for the particular geometries studied in this section justify not only the asymptotic estimate (4.94) universally valid for all domains but also Weyl’s hypothesis (4.87). The proof follows the one applied in Section 4.4 for rectangular domains (for more detail, see Makai, 1970).

4.6. Explicit Analytical Results for Some Polyhedra The method proposed in the previous section can be extended to some polyhedra. Here we also consider the Neumann problem only. In the class of polyhedral boundary surfaces S the only known classical geometry that gives an exact solution in explicit form is parallelepiped. The solution is well known and can be constructed by using a separation of variables, which involves again a modal representation. Since the latter does not admit extension to more complex geometries, we have developed an alternative approach. If a point source is placed at an arbitrary interior point (x 0 , y0 , z0 ) ∈ V : p0 = exp(ikr )/r with r = [(x0 – x)2 + (y0 – y )2 + (z0 – z )2 ]1/2 , then in order to satisfy the trivial boundary


Page 125

condition (1) for the normal derivative, one may arrange a three-dimensional periodic array of virtual sources (mirror images), which provides natural symmetry about the arbitrary boundary point. The response at the receiving point (x 1 , y1 , z1 ) ∈ V is then given by a three-dimensional series (0 < x < a; 0 < y < b; 0 < z < c is the size of the parallelepiped): p(x1 , y1 , z1 ) =

8 X

S (|ξj – x1 |, |ηj – y1 |, |ζj – z1 |, a, b, c),

(4.122)

j=1

where

(ξ1 , η1 , ζ1 ) = (x0 , y0 , z0 ); (ξ2 , η2 , ζ2 ) = (2a – x0 , y0 , z0 ); (ξ3 , η3 , ζ3 ) = (x0 , 2b – y0 , z0 ); (ξ4 , η4 , ζ4 ) = (x0 , y0 , 2c – z0 ); and the function S S (x, y , z , a, b, c) =

∞ X

m,n,l=–∞

(ξ5 , η5 , ζ5 ) = (2a – x0 , 2b – y0 , z0 ); (ξ6 , η6 , ζ6 ) = (2a – x0 , y0 , 2c – z0 ); (ξ7 , η7 , ζ7 ) = (x0 , 2b – y0 , 2c – z0 ); (ξ8 , η8 , ζ8 ) = (2a – x0 , 2b – y0 , 2c – z0 );

√ exp ik (x + 2am)2 + (y + 2bn)2 + (z + 2cl)2 √ (x + 2am)2 + (y + 2bn)2 + (z + 2cl)2

(4.123)

(4.124)

is given by a series, which, in the classical sense, does not converge. Therefore, the series (4.124) is absolutely unsuitable for any direct numerical treatment. Rapidly convergent representation for the function S . The series (4.124) can be regularized by means of Poisson–Abel summation (see Section 1.3). To follow his idea let us apply again the principle of extremely small attenuation. If a small imaginary component is added to the wave number: k ε = k + iε, 0 < ε 1, then the series (4.124) becomes exponentially convergent, and it has a finite limit when ε → 0. Since the function S (x, y , z , a, b, c) is periodic with respect to all its arguments, let us assume that 0 ≤ x < 2a, 0 ≤ y < 2b, 0 ≤ z < 2c. Then we apply the Weyl integral representation √ ZZ ∞ –i(α1 x+α2 y) 1 exp(ik x2 + y 2 + z 2 ) –γ(α) z e √ 2 = e dα1 dα2 , 2π γ (α) x + y2 + z 2 –∞ (4.125) q 2 2 2 α = (α1 , α2 ), γ (α) = α1 + α2 – kε .

Thus, the series (4.124) may be rewritten as # ZZ ∞ "X ∞ ∞ ∞ X 1 X –γ(α)(2cl+z) –γ(α)(2cl–z) e + e S (x, y , z , a, b, c) = 2π m,n=–∞ –∞ l=0 l=1

(4.126) exp[–i(α1 |2am + x| + α2 |2bn + y |)] dα1 dα2 . × γ (α) The branching function γ (α) (see Section 1.1) has a positive real part, Re[γ (α)] > 0, when ε > 0. Hence, both series in (4.126) are geometric progressions, which give ZZ ∞ ∞ exp[–γ (α)] + exp[γ (α)(z – 2c)] 1 X S (x, y , z , a, b, c) = 2π m,n=–∞ –∞ 1 – exp[–2cγ (α)] × =

exp[–i(α1 |2am + x| + α2 |2bn + y |)] dα1 dα2 γ (α) ZZ ∞ ∞ cosh[(c – z )γ (α)] exp[–i(α1 |2am + x| + α2 |2bn + y |)] 1 X 2π

m,n=–∞

–∞

sinh[cγ (α)]

γ (α)

(4.127) dα1 dα2 .


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Let us fix the value of the variable α1. Then the integrand in Eq. (4.127) is a meromorphic function of α2 , exponentially decaying at infinity. Therefore, with the help of the Jordan lemma, one can apply a standard expansion in residues at simple poles, by removing the integration contour down to infinity. The simple poles are solutions of the following transcendental equation: sinh[cγ (α1 , α2 )] = 0 ∼ γl = ∼ α2l = –iq (α1 , l),

πli cs

α21

q (α1 , l) =

+

πl c

2

(4.128) –

kε2 ,

l = 0, 1, . . . .

Hence, integration over α2 in (4.127) leads to the following representation: Z ∞ ∞ ∞ X X δl cos(πlz/c) dα1 exp(–iα1 |2am + x|) S (x, y , z , a, b, c) = c q ( α , l ) 1 –∞ m,n=–∞ l=0 × exp [–q (α1 , l) |2bn + y |] ,

δ0 =

1 , 2

δl = 1

(4.129)

(l = 1, 2, . . . ).

Since the real part of the root square function q (α 1 , l) in Eq. (4.129) is positive when ε > 0, the last series (taken over n) is again a geometric progression, its sum being equal to ∞ X n=0

exp[–q (α1 , l) (2bn + y )] +

∞ X

exp[–q (α1 , l) (2bn – y )]

n=1

h i p 2 2 – k2 cosh ( b – y ) α + ( πl/c ) 1 ε cosh[(b – y ) q (α1 , l)] h p i , = = sinh[b q (α1 , l)] sinh b α21 + (πl/c)2 – kε2

(4.130)

which involves

Z ∞ ∞ ∞ X X δl cos(πlz/c) S (x, y , z , a, b, c) = c –∞ m=–∞ l=0 i h p cosh (b – y ) α21 + (πl/c)2 – kε2 exp(–iα1 |2am + x|) h p i dα1 . × p α21 + (πl/c)2 – kε2 sinh b α21 + (πl/c)2 – kε2

(4.131)

The last integrand is a meromorphic function of the variable α 1 , which satisfies the necessary conditions of the Jordan lemma. Therefore, by removing the contour of integration down to infinity, the integral in (4.131) can be rewritten as an expansion in residues at simple poles. The latter are given as follows: q i h q πni 2 2 2 α21 + (πl/c)2 – kε2 = sinh b α1 + (πl/c) – kε = 0 ∼ b s (4.132) πn 2 πl 2 2 + – kε , ∼ α1 = –iqnl , qnl = b c

which yields the following representation in the form of a triple series:

∞ ∞ ∞ S (x, y , z , a, b, c) X X X δn δl cos(πny/b) cos(πlz/c) = exp(–qnl |2am + x|). (4.133) 2π b c q nl n=0 l=0 m=–∞


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Since Re(qnl ) > 0 when ε > 0, the last series with parameter m is a geometric progression: ∞ X m=0

exp[–qn,l (2am + x)] +

∞ X m=1

exp[–qn,l (2am – x)] =

cosh[(a – x) qn,l ] . sinh(a qn,l )

(4.134)

Therefore, the final representation (ε → +0) can be obtained from (4.133) and (4.134), as follows: ∞ X ∞ X δn δl cos(πny/b) cos(πlz/c) p S (x, y , z , a, b, c) = 2π b c (πn/b)2 + (πl/c)2 – k 2 n=0 l=0 i h p (4.135) cosh (a – x) (πn/b)2 + (πl/c)2 – k 2 h p i . × sinh a (πn/b)2 + (πl/c)2 – k 2 Note that with an arbitrary fixed wave parameter k the terms of the series (4.135) decay exponentially when n or/and l tend to infinity (recall that 0 < x < 2a ∼ –a < a – x < a). Thus, Eq. (4.135) is indeed an efficient rapidly convergent representation for the function S . It should also be noted that, although the original formula (4.134) is absolutely symmetric with respect to the coordinates x, y , z , the representation (4.135) has an asymmetric form. It thus admits a standard cyclic substitution: (x, a) → (y , b) → (z , c) → (x, a), with a final form under arbitrary substitution of this type being absolutely equivalent to (4.135).

Exact solution for triangular prisms. Apparently, in the case of a parallelepiped, the representation (1.435) can be directly obtained by using a separation of variables. However the application of this type of modal expansion is only possible for rectangular polyhedra. That is why in the previous section we developed a different approach, namely the virtual sources method, because this provides direct extension to more complex polyhedra, as soon as the complete disposition of the set of virtual sources is known. An appropriate exact solution can be constructed for polyhedra, for which every interior dihedral angle possesses a finite set of virtual sources, all situated outside of the polyhedron. In such cases an explicit solution can be obtained in the form of a finite superposition (compare with mode expansion for parallelepiped) of functions S with certain values of (ξj , ηj , ζj ) and (a, b, c). It is known that this holds only for wedges equal to π/n, n being an arbitrary positive integer. The case n = 2 for all dihedral angles has to be related to the above solution for the parallelepiped. The main statement of the present section is that one can construct an exact explicit solution for polyhedral spaces where each dihedral angle of the polyhedron is π/n. We do not aim to give a complete description of the all such polyhedra and the relevant solutions, but give instead some explicit results for three right-angled triangular prisms. If the base of the prism is a triangle, lying on the horizontal plane (see the previous section), then the corresponding sets of virtual sources are well known for 1) isosceles rightangled triangles; 2) equilateral triangles; and 3) right-angled triangle with acute angles 30 ◦ and 60◦ . Complete three-dimensional sets of virtual sources for every case are obtained as a certain periodic continuation (certain reflections) of the corresponding horizontal system along the vertical z -axis. Let (x0 , y0 , z0 ) denote the real point source, and (x1 , y1 , z1 ) the receiving point. Then for the three right-angled prisms we have: 1) The coordinates of the corners of this prism are (0, 0, 0); (a, 0, 0); (0, a, 0); (0, 0, h); (a, 0, h); (0, a, h) (a is a leg of the base). The exact solution is given as


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follows: p(x1 , y1 , z1 ) =

16 X

S (|ξj – x1 |, |ηj – y1 |, |ζj – z1 |, a, b, h),

b = a,

(4.136)

j=1

where (ξ , η , ζ )1 = (x0 , y0 , z0 );

(ξ , η , ζ )9 = (x0 , y0 , 2h – z0 );

(ξ , η , ζ )2 = (a – y0 , a – x0 , z0 );

(ξ , η , ζ )10 = (a – y0 , a – x0 , 2h – z0 );

(ξ , η , ζ )3 = (a + y0 , a – x0 , z0 );

(ξ , η , ζ )11 = (a + y0 , a – x0 , 2h – z0 );

(ξ , η , ζ )4 = (2a – x0 , y0 , z0 );

(ξ , η , ζ )12 = (2a – x0 , y0 , 2h – z0 );

(ξ , η , ζ )5 = (2a – x0 , –y0 , z0 );

(ξ , η , ζ )13 = (2a – x0 , –y0 , 2h – z0 );

(ξ , η , ζ )6 = (a + y0 , x0 – a, z0 );

(ξ , η , ζ )14 = (a + y0 , x0 – a, 2h – z0 );

(ξ , η , ζ )7 = (a – y0 , x0 – a, z0 );

(ξ , η , ζ )15 = (a – y0 , x0 – a, 2h – z0 );

(ξ , η , ζ )8 = (x0 , –y0 , z0 );

(ξ , η , ζ )16 = (x0 , –y0 , 2h – z0 );

(4.137)

and the function S in (4.136) is given by Eq. (4.135). 2) If the side of the equilateral √ base is c, then the corners of this prism √ are at the points (– 21 c, 0, 0); ( 12 c, 0, 0); (0, 21 3c, 0); (– 21 c, 0, h); ( 12 c, 0, h); (0, 12 3c, h). The exact solution is as follows: p(x1 , y1 , z1 ) =

24 X

S (|ξj – x1 |, |ηj – y1 |, |ζj – z1 |, a, b, h),

(4.138)

j=1

a=

3 2

c,

√ b = 21 3 c,

where (ξ , η , ζ )1 = (x0 , y0 , z0 );

√ √ √ 3y0 , 41 3c – 21 3x0 + 21 y0 , z0 ; √ √ √ (ξ , η , ζ )3 = 34 c – 21 x0 + 21 3y0 , 14 3c – 21 3x0 – 21 y0 , z0 ;

(ξ , η , ζ )2 =

3 c 4

– 21 x0 –

1 2

(4.139)

(ξ , η , ζ )4 = (ξ , –η , ζ )3 ; (ξ , η , ζ )5 = (ξ , –η , ζ )2 ; (ξ , η , ζ )6 = (ξ , –η , ζ )1 ; √ (ξ , η , ζ )j+6 = ξj + 32 c, ηj + 12 3c, ζj , j = 1, . . . , 6; (ξ , η , ζ )j+12 = (ξj , ηj , 2h – ζj ),

j = 1, . . . , 12.

3) If the hypotenuse √ of the base is c, then the Cartesian√coordinates of the prism corners are (0, 0, 0, ); (0, 12 3c, 0); (0, 21 c, 0); (0, 0, h); (0, 21 3c, h); (0, 21 c, h). The explicit form of the solution is p(x1 , y1 , z1 ) =

48 X

S (|ξj – x1 |, |ηj – y1 |, |ζj – z1 |, a, b, h),

a=

1 2

√ 3c,

b=

3 2

c, (4.140)

j=1


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where

(ξ , η , ζ )1 = (x0 , y0 , z0 ); √ √ √ (ξ , η , ζ )2 = 14 3c + 21 x0 – 12 3y0 , 34 c – 12 3x0 – 12 y0 , z0 ; √ √ √ (ξ , η , ζ )3 = 14 3c + 21 x0 + 12 3y0 , 43 c – 21 3x0 + 21 y0 , z0 ; √ √ (ξ , η , ζ )4 = 34 3c – 21 x0 – 21 3y0 , η3 , z0 ; √ √ (ξ , η , ζ )5 = 34 3c – 21 x0 + 12 3y0 , η2 , z0 ; √ 3c – x0 , y0 , z0 ; (ξ , η , ζ )6 = (ξ , η , ζ )j = (ξ , –η , ζ )13–j , j = 7, . . . , 12; √ (ξ , η , ζ )j+12 = (ξj + 21 3c, ηj + 23 c, ζj ), j = 1, . . . , 12; (ξ , η , ζ )j+24 = (ξj , ηj , 2h – ζj ), j = 1, . . . , 24.

(4.141)

It should be noted that if any value of (ξj , ηj , ζj ) in the above expressions appears to be outside the natural domain (0 ≤ ξj ≤ 2a, 0 ≤ ηj ≤ 2b, 0 ≤ ζ ≤ 2h), then one should use a periodicity of the function S , prior to applying formula (4.135). Explicit formulas for eigenfrequencies. The eigenvalues of the wave number k (related to resonance frequencies) can be obtained from Eq. (4.135) by setting its denominator equal to zero. Obviously, this holds for trivial values of the hyperbolic sine:   s πn 2 πl 2 + – k2  = 0 sinh a b c (4.142) s r πn 2 πl 2 2 2 2 n l m + – k 2 = πmi ∼ kmnl = π + 2 + 2, ∼ a b c a2 b c

with arbitrary natural numbers m, n, l. For a parallelepiped domain with sides a, b, c this result is well known in the literature. For triangular prisms it involves some new exact estimates: 1) Right-angled prism with the base in the form of an isosceles right-angled triangle (the leg of the base is a, the height of the prism is h). The eigenvalues of the frequency parameter are r l2 m2 + n 2 + 2, (4.143) kmnl = π 2 a h the same numbers as for a right-angled prism with a square base. 2) Right-angled prism with the base in the form of an equilateral triangle (the side ofpthe base is c, the height of the prism is h). Following Eq. (4.142), we obtain k mnl = π 4(m2 + 3n2 )/(9c2 ) + l2/h2 . Let us prove that this expression holds only if the sum (m + n) is even (compare with Section 4.5). We start from the evident notice that Eqs. (4.138), (4.139) show: for every point (ξ , η , ζ )j , j = 1, . . . , 24 in the sum (4.138) there p is also a point (ξ , η , ζ ) j±6 with ξj±6 = ξj ± a, ηj±6 = ηj ± b, ζj±6 = ζj . Now if we take (πn/b)2 + (πl/c)2 – k 2 = πmi/a, as in Eq. (4.142), part of the numerator in Eq. (4.135) becomes equal to p the following cos(πny/b) cosh (a – x) (πn/b)2 + (πl/c)2 – k 2 = cos(πny/b) cos[πm(a – x)/a], the term which has opposite signs for (ξ , η , ζ )j and (ξ , η , ζ )j±6 if (m + n) is odd. Therefore, these terms in the sum (4.138) cancel each other out whenever m + n is odd, which excludes this combination of n and m from the set of eigenvalues of the parameter k . The proof is


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complete. Further, in the same way as in the previous section, it can also be proved that using the obtained formula for kmnl with (m + n) even is equivalent to the following result: r 16(m2 + mn + n2 ) l2 + 2, (4.144) kmnl = π 9c2 h if one applies this with arbitrary positive integers m, n, l. 3) Right-angled prism with a right-angled triangular base having the acute angles 30 ◦ and 60◦ , the base hypotenuse is c and the height of thepprism h. Equation (4.142) involves the following eigenvalues of the wave number kmnl = π 4(3m2 + n2 )/(9c2 ) + l2/h2 . However, arguments analogous to those given for case 2) prove that the correct result is (m, n, l are arbitrary natural numbers): r 16(m2 + mn + n2 ) l2 + 2, (4.145) kmnl = π 9c2 h the same values as for the previous prism, given by (4.144). The proposed results are connected with the author’s works Sumbatyan (2000); Sumbatyan and Pompei (2001).

Helpful remarks 1◦ . The analytical results obtained are worthy of a special test to compare them with numerical calculations. To calculate the wave field at an arbitrary receiving point in the nonresonance case, one can apply, for instance, a finite-element or boundary-element technique. However, more precise results are obtained numerically for eigenfrequencies, since they are an integral measure of the polyhedron’s geometry. We have used a finite element method for this purpose, and we have achieved the relative error in comparison between analytical and direct numerical calculations less than 1% for all considered geometries. 2◦ . Weyl–Carleman theory (see Section 4.4) can also be applied in the 3D case to a volume of arbitrary shape. It can be strictly proved that the leading asymptotic term, for both the Dirichlet and the Neumann problem, is expressed as follows (see Courant and Hilbert, 1953): V 3/2 λ + O (λ ln λ), λ → ∞, (4.94) N (λ) = 6π 2 where V is the measure of the volume.


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Chapter 5

Integral Equations in Diffraction by Linear Obstacles 5.1. Integral Operators in Diffraction by Linear Screen and by a Gap in the Screen As in wave problems for acoustic layer (Chapter 3), the problem considered here demonstrates the power of the Fourier transform. Let a plane acoustic wave fall from below, propagating in direction of vertical axis y in the Cartesian coordinate system (x, y ): pinc (x, y ) = eiky . The problem is thus considered as two-dimensional. We will study here integral equations arising in diffraction by an acoustically hard linear obstacle placed over the axis x, i.e., somewhere at y = 0. If the obstacle occupies the whole axis x: y = 0, –∞ < x < ∞, then the problem is one-dimensional, and is reduced to a simple ordinary differential equation with constant coefficients. If there is a finite-length obstacle, then we consider simultaneously the two problems genetically related to each other: α) There is a rigid screen placed on |x| ≤ a, y = 0. β ) There is a gap (hole) in the infinite hard screen, occupying the interval |x| ≤ a, y = 0. In Hönl et al. (1961) you can find some general results, which establish a relationship between these two problems. For both problems we consider separately the lower (y ≤ 0, all functions here are marked with the subscript 1) and the upper (y ≥ 0, with the subscript 2) half-planes with appropriate (complex-valued) acoustic pressure p1 (x, y ) and p2 (x, y ), with both of them satisfying the Helmholtz equation ω (5.1) k= . ∆p1,2 + k 2 p1,2 = 0, c Normal incidence of a plane wave is considered here only for simplicity. Another form of the incident wave modifies the right-hand side of integral equations only. If we apply the Fourier transform with respect to variable –∞ < x < ∞, then it can be easily seen that a solution whose diffracted component satisfies Sommerfeld’s radiation condition is given by (cf. Chapter 3) α) P2 (s , y ) = A(s )e–γ(s)y + P inc (s , y ),

y ≥ 0,

+ P (s , y ),

y ≤ 0,

β ) P2 (s , y ) = C (s )e–γ(s)y , y ≥ 0, P1 (s , y ) = D (s )eγ(s)y + P inc (s , y ),

y ≤ 0,

P1 (s , y ) = B (s )e

γ(s)y

inc

γ (s ) =

√ α2 – s 2 ,

(5.2) (5.3)

where P inc (s , y ) = 2πδ (s )eiky . The coefficients A(s ) and B (s ) are to be defined from the boundary conditions on the line y = 0, which are posed in accordance with the impenetrability of the hard screen, i.e., trivial normal component of the velocity. Since


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v = (1/iωρ0 ) grad p, this involves the two problems in hand, for y = 0: α)

∂p1 ∂p2 = = 0, ∂y ∂y

β)

p 1 = p2 ,

|x| ≤ a,

∂p1 ∂p2 = , ∂y ∂y

∂p1 ∂p2 = , ∂y ∂y

|x| > a.

(5.4)

∂p1 ∂p2 = = 0, ∂y ∂y

|x| > a,

(5.5)

p 1 = p2 ,

|x| ≤ a,

where the second lines in Eqs. (5.4), (5.5) mean continuity of the wave field in the fluid. We aim to reduce both problems α) and β ) to an integral equation over the interval |x| ≤ a. For this purpose, let us introduce the new unknown functions g α (x) and gβ (x) defined just on this interval: 0, |x| > a α) y = 0: p1 – p2 = (5.6) gα (x), |x| ≤ a. β ) y = 0:

∂p2 ∂p1 = = ∂y ∂y

0, gβ (x),

|x| > a |x| ≤ a.

(5.7)

The boundary conditions in Eqs. (5.4), (5.5) are written so that in both problem α) and β ) we have ∂p1 /∂y = ∂p2 /∂y , y = 0, |x| < ∞, which with the help of Eqs. (5.1), (5.2) yields α) – γA = γB (5.8) β ) – γC = γD + 2πikδ (s ). Finally, Eqs. (5.6), (5.7) in Fourier transforms imply α) B – A = Gα (s ) β ) – γC = Gβ (s ),

where Gα,β =

Z

(5.9)

a

gα,β (ξ ) eisξ dξ .

(5.10)

–a

Expressions (5.8)–(5.10) allow us to write out coefficients A, B and C , D in terms of Gα , G β : Gα (s ) Gα (s ) , A=– , α) B = 2 2 (5.11) Gβ (s ) ik Gβ (s ) , D= – 2π δ (s ). β) C = – γ (s ) γ (s ) γ (s ) Therefore by applying the convolution theorem (see Section 1.1) we arrive at the following representations: Z ∞ 1 inc B (s )eγy e–isx ds α) p1 (x, y ) = p (x, y ) + 2π –∞ Z a Z ∞ 1 iky (5.12) gα (ξ ) dξ eγy e–is(x–ξ) ds , y ≤ 0, |x| < ∞, =e + 4π –a –∞ Z a Z ∞ 1 gα (ξ ) dξ e–γy e–is(x–ξ) ds , y ≥ 0, |x| < ∞. p2 (x, y ) = eiky – 4π –a –∞


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Z ∞ 1 C (s )e–γ(s)y e–isx ds p2 (x, y ) = 2π –∞ Z a Z ∞ –γ(s)y e 1 gβ (ξ ) dξ e–is(x–ξ) ds , y ≥ 0, |x| < ∞, =– 2π –a –∞ γ (s ) Z ∞ 1 inc (5.13) D (s )eγ(s)y e–isx ds p1 (x, y ) = p (x, y ) + 2π –∞ Z a Z ∞ γy Z ∞ γy e e 1 gβ (ξ ) dξ e–is(x–ξ) ds – ik δ (s )e–isx ds = eiky + 2π –a γ ( s ) γ ( s ) –∞ Z a Z–∞ ∞ γy e 1 gβ (ξ ) dξ e–is(x–ξ) ds + e–iky , y ≤ 0, |x| < ∞. = eiky + 2π –a –∞ γ (s ) In the last transformations we have used the basic property of Dirac’s delta (see Section 1.4), and the value γ (0) = –ik . The required integral equations are obtained from (5.12), (5.13) if we attract the remaining boundary condition, which has not yet been used: Z a ∂p1 gα (ξ )Kα (x – ξ ) dξ = –2ik , (x ≤ a), = 0, |x| ≤ a ∼ α) ∂y y=0 –a (5.14) Z ∞ 1 isx Lα (s ) e ds , Lα (s ) = γ (s ), Kα (x) = 2π –∞ Z a β ) p1 = p2 , y = 0, |x| ≤ a ∼ gβ (ξ )Kβ (x – ξ ) dξ = –1, (|x| ≤ a), –a Z ∞ (5.15) 1 1 Lβ (s ) eisx ds , Lβ (s ) = . Kβ (x) = 2π –∞ γ (s ) Properties of integral equations. Since Lα (s ) is unbounded for large s , the integral (5.14) for the kernel K α should be treated in the generalized sense as a value of divergent integral. This can be calculated following to ideas of Section 1.3: Z Z 1 ∞ √ 2 1 ∞ s cos(sx) ds + s – k 2 – s cos (sx) ds Kα (x) = π 0 π 0 Z ∞ Z k 2 ∞ cos(sx) ds 1 –εs √ lim e s cos(sx) ds – = π ε→+0 0 π 0 s2 – k2 + s (5.16) Z ∞ 2 2 2 ε –x k cos(kxs ) ds 1 √ lim – = π ε→ +0 (ε2 + x2 )2 π 0 s2 – 1 + s Z ∞ 2 k Ic (kx) cos(kxs ) ds 1 √ . , Ic (kx) = =– 2 – πx π s2 – 1 + s 0 The second term in (5.16) is bounded for finite x and k , because this integral is convergent in the classical sense if k , x > 0. For kx → 0 its asymptotic behavior can be estimated from formula (1.83): Ic (kx) ∼ O ln(k |x|) , kx → 0. Therefore the kernel in problem α) is a characteristic hyper-singular kernel plus a kernel with a (weak) logarithmic singularity. The kernel Kβ (x) can also be estimated in the same way; however it is more convenient to operate with the explicit representation β)

Kβ (x) =

i (1) H (k |x|), 2 0

(5.17)


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which is a kernel with a weak logarithmic singularity (cf. the asymptotic behavior of the Hankel function for small arguments, see Eq. (2.69)): i 1 x → 0. (5.18) Kβ (x) ∼ – [ ln(k |x|) + γ – ln 2] + + O x2 ln |x| , π 2 THEOREM 1. Integral equation (5.14) is uniquely solvable in the class of bounded functions at least for sufficiently small k . Proof. Let us extract, according to Eq. (5.16), characteristic hyper-singular part of the kernel Z Z k2 a 1 a gα (ξ ) dξ = –2ik + gα (ξ ) Ic (k |x – ξ |) dξ , |x| ≤ a, (5.19) – π –a (x – ξ )2 π –a and apply inversion formula (1.190) for the operator of the left-hand side. # "√ Z a Z ξ dξ a2 – x2 a √ 2 2 + Kc (x, ξ ) gα (ξ ) dξ , gα (x) = 2ik π a – ξ (x – ξ ) (5.20) –a –a

Kc (x, ξ ) = O (k ln k ), k → 0, where the estimate of the kernel Kc is uniform with respect to (x, ξ ) ∈ (–a, a). The first integral here can be calculated as follows Z a Z a Z a (ξ – x) + x dξ ξ dξ √ 2 2 √ 2 2 √ 2 2 = dξ = –π , since = 0, a – ξ (x – ξ ) a – ξ (x – ξ ) a – ξ (x – ξ ) –a –a –a (5.21) as the Cauchy principal value. We thus have arrived at the Fredholm integral equation of the second kind, which, due to the remark right after Theorem 5, Section 1.5, is uniquely solvable. This completes the proof. It should be noted that the solvability question for the arising equation (5.14) is a rather abstract issue. A much more practical question is related to the uniqueness, provided that correct conditions are set at infinity. If we have constructed a correct solution of this linear problem, it must be unique. Note that √ a bounded solution vanishes near the edges of the interval (–a, a) as, for example, a ± x, x → ∓a. THEOREM 2. For all 0 < k < ∞, the solution of equation√(5.14) is unique in the class of differentiable functions: gα (x) ∈ C1 (–a, a), with gα (x) ∼ a ± x, x → ∓a. Proof. This statement can be proved by the method of Section 1.6 adopted to the case when the symbolic function L(s ) is nonpositive (in our problem it is complex-valued). If there are two solutions u1 (x) and u2 (x) of the class described in the conditions of the theorem, then for the considered operator equation G α gα = fα we have Gα (u1 – u2 ) = 0. The scalar product of this relation with the function (u 1 – u2 ) results in Z ∞ √ (5.22) Lα (s )|U1 (s ) – U2 (s )|2 ds = 0, Lα (s ) = s 2 – k 2 . –∞

It is clear that under the conditions of this theorem U 1 (s ), U2 (s ) ∼ |s |–3/2 , s → ∞ being continuous for finite s , the integral (5.22) is finite. Separation of real and imaginary parts in relation (5.22) results in Z –k Z ∞ √ Z k√ 2 2 2 s – k |U1 (s ) – U2 (s )| ds – i k 2 – s 2 |U1 (s ) – U2 (s )|2 ds = 0. (5.23) + –∞

k

–k

Both the real and the imaginary terms in (5.23) should be equal to zero, which is possible only if U1 (s ) ≡ U2 (s ), |s | < ∞ ∼ u1 (x) ≡ u2 (x), |x| ≤ a. This proves the theorem. Equation (5.15) can be studied by absolutely the same technique.


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THEOREM 3. 1) The integral equation (5.15) is uniquely solvable in the class of continuous functions with a square root singularity at the ends of the interval at least for sufficiently small k , and 2) its solution is unique in the considered class for all k . Proof. The inversion of an integral operator with the characteristic logarithmic kernel (see Section 1.7) reduces equation (5.15) to a Fredholm equation of the second kind. Further procedure to prove the first statement of the theorem is absolutely the same as for the previous theorem. However, the second statement is also proved by analogy with Theorem 2. We have, instead of Eq. (5.22), Z ∞ 1 . (5.24) Lβ (s )|U1 (s ) – U2 (s )|2 ds = 0, Lβ (s ) = √ s 2 – k2 –∞ For functions u1 (x), u2 (x) from the considered class the separation of the real and imaginary parts in (5.24), Z –k Z ∞ Z k |U1 (s ) – U2 (s )|2 |U1 (s ) – U2 (s )|2 √ √ ds + i ds = 0, (5.25) + s2 – k2 k2 – s 2 –k –∞ k leads to some finite-valued integrals. If u1 , u2 ∈ C1 (–a, a) with behavior (a±)–1/2 , x → ∓a, the asymptotics of U1 , U2 is U1 (s ), U2 (s ) ∼ s –1/2 , s → ∞; therefore the first two integrals converge at infinity, which completes the proof.

Helpful remarks For applications, the theory discussed in this section permits the calculation of the far-field scattering pattern. At low frequencies this can be expressed in explicit form (compare with the results of Chapter 2). For the sake of brevity, we give here only the sketch. It is clear that for small frequencies (k → 0) the kernels (5.16), (5.18) of the integral operators (5.14), (5.15) are reduced to their characteristic parts, which are characteristic hyper-singular kernel and characteristic logarithmic one. Both of them admit exact √analytical inversion. As follows from Eq. (5.20) in the first problem we have g α (x) = –2ik a2 – x2 . The analogous low-frequency solution√ in the second problem, as a solution to a logarithmic characteristic equation, is gβ (x) = B/ a2 – x2 , with some constant B . Then this function may be substituted into Eqs. (5.12), (5.13) giving wave field in all space. The far-field asymptotics can be extracted from there if we estimate integrals Z ∞ Z ∞ –γ|y| e e–is(x–ξ) ds , (5.26) Iα = e–γ|y| e–is(x–ξ) ds , Iβ = γ ( s ) –∞ –∞ p for large R = x2 + y 2 . This is given by the stationary phase method (see Eq. (1.88), (1.92)): r Z ∞ p 2πk cos α ei(kR–π/4) F (k sin α), R = x2 + y 2 → ∞, J= F (s ) e–isx e–γ(s |y| ds ∼ R –∞ (5.27) hence r p 2πk i(kr–π/4) e cos α, r = (x – ξ )2 + y 2 → ∞, (5.28) Iα ∼ r r r p 2πk i(kr–π/4) cos α 2π i(kr+ π/4) e = e , r = (x – ξ )2 + y 2 → ∞. (5.29) Iβ ∼ r γ (–k sin α) kr © 2005 by CRC Press LLC

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Further treatment is based upon the classical far-field approximation for the quantity r in the argument of the exponential functions in (5.28), (5.29) (cf. Chapter 2): r p xξ 1 2xξ – ξ 2 2 2 2 ≈R 1– 2 +O r = x + y – 2xξ + ξ = R 1 – 2 R R R2 (5.30) p 1 2 2 , R = x + y → ∞. = R – ξ cos α + O R

5.2. Operator Equation in Diffraction Problem on a Crack in Unbounded Elastic Medium Here we will study an integral operator arising in diffraction by a crack in elastic medium. Again, only for simplicity, we consider the 2D case of the normal incidence of a plane longitudinal wave. Let a crack of length 2a be situated over the interval |x| ≤ a, y = 0, and the plane longitudinal wave fall to this crack from y = –∞: ϕ inc (x, y ) = eikp y , ψ inc (x, y ) = 0. The governing equations to this problem can be found in Section 1.9, and the application of the Fourier transform in the variable –∞ < x < ∞ is explained in Section 3.3. Let us consider again the two half-planes y ≤ 0 (labeled with the superscript (1)) and y ≥ 0 (superscript (2)) separately. Then in both half-planes the full wave field is the sum of the incident and scattered ones: ϕ = ϕinc + ϕsc ,

ψ = ψ sc

(ψ inc = 0).

(5.31)

In terms of Fourier images, a solution satisfying a correct radiation condition at infinity, which in the considered case of unbounded space can be applied simply as Sommerfeld’s radiation condition, has the following form (compare with Section 3.3): q –γ(s)|y| Φ1,2 (s , y ) = A1,2 (s ) e , γ (s ) = s 2 – kp2 , (5.32) p Ψ1,2 (s , y ) = B1,2 (s ) e–q(s)|y| , q (s ) = s 2 – ks2 ,

where both Lamé potentials are taken for the scattered component of the elastic wave field. Of course, the coefficients A1,2 , B1,2 will be found from the boundary conditions at the line y = 0; in the case of load free crack faces, these conditions are written out as y = 0:

(1,2) inc σxy = –σxy , (1) σxy

=

(2) σxy ,

(1,2) inc σyy = –σyy , (1) σyy

=

(2) σyy ,

u(1) x

|x| ≤ a, = u(2) x ,

(2) u(1) y = uy ,

|x| > a,

(5.33)

where the equations of the second line here mean the continuity of stresses and displacements. (1) (2) (1) (2) As usual, we notice that conditions (5.33) imply σxy = σxy , σyy = σyy , on the full line y = 0, |x| < ∞; hence ∂ 2 ψ1 ∂ 2 ψ1 ∂ 2 ϕ2 ∂ 2 ψ2 ∂ 2 ψ2 ∂ 2 ϕ1 + – =2 + – , 2 2 ∂x∂y ∂y ∂x ∂x∂y ∂y 2 ∂x2 2 2 ∂ 2 ψ1 ∂ ϕ2 ∂ 2 ψ2 ∂ ϕ1 2 + + k s ϕ1 = 2 + + ks2 ϕ2 , 2 ∂x2 ∂x∂y ∂x2 ∂x∂y

2

that is

– 2isΦ01 + Ψ001 + s 2 Ψ1 = –2isΦ02 + Ψ002 + s 2 Ψ2 , 2(–s 2 Φ1 – isΨ01 ) + ks2 Φ1 = 2(–s 2 Φ2 – isΨ02 ) + ks2 Φ2

(y = 0),

(5.34)

(5.35)


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where the sign of the derivative is related to y . Equations (5.34) imply – 2isγA1 + (2s 2 – ks2 )B1 = 2isγA2 + (2s 2 – ks2 )B2 ,

(5.36) – (2s 2 – ks2 )A1 – 2isqB1 = –(2s 2 – ks2 )A2 + 2isqB2 . In order to reduce the problem to an integral equation, let us introduce the new unknown functions 0, |x| > a, 0, |x| > a, (1) (2) (1) (2) y = 0: u x – ux = uy – u y = (5.37) gy (x), |x| ≤ a. gx (x), |x| ≤ a, Then in Fourier variables Z a

(–isΦ1 + Ψ01 ) – (–isΦ2 + Ψ02 ) = Gx (s ),

Gx (s ) =

(Φ01

Gy (s ) =

+ isΨ1 ) –

(–Φ02

+ isΨ2 ) = Gy (s ),

Z –aa –a

gx (ξ )eisξ dξ ,

(5.38)

gy (ξ )e

isξ

dξ ,

which is equivalent to –is A1 + qB1 + isA2 + qB2 = Gx , γA1 + isB1 + γA2 – isB2 = Gy . (5.39) Now the solution of the system of equations (5.35), (5.38) allows us to express the four coefficients A1 , B1 , A2 , B2 in terms of the functions Gx , Gy . is is 2s 2 – ks2 2s 2 – ks2 G + G , B = B = – G – Gx . (5.40) A1 = –A2 = – y x 1 2 y 2γks2 ks2 ks2 2qks2 Then the Fourier images of the normal and tangential stress over the crack faces are ∆(s ) ∆(s ) (1) (1) Gy , σ exy (s , 0) = – Gx , (5.41) σ eyy (s , 0) = 2 2γks 2qks2 where ∆(s ) = (2s 2 – ks2 )2 – 4s 2 γq (5.42) is the well-known Rayleigh function. Inversion of relations (5.41), with the help of the last remaining boundary conditions written in the first line of Eq. (5.33), leads to the following integral equations for the functions gx (x) and gy (x): Z a 2σ inc (x, 0)k 2 (|x| ≤ a), gy (ξ )Ky (x – ξ ) dξ = – yy 2 s ρcs –a (5.43) Z ∞ ∆(s ) 1 –isx Ly (s )e ds , Ly (s ) = , Ky (x) = 2π –∞ γ (s ) Z a inc 2σxy (x, 0)ks2 (|x| ≤ a), gx (ξ )Kx (x – ξ ) dξ = ρc2s –a (5.44) Z ∞ ∆(s ) 1 –isx Lx (s )e ds , Lx (s ) = . Kx (x) = 2π –∞ q (s ) inc inc Since σyy (x, 0) = (ks2 – 2kp2 ), σxy (x, 0) = 0, so the function gx is trivial: gx (x) ≡ 0, |x| ≤ a. Thus the only nontrivial function here is the relative opening of the crack’s faces in (2) the vertical direction gy (x) = (u(1) y – uy )(x, 0), |x| ≤ a, which after the change of variable s = es ks is reduced to the form (tildes are omitted below ) Z a kp 2(1 – 2β 2) , |x| ≤ a β= 0.

Proof. As usual, if there are two different solutions u 1 (x), u2 (x) of the mentioned class, then u(x) = u1 (x)– u2 (x) is a solution of the homogeneous equation (5.43), which is another form of (5.45). If we apply the scalar product of both sides of (5.43) with the same function u(x), we arrive at equality Z ∞ Ly (s )|U (s )|2 ds = 0. (5.48) –∞

The left-hand integral here is finite since the behavior of the integrand at infinity is ∼ s (s –3/2 )2 = O (1/s 2 ). Further steps of the proof are based on the classical properties of the Rayleigh function ∆(s ) (see, for example Achenbach, 1973). Usually, the Rayleigh equation ∆(s ) = 0 ∼ (2s 2 – ks2 ) – 4s 2 γq = 0

(5.49)

s > s0 : α < 0, β = 0; ks < s < s0 : α > 0, β = 0; kp < s < ks : α > 0, β > 0; 0 < s < kp : α = 0, β > 0;

(5.50)

has one real root s0 such that s0 > ks > kp . Further, the function Ly (s ) = ∆(s )/γ (s ) of (5.45), present in (5.48), is complex-valued: Ly (s ) = α + iβ , with

so separation of real and imaginary parts in (5.48) does not result in U (s ) ≡ 0, because of negative sign of the symbolic function for s > s0 .


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In order to provide that the Fourier image of the kernel is positive for s > s 0 , keeping positive signs for its real and imaginary parts on all other intervals, we apply the idea of Vorovich and Babeshko (1979). We develop below this approach to our case of hyper-singular equations. Let us represent equation (5.43) in the equivalent form: Z a Z ∞ 1 1 gy (ξ ) dξ (s02 – s 2 )Ly (s ) 2 2 e–is(x–ξ) ds = f , 2π –a s0 – s –∞ (5.51) 2 2 2(ks – 2kp ) 2 ks , |x| ≤ a, f = – µ that is

1 2π

Z

∞

Z

∞

v (ξ ) dξ (s02 – s 2 )Ly (s ) e–is(x–ξ) ds = f , ∞ Z a ∞ Z ∞ –is(x–ξ) e 1 gy (ξ ) dξ ds , v (x) = 2 2 2π –a –∞ s0 – s

|x| ≤ a, (5.52)

where the last integral is treated as a singular integral of the Cauchy type, which can be calculated as the sum of contributions by simple poles s = ±s 0 . We thus notice that Gy (s ) (5.53) V (s ) = 2 2 , s0 – s and so v (x) is even for all |x| < ∞ if Gy (s ) is even (the latter is even since gy (x) is even). Then is0 |x–ξ| Z Z e e–is0 |x–ξ| 1 a i a gy (ξ ) + dξ = gy (ξ ) sin(s0 |x – ξ |) dξ . (5.54) v (x) = – 2 –a 2s0 –2s0 2 –a Now we note that for |x| > a Z 1 a gy (ξ ) sin[s0 (x – ξ )] dξ v (x) = ± 2 –a Z a Z a (5.55) 1 sin(s0 x) gy (ξ ) cos(s0 ξ ) dξ + cos(s0 x) gy (ξ ) sin(s0 ξ ) dξ = 0, =± 2 –a –a since the second integral here is zero due to evenness of g y (x), and the first one must be trivial due to evenness of v (x) ∀x ∈ (–∞, ∞). This proves that Eq. (5.52) is a standard integral equation (not of Fredholm type), when there is the same interval of variation for interval (ξ ) and external x variables. After this auxiliary transformation the proof is completed by our usual reducing to relation of the type (5.48), which now reads Z ∞ (s02 – s 2 )Ly (s )|U (s )|2 ds = 0, (5.56) –∞

where the symbolic function (s02 – s 2 )Ly (s ) has positive real and imaginary parts over the total axis |s | < ∞. Then separation of the real and imaginary parts finally proves the Theorem.


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Helpful remarks As in the scalar problem of diffraction by a linear scatterer, the developed method allows us for low frequencies to obtain the far-field scattering pattern of the linear crack of finite length. Briefly speaking, the calculation scheme is as follows. For example, for the longitudinal scattered wave we have Z a Z ∞ 2 2s – ks2 –γ|y| –is(x–ξ) 1 g ( ξ ) dξ e e ds ϕ1,2 (x, y ) = ∓ y 2πks2 –a 2γ –∞ s (5.57) Z a 2π i(kR+π/4) 2 1 2 2 –ikp ξ cos α e (2kp cos α – ks ) gy (ξ )e dξ , ∼ ∓ 4πks2 kp R –a and the problem is then reduced to the calculation of the arisen integral, which was discussed in detail above. The scattering diagram of the transverse wave can be constructed in the same way.

5.3. High-Frequency Asymptotics in Diffraction by Linear Obstacles in Unbounded Medium The integral equations (5.14), (5.15) arising in diffraction by linear obstacles, as a Fredholm equation of the first kind with a weak singularity in the kernel, and a hyper-singular equation, admit the application of direct numerical methods (see Chapter 9) for low and moderate frequencies. For extremely high frequencies, when the wavelength λ = 2π/k becomes too short, the application of any numerical method requires very large number of nodes, since stable computations with reliable results need at least 10 nodes per wavelength. This is why the high-frequency (i.e., short-wave) regime requires a special analytical asymptotic approach. In our investigation we follow ideas suggested by Koiter (1954) and Aleksandrov (1968). Let us rewrite equations (5.14), (5.15) in a uniform manner as follows (we apply here ˜ , and tildes are omitted below) the change of variables s = s˜ k , x = x/k ˜ , ξ = ξ/k Z ak g (ξ )K (x – ξ ) dξ = f , |x| ≤ ak (f = const), –ak √ ( (5.58) Z ∞ α ) L ( s ) = γ ( s ) = s 2 – 1, f = –2i, 1 isx L(s ) e ds , L(s ) = K (x) = √ 2π –∞ β ) L(s ) = 1/γ (s ) = 1/ s 2 – 1, f = –k . Let us represent Eq. (5.58) in equivalent form, where the key point is a representation of the integral over finite-length interval as a combination of several integrals, with each of them being taken over some semi-infinite integral. In order to derive this representation, let us consider the three equalities Z ∞ Z –ak v (ak + ξ )K (x – ξ ) dξ = f + [v (ak – ξ ) – w]K (x – ξ ) dξ , –ak ≤ x < ∞, Z–aa Z –∞ ∞ v (ak – ξ )K (x – ξ ) dξ = f + [v (ak + ξ ) – w]K (x – ξ ) dξ , –∞ < x ≤ ak , –∞ ak Z ∞ w(ξ )K (x – ξ ) dξ = f , –∞ < x < ∞, –∞

(5.59) with each of them holding, among other domains, also on the basic interval |x| ≤ ak .


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Note that a solution of the last equation is simply constructed by the direct application of the Fourier transform: α) 2, f δ (s ) ∼ w(x) = ≡ const = (5.60) W (s )L(s ) = 2πf δ (s ) ∼ W (s ) = 2πf β ) ik , L(s ) L(0) and for this reason we could write in Eqs. (5.59) the function w(x) as a constant w. After this remark, let us take the sum of the first two equations (5.59) and subtract this sum by the third equation (5.59). By cancelling some identical integrals, we can see that g (x) = v (ak + x) + v (ak – x) – w

(5.61)

satisfies initial equation (5.58). Note that the (constant) function w is already found in Eq. (5.60), so the only remaining unknown function to be found is v (x). The latter satisfies Eq. (5.59) which in a rewritten form is Z ∞ Z ∞ v (ξ )K (x – ξ ) dξ = f + [v (2ak + ξ ) – w]K (x + ξ ) dξ , 0 < x < ∞. (5.62) 0

0

The right-hand integral here represents a tail, and we will prove that this is asymptotically small when k → ∞, and so Eq. (5.62) becomes a Wiener–Hopf equation, which admits solution in explicit form (see Section 1.2). A rather widespread mathematical approach consists in accepting first this statement as a hypothesis, to construct the full solution under this hypothesis, and finally to make sure that the obtained solution validates this hypothesis. If we reject the tail integral in (5.62), then the solution of the Wiener–Hopf equation Z ∞ v (ξ )K (x – ξ ) dξ = f , 0 ≤ x < ∞, (5.63) 0

can be derived by the method described in Section 1.2 as follows: √ f 2 + F– (s ) α) V+ (s ) s – 1 = –is + √ √ f + F– (s ) ∼ V+ (s )( s + 1)+ ( s – 1)– = –is + √ F– (s ) 1 f + √ + √ ∼ V+ (s )( s + 1)+ = –is + s – 1 – ( s – 1)– √ 1 f + N– (s ) ∼ V+ (s )( s + 1)+ = –i –is + √ 1 f = N– (s ) ≡ 0 ∼ V+ (s )( s + 1)+ – –i –is + f 1 f f 1 √ √ = eπi/4 √ = , ∼ V+ (s ) = –i (–is ) s + 1 –i p 1 + ip p p–i

(5.64)

p = –is ,

where we have passed from the Fourier transform variable s to the Laplace transform variable p, in order to take advantage of the tables of inverse Laplace transforms (Bateman and Erdelyi, 1954): √ √ f v (x) = eπi/4 √ Erf –ix = eπi/4 f 2 [C (x) + iS (x)], –i

(5.65)


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where Erf(x) is the probability integral (error function) and C (x), S (x) are the Fresnel integrals (Bateman and Erdelyi, 1953). Now, having constructed a solution of equation (5.63) in explicit form, we can prove that the right-hand integral in (5.62) is asymptotically small. Indeed, since 1 x → +∞, (5.66) C (x), S (x) = + O (x–1/2 ), 2 we can see that v (x) = if + O (x–1/2 ) = 2 + O (x–1/2 ) = w + O (x–1/2 ),

so expression in the square brackets in (5.62) is

v (2ak + ξ ) – w = O (ak )–1/2 ,

x → +∞,

k → ∞.

(5.67) (5.68)

If the kernel K (x) possesses integrable behavior at infinity, then the estimate (5.68) takes place for the full right-hand integral (5.62). We have Z 1 Z ∞ √ Z 1 1 ∞√ 2 + s – 1 cos(sx) ds = s 2 – 1 cos(sx) ds K (x) = π 0 π 0 1 Z 1 (5.69) Z ∞p p 1 s (2 – s ) cos[(1 – s )x] ds + s (2 + s ) cos[(1 + s )x] ds . = π 0 0

Now Erdelyi’s lemma (see Section 1.4) shows that K (x) ∼ x –3/2 , x → +∞, if we set there α = 1, β = 3/2 and express cosines in Eq. (5.69) in terms of exponential functions of imaginary arguments with the help of Euler’s formula. β ) By analogy with the problem α) we can find here that √ √ √ p–i –if s + 1 –if 1 + ip –πi/4 = =e f , p = –is , (5.70) V+ (s ) = –is p p

hence (Bateman and Erdelyi, 1954) ix ix √ √ √ e e + –i Erf( –ix) = e–πi/4 f √ – 2 i[C (x) + iS (x)] . v (x) = e–πi/4 f √ πx πx (5.71) Now, having derived explicit expression for v (x), estimate of the right-hand integral is based upon the following asymptotic behavior at infinity: eix eix v (x) ∼ –if – √ 3/2 + O x–3/2 = ik – √ 3/2 + O x–3/2 πx πx (5.72) ix e –3/2 , x → +∞. = w – √ 3/2 + O x πx If we add to this estimate also the asymptotics of the kernel Z ∞ i eisx ei(x+π/4) 1 √ ds = H0(1) (|x|) ∼ √ , x → +∞, K (x) = 2π –∞ s 2 – 1 2 2πx it becomes clear that the tail in (5) is asymptotically small: Z ∞ [v (2ak + ξ ) – w]K (x + ξ ) dξ 0 Z ei(2ak+x–π/4) ∞ e2iξ √ dξ = O (ak )–3/2 , ak → ∞. ∼ –√ 3/2 π (2ak ) x+ξ 0

(5.73)

(5.74)


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Helpful remarks 1◦ . The structure of the derived solution is such that there is a degenerate solution w, which asymptotically represents the actual solution over the total interval |x| ≤ ak , except small boundary layers near the screen (or hole) ends x ∼ ±ak , where the structure of the solution is affected by these sharp edges. This idea is quite clear since if |x ± ak | = O (1), then v (ak ± x) → w, k → ∞, and hence in (5.61) we have g (x) → w, k → ∞. We thus can conclude that the degenerate solution, which is obtained from the integral equation over infinite axis |x| < ∞, is valid over the considered interval. This should be corrected only in a small neighborhood of the edges, and correct structure of the solution in these small boundary layers can be determined by solving the corresponding Wiener–Hopf equation. 2◦ . Heuristically, the degenerate solution, which is valid outside small boundary layers, must be for high frequencies in accordance with Kirchhoff’s physical diffraction theory. We can verify that these heuristic ideas are correct. Indeed, in problem α) the pressure “on the light” is, according to Kirchhoff’s theory, p(x, –0) = p 1 (x) = 2pinc = 2, |x| ≤ a, and solution “in the shadow” is p(x, +0) = p2 (x) = 0. Then g2 (x) = p1 (x) – p2 (x) = 2, which coincides with our obtained value w = 2. In problem β ) the physical diffraction theory may be interpreted so that in the case, when the width of the opening is much longer than the wavelength, the structure of the solution over the opening is as the one in the incident wave, i.e., the edges of the screen modify the latter very poorly. Therefore, in this problem gβ (x) = (∂p1 /∂x)(x, 0) = (∂p2 /∂x)(x, 0) = (∂pinc/∂x)(x, 0) = ik , which coincides with the degenerate solution w = ik . 3◦ . In the literature, a standard approach to improve simple ideas of physical (and geometrical, see below Section 6.3) diffraction theory is called the method of boundary (or edge) waves (see Keller, 1962; Ho¨ nl et al., 1961). This is based on some nonstrict physical observations, and is also reduced to a Wiener–Hopf equation. The method described here gives a direct mathematical instrument, which admits a clear mathematical justification. 4◦ . From results of Section 5.1 we could come to an observation that the behavior of the solution near the interval ends is ∼ (ak ± x)1/2 (x → ∓ak ) in problem α). It is interesting to check if this qualitative property is provided by the constructed solutions. As follows from Eq. (5.61), (5.65), if x → ak , then g (x) → v (2ak ) – w + v (ak – x) = O (ak – x)1/2 , since Erf(ax) = O (x), x → 0. By analogy, it can be easily seen from formula (5.71) that this possesses the required square root singularity near the screen edges in problem β ).

5.4. High-Frequency Asymptotics for Diffraction by Linear Obstacles in Open Waveguides Many efforts of researchers were devoted to study wave propagation in waveguides (see, for example, Chapter 3; Jones, 1986; Achenbach, 1980; Mittra and Lee, 1971). From the theoretical point of view the case of a homogeneous infinite waveguide of constant width has been studied by means of the Fourier transform and the obtained results have now become classical (Collin, 1960). Through the application of the Wiener–Hopf method (Section 1.2), many problems for open-ended waveguides have been successfully investigated (see Vajnshtejn, 1969). It was clarified that the wave field in elongated waveguides is locally similar to that in infinite ones (results contiguous to what is obtained in the previous section). The edge waves method developed in the previous section cannot be transferred directly to the diffraction by a finite-length waveguide, due to presence of the propagating mode waves that lead to a strong interaction between the ends of the waveguide. Nevertheless, for specific values of the frequency, close to these mode values, Vajnshtejn (1969) first


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proposed, on the basis of some heuristic physical ideas, an approach similar to the method of edge waves. He discovered that at these frequencies the wave propagating along the waveguide does not permit any mode distinct from the incident one, when reflecting from the open edge. Hence, it is a transformation of a one-type wave, as in diffraction by the single slit (see Section 5.3), and so successive reflections from the edges correctly describe the wave picture. Nobody knows if this idea is correct for arbitrarily high frequency, different from these critical mode values. Here we develop an asymptotic approach valid for high frequencies antipodal to those of Vajnshtejn.

y h

pinc q a

-a

x -h

Figure 5.1. High-frequency diffraction by the open waveguide Let the plane incident acoustic wave pinc (x, y ) = e–ik(x cos θ+y sin θ)

(5.75)

be scattered by a finite-length waveguide with incident angle θ , as illustrated in Fig. 5.1. The full wave field, as usual, consists of the incident and the scattered one: p(x, y ) = pinc (x, y ) + psc (x, y ) and all three functions satisfy the Helmholtz equation ∆p + k 2 p = 0,

(5.76)

everywhere outside the two finite parallel plates of length 2a, with the distance between them being equal to 2h. Assume, for instance, that the plates are acoustically soft: p(x, y ) = 0:

psc (x, y ) = –pinc (x, y ),

|x| ≤ a, y = ±h.

(5.77)

By applying the Fourier transform along the x-axis to the Helmholtz equation (5.76), we obtain the following representation for the function P sc (s , y ) (cf. Sections 3.1, 5.1) in the form satisfying radiation condition at infinity:  –γy y > h,  A(s ) e , –γy sc P (s , y ) = C (s ) e + D (s ) eγy , |y | ≤ h, (5.78)  γy B (s ) e , y < –h,

where A(s ), B (s ), C (s ), D (s ) are unknown functions which √ should be defined from the boundary conditions over the screens, and γ = γ (s ) = s 2 – k 2 . In the definition of the branching function γ (s ) we use, as usual, the branch cut such that Re(γ ) ≥ 0 for arbitrary complex-valued s . It is evident that the boundary condition (5.77) provides continuity of the potential p(x, y ) over the plates: psc (x, h + 0) = psc (x, h – 0) = –pinc (x, h); psc (x, –h + 0) = psc (x, –h – 0) = –pinc (x, –h),

|x| ≤ a.

(5.79)


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Together with the continuity of the wave field outside the plates, this involves the equality p(x, h + 0) = p(x, h – 0); p(x, –h + 0) = p(x, –h – 0), –∞ < x < ∞, which gives the two relations A(s ) e–γh = C (s ) e–γh + D (s ) eγh , (5.80) B (s ) e–γh = C (s ) eγh + D (s ) e–γh . Let us introduce the new unknown functions u(x) and v (x) as follows: ( 0, |x| > a, ∂p ∂p – = ∂y y=h+0 ∂y y=h–0 u(x), |x| ≤ a, ( 0, |x| > a, ∂p ∂p – = ∂y y=–h–0 ∂y y=–h+0 v (x), |x| ≤ a.

The differences vanish here along |x| > a, due to continuity of the wave field. In terms of Fourier images the last equations are equivalent to Z a –γh γh –γh γ A(s ) e + D (s ) e – C (s ) e =– u(ξ ) eisξ dξ , –a Z a γ B (s ) e–γh + C (s )eγh – D (s ) e–γh = v (ξ ) eisξ dξ .

(5.81)

(5.82)

–a

It follows from (5.80), (5.82) that Z e–γh a v (ξ ) eisξ dξ , C (s ) = 2γ –a

D (s ) = –

e–γh 2γ

Z

a

u(ξ ) eisξ dξ ,

(5.83)

–a

with A(s ) and B (s ) being defined from (5.80). Lastly, the boundary condition (5.79) yields the following system of two integral equations (|x| ≤ a): Z a Z ∞ is(ξ–x) Z a Z ∞ is(ξ–x)–2γh e 1 e 1 u(ξ ) dξ ds – v (ξ ) dξ ds = e–ik(h sin θ+x cos θ) , 2π –a 2 γ 2 π 2 γ Z a Z –∞ Z –aa Z–∞ ∞ is(ξ–x) ∞ is(ξ–x)–2γh e 1 e 1 v (ξ ) dξ ds – u(ξ ) dξ ds = –eik(h sin θ–x cos θ) . 2π –a 2 γ 2 π 2 γ –∞ –a –∞ (5.84) By applying summation and subtraction to (5.84), we can reduce the last system to a pair of two independent integral equations: Z a [u(ξ ) + v (ξ )]K1 (x – ξ ) dξ = –4i sin(kh sin θ ) e–ikx cos θ , |x| ≤ a, (5.85) –a

Z

a

–a

where

[u(ξ ) – v (ξ )]K2 (x – ξ ) dξ = 4 cos(kh sin θ ) e–ikx cos θ ,

1 K1,2 (x) = 2π

Z

∞

–∞

G1,2 (s ) e–isx ds ,

G1,2 (s ) =

|x| ≤ a,

1 ∓ e–2γh . γ

(5.86)

(5.87)

Note that the symbolic functions G1 (s ) and G2 (s ) have the branch points s = ±k , and there are no singularities, except these two points, which is an intrinsic property of open structures. The closed ones may possess a countable set of simple poles, with a finite number of them being distributed on the real axis (compare with the results of Chapter 3).


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When any pole coincides with the origin, we have a resonant case (see Vorovich and Babeshko, 1979). As shown below, the distribution of the zeros of the symbol G 1,2 (α) is important too. To treat correctly disposition of (possibly existing) real zeros, let us apply the principle of extremely low absorption (Section 3.2). If a small attenuation is added to the medium, then the positive zeros and the branch point s = k move upwards, and the negative ones downwards. This implies that for an ideal medium, for which these points are situated on the real axis, the integration contour in (5.87) should bend around the positive points from below and the negative ones from above. It is conventionally recognized that presence of the branch points in the symbolic function complicates analytical treatment. The algorithm developed below is insensitive to the influence of the branching effect. In order to solve equations (5.85) and (5.86), it suffices to treat the following equation written in a nondimensional form: Z b w(ξ )K (x – ξ ) dξ = f (x), |x| ≤ b, χ = kh, b = a/h, –b ( ) –4i sin(χ sin θ ) w(x) (5.88) , f (x) = e–iχx cos θ , u(x) ± v (x) = h 4 cos(χ sin θ ) Z ∞ p 1 ∓ e–2γ 1 G1,2 (s ) e–isx ds , G1,2 (s ) = , γ = s 2 – χ2 . K (x) = 2π –∞ γ

There are two independent dimensionless parameters in this problem: the frequency parameter χ and the relative length of the waveguide b = a/h. The asymptotic analysis undertaken by Jones (1952) operates with a one-mode regime (π < χ < 2π ) (see Section 3.1) when b → ∞. We study the opposite case: the high-frequency regime χ = kh → ∞ with bounded b (so ka → ∞ too), which is more complex, since the number of propagating modes grows with the frequency increasing. To attract the powerful Wiener–Hopf technique, we represent again the basic equation (5.88) as a superposition of the three functions: w(x) = w1 (b + x) + w2 (b – x) – w0 (x),

(5.89)

with the new ones satisfying the following system equivalent to the initial equation (5.88): Z ∞ Z –b w1 (b + ξ )K (x – ξ ) dξ = f (x) + [w2 (b – ξ ) – w0 (ξ )]K (x – ξ ) dξ , –b < x < ∞, –b –∞ Z b Z ∞ w2 (b – ξ )K (x – ξ ) dξ = f (x) + [w1 (b + ξ ) – w0 (ξ )]K (x – ξ ) dξ , –∞ < x < b, –∞ b Z ∞ w0 (ξ ) K (x – ξ ) dξ = f (x), –∞ < x < ∞. –∞

(5.90) Here again, from the physical point of view the integral operators on the left-hand sides of (5.90) are related to one infinite waveguide and two semi-infinite ones. As we could see in previous section, for a wide class of physical problems additional integrals on the right-hand side of the first two equations (5.90), the so-called tails, appear to be small in the asymptotic sense. For instance, if the frequency is low enough, so that there is no real zero of the symbol G1,2 (s ) (k is less than the first critical value), then for a long waveguide (b = a/h 1 is the asymptotic case considered by Jones, 1952) the differences in the square brackets vanish when ξ → ±∞. Thus, it can be proved that the integrals on


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the right tend to zero with a → ∞. Physically, this means that the edges of the narrow open-ended waveguide do not influence each other in the asymptotic sense. A similar approach is possible for the short-wave diffraction problem for a single isolated plate (as well as for a single slot in the infinite plate). The symbolic function G(α) there may be formally obtained from (5.87) at h → ∞ to be G(α) ∼ 1/γ ; thus there are no zeros of the function G(α). It involves again an asymptotic disappearance of the right-hand integrals in (5.90), with the same physical meaning: the edges of the plate influence each other weakly when a/λ → ∞ (λ = 2π/k is the wavelength). The described approach generates again the edge-waves method (see Keller, 1962), which was developed by means of a different mathematical technique. In the problem at hand the zeros ±s1m , ±s2m of the function G1,2 (s ): s1m

p = χ2 – (πm)2 ,

s2m =

r

χ2

π 2 , – πm – 2

m = 1, 2, . . . ,

(5.91)

play the most important role. The number of positive zeros (5.91) increases with the growth of the frequency parameter χ = kh. It generates more and more propagating mode waves inside the waveguide. Thus, the ends of the open finite-length waveguide just affect each other by these propagating waves, and the internal wave process differs considerably from that in semi-infinite structures. As a result, the differences [w2 (b – ξ ) – w0 (ξ )], [w1 (b + ξ ) – w0 (ξ )] do not vanish as χ → ∞, in contrast to case of diffraction in unbounded medium, Section 5.3. In spite of the finite-length structure cannot be represented as a composition of two semi-infinite ones, there is a classical idea that diffraction by the open end of the finitelength waveguide is similar to diffraction by the edge of a semi-infinite one, when the wave number approaches any mode value. We mean the optical range of electromagnetism, where both geometric sizes of the structure (a and h) are several thousand times larger than the wavelength (see Vajnshtejn, 1969). Below we determine in which sense this point of view is correct. To begin with, let us write out the explicit representation for the kernels K1,2 (x) =

h p io i n (1) H0 (χ|x – ξ |) ∓ H0(1) χ 4 + (x – ξ )2 , 2

(5.92)

where the asymptotic behavior of the Hankel function H 0(1) is given by Eq. (3.115). It is clear from (5.92) that at high frequencies (χ 1) the kernel K (x) on the right-hand sides of (5.90) is strongly oscillating. So the right-hand integrals in (5.90) may be small if the differences in square brackets are asymptotically bounded. This heuristic idea should be verified. Let us assume, as usual, that these integrals are asymptotically small indeed (correctness of this assumption should be verified afterwards). Then we can construct the solution of the equations (5.90) over the semi-infinite intervals by using the Wiener–Hopf method (Section 1.2). It gives the following result: ∞ X 1 (πm)2 e–iχ(x–b) cos θ – G+ (sm ) ei(sm x+χb cos θ) , w1 (x) = G(χ cos θ ) 2G+ (–χ cos θ ) m=1 sm (sm + χ cos θ )

w2 (x) =

∞ X 1 (πm)2 eiχ(x–b) cos θ – G+ (sm ) ei(sm x–χb cos θ) , G(χ cos θ ) 2G+ (χ cos θ ) m=1 sm (sm – χ cos θ )

where G(s ) = G+ (s ) G– (s ) is a factorization of the function G(α).

(5.93)


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A solution of the third equation (5.90) is obtained by the application of the Fourier transform as follows: e–iχx cos θ , (5.94) w0 (x) = G(χ cos θ ) hence we can give the final representation for w(x), under the accepted assumption, as w(x) =

×

∞ 1X G+ (sm ) e–iχx cos θ – (πm)2 G(χ cos θ ) 2 m=1 sm

ei[sm (b+x)+χb cos θ] ei[sm (b–x)–χb cos θ] + (sm – χ cos θ ) G+ (χ cos θ ) (sm + χ cos θ ) G+ (–χ cos θ )

+O χ

–1/2

(5.95) .

The asymptotic estimate of the error of this formula is proved below. The structure of the obtained solution (5.95) is the following. The first term corresponds to the case of an infinite waveguide and has the order O (χ), χ → ∞ (so far as G(χ cos θ ) = i(1 ∓ exp(2χi sin θ ))/(χ sin θ )). Several oscillating terms under the sum sign may appear to have the same order O (χ) with respect to the parameter χ (if α m ∼ χ cos θ ). Thus, the behavior of the solution is also highly oscillating. When estimating the value of the right-hand side integrals in (5.90), let us note that the argument (x – ξ )√in the kernel K (x – ξ ) is of a constant sign for all x and ξ there, so K (x – ξ ) ∼ O (1/ χ ) uniformly on x and ξ . Thus, the integrand is of the order of O (χ1/2 ), χ → ∞. Besides, there is no stationary point, in contrast to the left-hand side kernel with its stationary point x – ξ = 0 ∼ ξ = x. Therefore, it can be shown by integration by parts that, owing to the oscillating structure of the integrand, the value of the integrals has the order O (χ1/2 )/χ = O (χ–1/2 ), χ → ∞. This proves the correctness of the method when the right-hand integrals in (5.90) are neglected. The only breakdown takes place when the value of the variable x coincides with the edge point (x = –b in (5.90) 1 and x = b in (5.90)2 ). For these values the argument (x – ξ ) may be small, and the asymptotic estimate (5.93) for the kernel is not valid. Thus, the developed approach gives the correct representation (5.95) all over the interval –b < x < b, excluding the boundary layers where b ± x ∼ 1/χ. Another restriction is related to the cases when the incident angle θ approaches those values where the incident wave generates standing modes. In these cases G(χ cos θ ) tends to zero, w 1 (x) and w2 (x) grow to infinity, and the right-hand integrals in (5.90) become unbounded. Thus, for some particular values of the frequency and the angle of incidence the main asymptotic result (5.95) fails.

Helpful remarks 1◦ . The main obtained formula (5.95) is worthy of special discussion. An interesting fact is that the mode waves do not decay with distance; however some integrals of these wave functions decay, so they could be neglected in (5.90). The nondecaying structure of the wave field in (5.95) shows that the solution has no relation to the problem of semi-infinite waveguide where only one type of edge-wave, instead of two, is present. It should also be noted that the asymptotic contribution to the sum in (5.95) is given only by nondecaying terms in the sum (when sm is real). Therefore, in practice the summation can be restricted by m = n if the frequency parameter is χ = π (n + δ ) (0 < δ < 1/2 or 1/2 < δ < 1), where n is a large positive integer. Thus, formula (5.95) is quite appropriate to evaluate the split function w(x) at high frequencies, because efficient high-frequency representations for the coefficients G+ (sm ) are proposed in Mittra and Lee (1971).


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2◦ . Whenever a solution of the main integral equation is determined by (5.95), the scattered wave field psc (x, y ) can be obtained directly from (5.78) and (5.83) Z p i a sc v (ξ )H0(1) k (x – ξ )2 + (y + h)2 dξ p (x, y ) = 4 –a Z (5.96) p i a u(ξ )H0(1) k (x – ξ )2 + (y – h)2 dξ , – 4 –a

uniformly over (x, y ).

5.5. High-Frequency Diffraction by a Linear Discontinuity in the Waveguide Here we consider a problem contiguous to that studied in the previous section. We will see that qualitative properties of the main symbolic function (i.e., the Fourier transform of the kernel of the basic integral equation) for this closed waveguide are absolutely different from the one for open structure studied in Section 5.4.

y

p=0

b

h -l p=0

p = p0 l

a x

Figure 5.2. Oscillating linear plate in a closed infinite waveguide To be more specific, we consider acoustic waveguide (i.e., a layer of constant thickness) with acoustically soft faces (see Fig. 5.2). Let a plane rigid finite-length plate be located inside a layer. Let its faces y = ±0, |x| < l oscillate with an amplitude p 0 , which is assumed, for definiteness, to be uniform, p0 = const, without any restriction. As usual, we study the wave process harmonic in time, where the wave pressure p(x, y ) satisfies the Helmholtz equation. By applying the Fourier transform along the x axis to the wave (Helmholtz) equation, we arrive at the following representation for the function P (α, y ), which is chosen so that the boundary conditions p(x, b) = p(x, –a) = 0 on the free boundary surfaces are automatically satisfied: A(s ) sinh[γ (y + a)], –a < y < 0, P (s , y ) = (5.97) B (s ) sinh[γ (y – b)], 0 < y < b, where A√ (s ) and B (s ) are unknown functions to be defined from the boundary conditions, and γ = s 2 – k 2 . Obviously, the wave field is continuous at |x| > l along the line y = 0. On the other hand, the boundary condition p = p0 , y = ±0, |x| ≤ l implies its continuity on the interval |x| ≤ l also. Thus we have p(x, –0) = p(x, +0), |x| < ∞, and hence sinh(aγ ) . (5.98) A(s ) sinh(aγ ) = –B (s ) sinh(bγ ) ∼ B (s ) = –A(s ) sinh(bγ ) Let us introduce a new unknown function u(x), |x| ≤ l, so that 0, |x| > l, ∂p ∂p – = u(x), |x| ≤ l. ∂y ∂y y=–0

(5.99)

y=+0


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Then the function A(s ) can be expressed in terms of u(x): Z sinh(bγ ) l u(ξ ) eisξ dξ , h = a + b. A(s ) = sinh(hγ ) –l

(5.100)

At last, the boundary condition p = p0 , |x| ≤ l, with the help of Eqs. (5.97)–(5.100), yields the following integral equation for the function u(x): Z l u(ξ )K (x – ξ ) dξ = p0 , (|x| ≤ l), –l (5.101) Z ∞ sinh(aγ ) sinh(bγ ) 1 –isx G(s )e ds , G(s ) = . K (x) = 2π –∞ γ sinh(hγ ) The symbolic function G(s ), given by (5.101), is well known in the theory of branching electromagnetic waveguides. However there is no mathematical theory appropriate for Eq. (5.101) at high frequencies. For the sake of brevity, further consideration is restricted, without any loss of generality, to a particular case a = b, i.e., for the discontinuity positioned parallel to and midway between the faces of the waveguide. Let us rewrite equation (5.101) in a nondimensional form (h = 2a = 2b) Z d u(x) , v (ξ )K (x – ξ ) dξ = 1, |x| ≤ d, χ = ka, d = l/a, v (x) = 2p0 –d (5.102) Z p tanh γ 1 –iαx 2 2 G(α)e dα, G(α) = , γ = α –χ . K (x) = 2π σ γ

The short-wave range implies that the parameter d is fixed, and the frequency parameter χ 1. The function G(s ) in Eq. (5.102) is a meromorphic function of the complex variable s (see Section 1.1). There is a countable set of its zeros ±s m : p m = 1, 2, . . . (5.103) αm = χ2 – (πm)2 , and a countable set of its poles ±βm : p βm = χ2 – [π (m – 1/2)]2 ,

m = 1, 2, . . . ,

(5.104)

so the kernel (5.102) can be explicitly represented as a series in residues at simple poles (Section 1.1): ∞ X exp(iβm x) . (5.105) K (x) = i βm m=1

Such a representation permits reducing the problem to an infinite system of linear algebraic equations (see Mittra and Lee, 1971). For arbitrary large value of the frequency parameter χ there is a finite number of real zeros and poles. When a pole approaches the origin with the frequency change, then we encounter a resonance. If a small attenuation is added to the medium, then the positive poles move upwards, and the negative ones move downwards. It implies that for the ideal medium, when these singular points are situated on the real axis, the integration contour σ in (5.102) should bend round the positive poles from below and the negative ones from above.


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It should be noted that with the frequency increasing (the parameter χ tends to infinity) more and more zeros and poles appear on the real axis. This essential feature complicates the high-frequency analysis. Let χ = π (n + δ ),

0 ≤ δ < 1,

δ ≠ 1/2,

(5.106)

with n being a large positive integer: n 1. The critical value δ = 1/2, related to the resonance case, should be excluded from the consideration. As in the previous sections, it is easily seen that the basic equation (5.102) is equivalent to a pair of equations for the new unknown functions p(x) and w(x), which hold, respectively, on semi-infinite and infinite interval: Z ∞ Z ∞ w1 (ξ )K (x – ξ ) dξ = 1 + [w1 (2d + ξ ) – w]K (x + ξ ) dξ , 0 < x < ∞, 0 Z0 ∞ (5.107) w(ξ )K (x – ξ ) dξ = 1, –∞ < x < ∞, –∞

if

v (x) = w1 (d + x) + w1 (d – x) – w,

|x| ≤ d.

(5.108)

The main result of the present work is given by the following THEOREM. If δ ≠ 0, 1/2, 1 and x > 0, then the right-hand integral in Eq. (5.107) is asymptotically small: Z ∞ [w1 (2d + ξ ) – w]K (x + ξ ) dξ → 0, n → ∞, (5.109) 0

and so can be rejected.

The proof is based again on the asymptotic properties of the kernel K (x). Representation (5.105) is not suitable for this purpose, and we use here a different expansion. The function G(s ) in Eq. (5.102) can be rewritten as ∞ X 1 2 exp(–2γ ) 1 exp(–2mγ ) 1 – exp(–2γ ) = – = –2 (–1)m–1 . (5.110) G(x) = γ [1 + exp(–2γ )] γ γ [1 + exp(–2γ )] γ γ m=1

Taking into account the integral identity (see Eq. (1.17)) Z √ exp(–|z |γ ) –isx e ds = πiH0(1) χ z 2 + x2 , γ σ

(5.111)

we then arrive at the following expression for the kernel K (x):

∞ X √ i (1) (–1)m–1 H0(1) χ 4m2 + x2 , K (x) = H0 (χ|x|) – i 2 m=1

(5.112)

where H0(1) is the Hankel function. It is evident that with χ → ∞ the argument of each term under the sum in Eq. (5.112) increases infinitely, uniformly with respect to x. Hence, with the help of the asymptotic formula (cf. Eq. (3.115)) r 2 i(z–π/4) e , z → ∞, (5.113) H0(1) (z ) ∼ πz


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the uniform high-frequency representation for the kernel can be finally obtained as follows: s √ ∞ 2 2 2 πi/4 X i (1) m–1 exp(iχ 4m + x ) (–1) . (5.114) e K (x) ∼ H0 (χ|x|) – 2 2 1/4 2 πχ (4m + x ) m=1

For the open finite-length waveguide (cf. Section 5.4) there is only the first term of the infinite series (5.114). It should also be noted that the term (i/2)H 0(1) (χ|x|) cannot be simplified asymptotically at χ → ∞, since the argument of the kernel K (x + ξ ) in the considered right-hand integral (5.107) may be arbitrarily small when ξ ∼ x ∼ 0. To prove the theorem let us suppose, as usual, that its statement is valid. Then the function w1 (x) is a solution of the Wiener–Hopf integral equation (5.107) holding over a semi-infinite interval. Omitting some routine transformations, one comes to the following expression for its solution: ∞ X (πm)2 G+ (αm ) iαm x 1 1 – e w1 (x) = G(0) 2G+ (0) m=1 α2m

where

∞ 1 X G+ (αm ) (πm)2 iαm x χ – e , = tan δ 2 m=1 G+ (0) α2m

G(s ) = G+ (s ) G– (s ),

G– (–s ) = G+ (s ),

(5.115)

x > 0,

(5.116)

is a result of the factorization of the symbolic function G(s ) given by Eq. (5.102). The solution of the convolution equation (5.107) can be trivially obtained by using the Fourier transform, as follows: w(x) ≡ w =

χ 1 = . G(0) tan δ

(5.117)

It now becomes clear from Eqs. (5.115), (5.117) that the values δ = 0 and δ = 1 (as well as the resonance value δ = 1/2) should be excluded from the consideration, which is stated in the body of the theorem. If δ does not coincide with any of those critical values, then w1 (x) and w(x) are uniformly of the order of O (χ), χ → ∞. It is interesting to note that a few first terms in the series (5.115), in addition to the first one, have the same order O (χ), χ → ∞ (for m such that πm ∼ χ). Note also that the summation in the series (5.115) may be taken over 1 ≤ m ≤ n, because, in the asymptotic sense, only for these values of m the quantities αm are real. All other terms are exponentially small (iα m ∼ –χ1/2 , m > n), since the argument of the function w1 (2d + ξ ), ξ > 0 in Eq. (5.107)1 is greater than 2d (i.e., x > 2d in (5.115)). Let us estimate the proper asymptotic order of the right-hand integral in Eq. (5.107): Z Z ∞ i ∞ [w1 (2d + ξ ) – w]H0(1) [χ(x + ξ )] dξ [w1 (2d + ξ ) – w]K (x + ξ ) dξ = 2 0 0 s √ (5.118) Z ∞ ∞ exp[iχ 4m2 + (x + ξ )2 ] 2 πi/4 X m–1 e (–1) [w1 (2d + ξ ) – w] dξ . – πχ [4m2 + (x + ξ )2 ]1/4 0 m=1

The right-hand integrals in Eq. (5.118) should be estimated by different ways for the free term and the terms in the series. For the last ones we introduce the phase function p (5.119) S (x) = 4m2 + (x + ξ )2 , © 2005 by CRC Press LLC

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which determines highly oscillating nature of the exponential function exp[iχS (x)] in (5.118). It is evident that S 0 (x) = (x + ξ )/S (x) > 0, when x > 0, ξ ≥ 0, hence this function has no stationary point. Thus a standard multiple integration by parts proves that the order of each term under the sum sign in Eq. (5.118) is less than O (χ –β ), with β being an arbitrarily large positive number: β 1. Further, the free right-hand side term in Eq. (5.118) can be estimated as Z ∞ Z ∞ w1 (2d + ξ/χ) – w (1) H0 (χx + ξ ) dξ [w1 (2d + ξ ) – w] H0(1) [χ(x + ξ )] dξ = χ 0 0 Z ∞ eiχx 1 w1 (2d + ξ/χ) – w iξ e dξ √ = O √ , ∼ χ χx χx 0 (5.120) if the value x is outside of a boundary layer (where x ∼ 1/χ). Thus the theorem is proved. The problem in hand seems to be a typical asymptotic problem with singular perturbations. The structure of the solution outside of the boundary layer is absolutely different from that valid near the boundary point x = 0. As can be seen, the theorem indicates how to construct an external solution only. Indeed, by substituting (5.115) and (5.117) into Eq. (5.108), one comes to the following high-frequency representation: v (x) =

n 1 X G+ (αm ) (πm)2 iαm (d+x) χ iαm (d–x) – e + e , tan δ 2 m=1 G+ (0) α2

|x| < d.

(5.121)

The solution in the boundary layer d ± x ∼ 1/χ cannot be obtained by this method. Note that a finite number of terms 1 ≤ m ≤ n are taken in the series (5.121), since all other terms are exponentially small outside of the boundary layer (α m becomes imaginary). Efficient high-frequency representation for G+ (α). The main asymptotic formula (5.121) contains some values of the function G + (s ) (s is real) which is a result of the factorization of the main symbolic function G(s ) (5.102). It can be written as the product of three functions: G = G 1 G–1 2 G3 , where G1 = 1 – exp(–2γ ), √ 2 –1 2 G2 = 1 + exp(–2γ ), G3 = (γ ) , and γ = s – χ , with the trivial factorization of the third of them: G3 (s ) = (s + χ)–1/2 (s – χ)–1/2 . For the first two functions one can deal with the general representation (see Section 1.2; Noble, 1962) √ ( ) Z –2 u2 –χ2 ln 1 ∓ e 1 + du , (5.122) G1,2 (s ) = exp 2πi σ u–s which has too complex form. A good high-frequency approximation proposed by Vajnshtejn (1969) is known: r 1–i s δ + + exp –0.824 √ s , χ = π (n + δ ), 0 < δ < 1. (5.123) 1+ √ G1 G3 (s ) ≈ 2 n 2χ 2πδχ The last approximation is correct when |α| χ, and 0 < δ 1, i.e., for high frequencies near eigen-modes. Hence, that is of no interest for the present investigation since the asymptotic approach proposed fails for δ → 0, as discussed above. We start from the following approximate equality: r Z ∞ s∗ χ ln {1 – exp [2i(πδ + t)]} dt eπi/4 + √ exp (s∗ = s/χ), (5.124) G1 (s ) ≈ √ s+χ π 2 0 2t + χs∗2 t


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which is known to be valid at χ 1 and arbitrary Im(s ) ≥ 0 (see Mittra and Lee, 1971). By the application of the Taylor expansion to the logarithmic function, the last integral can be rewritten as Z ∞ X e2πδji ∞ e2tji dt √ – j 2t + χs∗2 t 0 j=1 (5.125) ∞ o –iπ/4 π X e2πδji n√ 2 2 –iχs∗2 j 2 S (χs∗ j )i + C (χs∗ j ) e –1 e , = √ s∗ 2χ j=1 j

where S (x) and C (s ) are the Fresnel integrals. These possess some rational approximation (see Abramowitz and Stegun, 1965) so that the integral in (5.124) can be reduced to the form " ! !# r r ∞ 2πδji X e 2 2 π if α∗ χj + g α∗ χj , (5.126) = – √ e–πi/4 s∗ χ j π π j=1 where f (x) ≈

1 + 1.22x , 2 + 2.17x + 1.22πx2

g (x) ≈

1 , 2 + 4.14x + 3.49x2 + 6.67x3

(5.127)

uniformly over 0 < x < ∞. P∞ If we extract the most slowly decaying term in the sum (5.126), 1 [exp(2πδji)/j 3/2 ], then the series becomes rapidly convergent. The extracted series can be calculated by using our efficient approximation for the functions Cβ (x), Sβ (x) (see Section 3.4). Finally, we arrive at the following high-frequency representation: * ( p p πi/4 iC ( s 2 χ/π ) – S ( s 2χ/π ) i – 1 e 3/2 ∗ 3/2 ∗ √ exp G+1 (s ) ≈ √ s+χ 2 s∗ 2πχ #)+ " ! ! (5.128) r r ∞ X e2πδji i 2 2 , if s∗ χj + g s∗ χj – √ + j π π s∗ 2πχj j=1 where the functions f , g are determined by (5.127) and C 3/2 , S3/2 in Section 3.4. The corresponding high-frequency approximation for G +2 (s ) can be derived likewise. The common term of the series (5.128) decreases as O (j –5/2 ), so this formula permits efficient calculations indeed. The convergence of the series √ slower √ in (5.128) becomes when δ → 0, because there is a number of small s∗m = (πm)2 – χ2 /χ ∼ δ , so the acceleration procedure in this range becomes inefficient. The discussion in the last two sections follows the authors’ works (Scalia and Sumbatyan, 1999, 2001).

Helpful remarks As we have already discussed, the problems considered in the two last sections lead to integral equations containing two independent dimensionless parameters. The first of them is purely geometric, and this characterizes relative length of the slits, that is, b = a/h (Section 5.4) or d = l/a (this section). The second parameter χ is related to the frequency, being large for high frequencies. This type of high-frequency problem gives a bright example when heuristic ideas may lead to mistaken conclusions. Indeed, in the considered


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case of high-frequency diffraction the wavelength becomes asymptotically small when compared with the slits length or the width of the structure. So, there are two very widespread (mistaken) heuristic opinions. The first of them is based on the fact that if the width (thickness) of the waveguide is much greater than the wavelength, then the leading asymptotic term of the solution on the slit coincides with that in the problem for a half-plane. Another (mistaken) point of view declares that the solution for the finitelength slit is asymptotically the same as for semi-infinite one, since the slit is much longer than the wavelength. Both viewpoints ignore the fact that with frequency increasing both dimensional sizes become large with respect to the wavelength simultaneously. Formulas derived in the present chapter represent mathematical structure, which provides actual behavior with multiple re-reflections of waves between the faces of the waveguide.

5.6. Waves in Elastic Half-Space. Factorization of the Rayleigh Function Let us consider vertical harmonic oscillations of a rigid punch coupled without friction with an elastic half-plane (2D problem). For this problem, as in many problems studied above, the choice of methods of investigation depends on the value of frequency. As usual, the case of low and moderate frequencies can be investigated by using regular (analytical or numerical) methods. With the frequency increasing regular methods lead to a loss of the computational stability. That is why we undertake here again an asymptotic analysis appropriate just for the high frequency case. In order to derive the basic integral equation of the considered contact problem of dynamic elasticity, let us come back to equalities (5.41), which establish relations between some components of the stress tensor and those of the displacement vector. In the case when there is known a normal component of the stress, we arrived at the integral equation (5.45) of the dynamic cracks theory. If, in the opposite case, we know the amplitude of normal displacement, then it can be easily seen from Eq. (5.41) that we arrive at the following integral equation of the first kind: Z ∞ Z 1 1 µ L∗ (s )e–isx ds , p(ξ )K (χ|x – t|) dξ = w, |x| ≤ 1, K (x) = a 2π –∞ –1 √ 2 u – β2 √ , χ = aks , β = kp /ks = cs /cp < 1. L∗ (u) = √ 4u2 u2 – β 2 u2 – 1 – (2u2 – 1)2 (5.129) In Eq. (5.129), p(x) is the amplitude of the normal contact pressure (the contact area is free of friction), w is the amplitude of the punch oscillations, χ is a parameter, which is large at high frequencies, µ is the shear elastic modulus, and a is the semi-width of the punch. As in the previous sections, we will apply here our idea based upon the fact that initial equation (5.129) is (exactly) equivalent to the system: Z ∞ Z ∞ ϕ(ξ )K (x – ξ ) dξ = χ + [ϕ (2χ + τ ) – v (τ )] K (x + τ ) dτ , x ≥ 0, 0 0 Z ∞ (5.130) v (t)K [χ|x – ξ |) dξ = 1, |x| < ∞, –∞

if p(x) =

µ w {ϕ [χ(1 + x)] + ϕ [χ(1 – x)] – v (x)} , a

|x| ≤ 1.

(5.131)


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As we could see above, in many problems the last integral in (5.130) may be rejected, and equation (5.129) becomes asymptotically as a Wiener–Hopf equation on the semiinfinite interval. Usually, this brings the error of the order of exp(–χε) (ε > 0), uniformly with respect to x. Sometimes this error has a power character, i.e., becomes more essential. Below we will show that in the present problem the tail on the right-hand side of Eq. (5.129) is an (asymptotically) small quantity just of some power degree. Namely, the estimate Z ∞ χ → ∞, (5.132) [ϕ (2χ + τ ) – v (τ )] K (x + τ ) dτ = O (χ–1/2 ), 0

holds uniformly over x ≥ 0. As in the previous sections we aim at construction of the leading asymptotic term. Heuristically, the estimate (5.132) means again that the leading asymptotic term (5.131) is determined from the solution for a pair of semi-infinite punches (5.130) and one infinite punch. Physically, this means that the high-frequency oscillations generate the waves of so small length that perturbations from the right edge of the punch do practically not influence the wave process near the left edge, and vice versa. A solution of the convolution equation (5.130) is constructed immediately by applying the Fourier transform, and is given as follows: v (x) = –

i ≡ v. βχ

In order to successfully solve the equation on the semi-axis Z ∞ ϕ(ξ )K (x – ξ ) dξ = χ,

(5.133)

(5.134)

0

it is necessary to arrange factorization of the symbolic function (Fourier image of the kernel). As known (see Babeshko, 1971; Vorovich and Babeshko, 1979), the solution of equation (5.134) is stable with respect to small perturbations of the symbol on the real axis, so let us apply an approximate factorization (cf. Section 1.2). √ √ s + β , s + 1, The symbolic function L ( u ) is a combination of the four root squares ∗ √ √ s – β , s – 1, with each of them having branching points. Let us draw in the complex plane s the cuts (cf. Section 1.1) that join the points s = –β and s = –1 with infinity in the lower half-plane, and the points s = β and s = 1 with infinity in the upper half-plane. In addition to the branching points, on the real axis Im(s ) = 0 there are two Rayleigh poles of the symbolic function: a = ±s1 , s1 > 1. According to the principle of extremely small absorption (see Section 3.2), the integration contour in the integral representation of the kernel in Eq. (5.129) coincides with the real axis, bending around positive singularities from below, and negative ones from above. The function L∗ (s ) has qualitatively different behavior on different intervals of the real axis. For |s | ≥ 1 it is real-valued, for β < |s | < 1 it is complex-valued, and for |s | ≤ β it is imaginary-valued. Let us approximate the symbol L∗ (s ) by the expression √ √ p 4s 2 s 2 – β 2 s 2 – 1 + (2s 2 – 1)2 2 2 L∗ (s ) = s – β 16s 4 (s 2 – β 2 )(s 2 – 1) – (2s 2 – 1)4 √ A s2 – β2 (5.135) M+ (s )M– (s ), ≈ L(s ) = 2 2 2 2 2 (s – s1 )(s – z )(s – z¯ 2 ) p √ √ 2 2s ±1 (A, B , Im z > 0). M± (s ) = Bs s ± β s ± 1 + © 2005 by CRC Press LLC

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This function L(s ), like L∗ (s ), is even, has two Rayleigh poles s = ±s1 and possesses the same qualitative behavior on various intervals of the real axis. Moreover, it catches the behavior at the origin and at the infinity exactly. Besides, on all these intervals it has a true sign of the imaginary part, which is important for the uniqueness theorem to be valid. The expression M+ (s ) has a zero in the upper half-plane, which should be liquidated by the zero of the denominator s = –z¯ . Let us note that the point s = –z is another zero. With the help of this approximation factorization of the function L(s ) can be made very easily, in the following way: √ A s+β M+ (s ). (5.136) L+ (s ) = (s + s1 )(s + z )(s + z¯ ) The realization of the Wiener–Hopf method for equation (5.134), with the help of factorization (5.136), leads to the following expression for the Fourier image of the function ϕ(x): (s + s1 )(s + z )(s + z¯ ) √ s s + β M+ (s ) i √ √ 2 (s + s1 )(s + z )(s + z¯ ) h p √ Bs s + β s + 1 – ( 2 s + 1) , = C s s + β ∆(s ) √ 4 ∆(u) = B 2 u2 (u + β )(u + 1) – ( 2 u + 1) = d(s + z )(s + z¯ )(s + η )(s + η¯ ), Φ+ (s ) = C

2

d = B – 4,

(5.137)

s1 |z |2 C = √ χ. A β

Therefore,

# " √ √ a4 a5 a3 s+1 s+1 C √ √ + + , a1 + a2 + √ Φ+ (s ) = d s+η s + η¯ s s + β (s + η ) s + β (s + η¯ ) u + β √ 2 u1 (u1 – η )( 2 η – 1) u1 – η , a2 = a¯ 1 , a3 = – 2 , a4 = , a5 = a¯ 4 . a1 = B η¯ – η η (¯η – η ) |η | (5.138) This permits explicit representation for the function ϕ(x), since the following inversion formulas hold (see Bateman and Erdelyi, 1954): p eiηx 1 √ ⇐= –i √ Erf –ix(β – η ), x > 0, (s + η ) s + β β–η √ (5.139) p p e–i(π/4–x) s+1 iηx ⇐= √ – ie 1 – η Erf –ix(1 – η ), x > 0. s+η πx

Note that this solution can be expressed in terms of the Fresnel integrals (compare with Section 5.3). From explicit form of the function ϕ(x), as well as directly from Eq. (5.138), we can obtain the estimate ϕ(x) ∼ v + a eiβx x–1/2 , x → +∞, (5.140) which indicates that the boundary layer solution, as we did expect, turns in the outer zone into a solution for an infinite punch (5.133). However, we can see that this process is very slow. If we add to the estimate (5.140) also the evident estimate K (x) ∼ c1 eiu1 x + c2 eiβx x–3/2 + c3 eix x–3/2 ,

then we justify the asymptotic relation (5.132).

x → +∞,

(5.141)


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It is interesting for many applications to define the relation between the applied force and the settlement of the punch base (we assume the punch to be weightless) Z

Z

1

1

p(x) dx = µ w {ϕ [χ(1 + x)] + ϕ [χ(1 – x)] – v } dx –1 –1 Z ∞ = µ w 2v + 2 {ϕ [χ(1 + x)] – v } dx .

P =b

(5.142)

–1

The last equality holds, due to the estimate (5.140), with the error O (χ –1/2 ), which is equal to the error of the constructed solution of the equation (5.129). Further, since Z Z ∞ 1 1 ∞ [ϕ(ξ ) – v H (ξ )] dt = [Φ+ (s ) – V+ (s )]s=0 (5.143) {ϕ [χ(1 + x)] – v } dx = χ –∞ χ –1 (H (ξ ) is the Heaviside function), then finally we obtain the following expression for the pliability of the foundation: a3 β 1 1 a4 P – 3/2 . , D= 2 Re a1 + √ (5.144) w= 2µ CDβ – iχ d η 2β β Let us note that the constants C and D in Eq. (5.144) are real-valued, so, for instance, the shift in phases between the settlement and the applied force is given by the formula χ . (5.145) θ = – arctan CDβ

Helpful remarks For a long time the problem of factorization of the classical Rayleigh function remained unsolved, and in many problems where there was a need of this factorization some authors applied exact factorization formulas. This leads to extremely cumbersome calculations with very complex integrals. Instead, we use here an approximate factorization (5.135), which keeps all qualitative properties of the (complex-valued) initial Rayleigh function (compare with what is discussed in Section 1.2). Quantitatively, an arbitrarily precise approximation can be achieved by a regular factor representing a combination of some polynomials (as discussed in Section 1.2). Our calculations show that the proposed approximation, as follows from Eq. (5.135), provides an accuracy of few percent.

5.7. Integral Equation of the Mixed Boundary Value Problem for Elastic Layer In this section we will demonstrate that the asymptotic approach developed in Section 5.5 for scalar problems in acoustic layer with mixed boundary conditions can be expanded to the case of elastic layer (2D in-plane problem). Let a rigid punch of the width 2b be undertaken by vertical harmonic oscillations being placed on the free surface of elastic strip. Let the latter, to be more specific, be placed on an absolutely rigid basis. We assume that the friction in the contact zone, as well as between the basis and the strip, is absent. As follows from results of Section 3.3, the considered


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problem can be reduced to the integral equation for the unknown normal contact stress p(x), related to the quantity µw/h: Z ∞ Z a 1 L(s ) e–iχxu ds , p(ξ )K (x – ξ ) dξ = 1, |x| < a, (a = b/h), K (x) = 2π –∞ –a L(s ) = L1 (s ) – L2 (s ), L1 (s ) = σ1 /∆(s ), L2 (s ) = σ1 P1 (s )/∆(s ), √ p P1 (s ) = e–2χσ1 + e–2χσ2 – e–2χ(σ1 +σ2 ) , χ = hks , σ1 = s 2 – β 2 , σ2 = s 2 – 1, ∆(s ) = 4s 2 σ1 σ2 G1 (s )F2 (s ) – (2s 2 – 1)2 G2 (s )F1 (s ),

Gj (s ) = 1 – e

–2χσj

,

Fj (s ) = 1 + e

–2χσj

,

β = kp /ks = cs /cp < 1,

j = 1, 2,

(5.146) where h is the thickness of the layer. Equation (5.146) is written again in a dimensionless form, which is more convenient to be represented in a little different manner compared with the previous section. Note that the function ∆(s ) represents again the Rayleigh–Lamb function, written in a slightly different manner. Similarly to the scalar problem (Section 5.5) with increasing frequency (i.e., with the growth of dimensionless parameter χ) the number of zeros and poles of the symbolic function L(s ), situated on the real axis, unboundedly grows. Location of real zeros of the symbol ±αk and ±βk : s s πk 2 πk 2 (k = 1, . . . , nα ), βk = β 2 – (k = 1, . . . , nβ ). (5.147) αk = 1 – χ χ In principle, the method of Sections 5.4 and 5.5 can be directly applied to the integral equation (5.146). However, justification of the using approach will be more complex and difficult. So we apply a slightly different idea, which leads to the leading asymptotic term more directly and more rapidly. Let us extract from the kernel K (x) the function K 1 (x) with the symbol L1 (s ), which does not have any zeros on the real axis and whose behavior is algebraic at u → ±∞: Z a Z a p(ξ )K1 (x – ξ ) dξ = 1 + p(ξ )K2 (x – ξ ) dξ . (5.148) –a

–a

So far as symbolic the function L1 (s ) of the kernel K1 (x) has no real pole, our standard representation of initial equation holding over the finite-length interval as a combination of three equations allows us to construct the outer (i.e., outside of the asymptotically small boundary layers near the punch edges) solution at χ → ∞ by extending the integral operator with kernel K1 to the full real axis. We thus arrive at a convolution equation whose solution is easily constructed by the application of the Fourier transform. This technique reduces the equation to the simpler form: Z ∞ Z a 1 G(s )e–iχsx ds , p(ξ )Q(x – ξ ) dξ = l0 , |x| < a, Q(x) = 2 π –∞ –a G(s ) = G1 (s )G2 (s ) = 1 – e–2χσ1 1 – e–2χσ2 , (5.149) i ∆ = – (1 + e2iχβ )(1 – e2iχ ). l0 = σ1 s=0 β Taking into account the relation 1 2π

Z

∞

–∞

e–iχsx ds =

1 δ (x), χ

(5.150)


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Eq. (5.149) can be reduced to the following form: Z a p(x) – χ p(ξ )Q1 (x – ξ ) dξ = l0 χ, |x| < a, –a Z ∞ 1 P1 (s )e–iχsx ds , P1 (s ) = 1 – G(s ). Q1 (x) = 2π –∞

(5.151)

Obviously, the kernel Q1 (x) is continuous (moreover, it is even infinitely differentiable). Now, by analogy with integral equations of the first kind, here we will apply our standard idea that equation (5.151) is equivalent to the pair of equations for two new unknown functions ϕ(x) and v (x): Z ∞ Z ∞ ϕ(x)– χ ϕ(ξ )Q1 (x – ξ ) dξ = l0 χ – χ [ϕ(2a + ξ )– v ]Q1 (x + ξ ) dξ , x > 0, (5.152) 0

0

v (x) – χ

if

Z

∞ 0

v (ξ )Q1 (x – ξ ) dξ = l0 χ,

p(x) = ϕ(a + x) + ϕ(a – x) – v ,

|x| < ∞, |x| < a.

(5.153) (5.154)

Let us prove, as usual, that integral on the right-hand side of Eq. (5.152) is asymptotically small, starting again from the properties of the kernel Q 1 (x). Similarly to the previous sections, we first obtain an asymptotic expression for Q 1 (x) at χ → ∞. The principal trouble is connected with the estimate of the integral Z ∞ 1 e–iχsx e–2χ(σ1 +σ2 ) ds . (5.155) J= 2π –∞

It can be simply shown that the main contribution to J is given by a neighborhood of the stationary point s∗ of the phase F : p √ (5.156) F (s , x) = sx + 2 1 – s 2 + 2 β 2 – s 2 ,

where 0 < s∗ < β . It can also be shown that Fs0 is a monotonically decreasing continuous function for any fixed x > 0, where Fs0 > 0 for u = 0 and Fs0 < 0 for s = β – 0. Therefore, equation Fs0 = 0, which determines the stationary point s∗ , always has a unique solution. As a result, we obtain J∼

exp(–iπ/4) exp(iχF (s∗ , x)) √ p , 2πχ |Fss00 (s∗ , x)|

χ → ∞.

(5.157)

Asymptotic estimates for two other integrals present in the kernel Q 1 (x) can be derived by using the asymptotic behavior of the Hankel function for large arguments (compare with Section 5.5). Then we finally arrive at the following estimate: i h √ √ 2 2 (5.158) Q1 ∼ χ–1/2 A1 (x)eiχF (s∗ ,x) + A2 (x)eiχβ 4+x + A3 (x)eiχ 4+x ,

where A1 (x), A2 (x), A3 (x) are some smooth functions free of the parameter χ. Further estimate of the integral on the right-hand side (5.152) will be based on the application of the method of integration by parts (see Section 1.4) well known in such cases. Since for x > 0 the phase function (4 + x2 )1/2 has no stationary point, such integration by parts shows that contribution to the considered tail from the terms related to the functions A 2


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and A3 has the order χ–1/2 . In order to establish the asymptotic behavior of the term related to A1 , let us first clarify that the function F (s∗ , x) has no stationary point (as a function of x) for x > 0. Indeed, the equation Fs0 (s , x) = 0 defines the stationary point s∗ = s∗ (x), so F (s∗ , x) = F [s∗ (x), x]. It follows from this relation that dF/dx = F s 0 s∗x + Fx0 = Fx0 = s = s∗ (x). We thus have discovered that dF/dx = 0 only for s∗ = 0, which is possible only at x = 0, which can be proved directly. It follows from the last fact that the contribution of the term with A1 to the right-hand side integral in Eq. (5.152) is also of the order of χ –1/2 . Omitting details of the Wiener–Hopf method for equation (5.152) and trivial treatment of the convolution equation (5.153), we can write out the final result p(x) =

iπl0 χ χ – β tan χβ 2G– (0) (n α X j G+ (αj )Hj (x, 1)

) nβ X j G+ (βj )Hj (x, β ) + , × 1 – exp[–2χσ1 (αj )] j=1 1 – exp[–2χσ2 (βj )] j=1 √ √ exp i (χβ )2 – (πj )2 (a + x) + exp i (χβ )2 – (πj )2 (a – x) , Hj (x, β ) = (χβ )2 – (πj )2

(5.159)

where G(s ) = G+ (s ) G– (s ) is a factorization of the function G(s ).

Helpful remarks Direct numerical solution of the initial equation (5.146) is faced with essential difficulties. This is caused by the evident remark that for χ 1 the kernel represents the sum of a delta-like and several strongly oscillating functions, which makes the calculation procedure very unstable. With increasing of the frequency parameter χ the contact pressure diagram becomes wave-like. This phenomenon, just as in the scalar acoustic case, can be explained by multiple re-reflections of rays from the bottom of the layer and, as noted in the previous sections, differs from analogous problems for the half-plane (Section 5.6), where the asymptotics of the contact pressure tends with the frequency increasing to a constant distribution.


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Chapter 6

Short-Wave Asymptotic Methods on the Basis of Multiple Integrals 6.1. Schoch’s Method: Exact Representation of 3D Wave Fields by One-Dimensional Quadratures It is conventionally recognized that modern computers permit, in principle, efficient calculation of wave fields by direct numerical methods for arbitrary boundary value problems. This is especially true for the cases of low and moderate frequencies. In the Preface to this book it was outlined why the high-frequency regime is so distinct from ordinary cases, when implemented on computers. Briefly speaking, this requires too large number of mesh nodes with the frequency increasing. This chapter will be devoted to various methods to achieve efficient calculation of high-frequency wave fields. In the previous chapters we could see a number of unexpected features of high-frequency wave processes. Another interesting property is also rather unexpected: some exact solutions written in explicit form are absolutely unserviceable for computer calculations. Example. Let us consider an elastic half-space z ≥ 0, whose boundary plane surface z = 0 is free of tangential load, and an oscillating (with angular frequency ω ) normal pressure of the amplitude p0 is applied to this surface over the rectangular domain |x| ≤ a, |y | ≤ b. Equations of motion (see Section 1.9) in this 3D problem have the following form: 2 2 ∂ ∂ ux ∂ 2 ux ∂uy ∂uz 2 2 2 2 ∂ ux + cs + + cp – c s + + ω 2 ux = 0, cp ∂x2 ∂y 2 ∂z 2 ∂x ∂y ∂z 2 2 ∂ ∂ uy ∂ 2 uy ∂uz ∂ux 2 2 2 2 ∂ uy (6.1) + cs + + cp – c s + + ω 2 uy = 0, cp ∂y 2 ∂z 2 ∂x2 ∂y ∂z ∂x ∂ ∂ 2 uz ∂ 2 uz ∂ux ∂uy ∂ 2 uz + c2s + + c2p – c2s + + ω 2 uz = 0. c2p 2 2 2 ∂z ∂x ∂y ∂z ∂x ∂y

The boundary conditions on the plane boundary are (z = 0): p0 , (x, y ) ∈ S = {|x| ≤ a} ∩ {|y | ≤ b}, σxz = σyz = 0, σzz = p0 (x, y ) = 0, (x, y ) ∈ R2 \ S .

(6.2)

The double Fourier transform with respect to the variables x, y , |x| < ∞, |y | < ∞, applied to these equations, leads to the Fourier images of the displacement components to the system of ordinary differential equations of the second order with constant coefficients c2s Ux00 + ω 2 – c2p s12 + c2s s22 Ux + c2p – c2s (–is1 ) –is2 Uy + Uz0 = 0, c2s Uy00 + ω 2 – c2p s22 + c2s s12 Uy + c2p – c2s (–is2 ) –is1 Ux + Uz0 = 0, (6.3) 2 2 2 2 00 2 2 2 2 2 0 0 s = s1 + s2 , cp Uz + ω – cs s Uz + cp – cs –is1 Ux – is2 Uy = 0,


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where Ux = Ux (s1 , s2 , z ), Uy = Uy (s1 , s2 , z ), Ux = Ux (s1 , s2 , z ) are the Fourier transforms of respective components of the displacement vector, and the sign of the derivative is of course related to the variable z . A solution of this system can be found by a standard operation with the characteristic polynomial. To this end, we should seek the solution in the form     Ux Ax  Uy  =  Ay  eαz , (6.4) Uz Az where the characteristic parameter α should be determined as a root of the characteristic polynomial. The latter is obtained by the substitution of the representation (6.4) into system (6.3), which leads to a 3 × 3 homogeneous linear algebraic system. Since we seek a nontrivial solution of this system, its principal determinant must be trivial. This yields the characteristic equation, which in the considered problem after some transformations can be reduced to the following simple form: (s 2 – α2 – ks )2 (s 2 – α2 – kp ) = 0, (6.5) where kp = ω/cp and ks = ω/cs are longitudinal and transverse wave numbers, respectively. As can be seen from Eq. (6.5), there are two simple α1 , α2 and two double α3 , α4 roots of this bi-cubic equation: q p (6.6) α1,2 = ±γ (s ), α3,4 = ±q (s ), γ (s ) = s 2 – kp2 , q (s ) = s 2 – ks2 .

It is easily seen from Eq. (6.6) that the general solution is a combination of some structures of the type (6.4), and in our case can be expressed as follows: Ux = iAs1 e–γ(s)z – Bq (s ) e–q(s)z , Uy = iAs2 e–γ(s)z – Cq (s ) e–q(s)z , –γ(s)z

(6.7) –q(s)z

Uz = Aγ (s ) e + (is1 B + is2 C ) e , where we keep only those characteristic values of α that provide the satisfaction of the radiation boundary conditions at z → +∞, that is, we take the sign minus in all pairs ±γ (s ), ±q (s ). The remaining unknown constants A, B , C also depend on the Fourier transform parameters (s1 , s2 ) and should be determined by satisfying the boundary conditions (6.2) over the boundary plane z = 0. Note that the components of the stress tensor participating in Eqs. (6.2) are expressed in terms of components of the displacement vector in the following way: ∂ux ∂uz σyz ∂uy ∂uz σxz 2 2 = cs + , = cs + , ρ ∂z ∂x ρ ∂z ∂y (6.8) ∂ux ∂uy ∂uz σzz = c2p + c2p – 2c2s + . ρ ∂z ∂x ∂y After some transformations, expressions (6.8) can be rewritten in Fourier images as (a tilde here denotes the corresponding Fourier transform of the stress tensor components) σ˜ xz = Ux0 – is1 Uz = –2is1 γA e–γ(s)z + s12 + q 2 B + s1 s2 C e–q(s)z , 2 ρcs σ˜ yz = Uy0 – is2 Uz = –2is2 γA e–γ(s)z + s1 s2 B + s22 + q 2 C e–q(s)z , 2 ρcs (6.9) ks2 0 ks2 σ˜ zz = 2 Uz + 2 – 2 is1 Ux + is2 Uy ρc2s kp kp = ks2 – 2s 2 A e–γz – 2iq (s1 B + s2 C ) e–qz . © 2005 by CRC Press LLC

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Now, satisfaction by these expressions to the boundary conditions (6.2), which, in Fourier transforms, become σ˜ xz (s1 , s2 , 0) = 0,

σ˜ yz (s1 , s2 , 0) = 0,

σ˜ zz (s1 , s2 , 0) = P0 (s1 , s2 ),

(6.10)

leads to a 3 × 3 linear algebraic system to find coefficients A, B , C . By solving this system and taking into account that for the uniformly distributed applied constant normal oscillating pressure p0 ≡ const Z aZ b sin(as1 ) sin(bs2 ) , (6.11) P0 (s1 , s2 ) = p0 ei(s1 x+s2 y) dx dy = 4p0 s1 s2 –a –b we then finally arrive at the following exact analytical representations for the most interesting physical quantities, the components of the stress tensor given explicitly by Z Z (e–γz – e–qz ) sin(as1) sin(bs2) –i(xs1+ys2) 2ip0 ∞ ∞ γ (2s 2 – ks2) e ds1 ds2, σxz(x, y , z ) = 2 π –∞ –∞ s2 ∆(s ) Z ∞Z ∞ (e–γz – e–qz ) sin(as1) sin(bs2) –i(xs1+ys2) 2ip0 γ (2s 2 – ks2) e ds1 ds2, (6.12) σyz(x, y , z ) = 2 π –∞ –∞ s1 ∆(s ) Z Z p0 ∞ ∞ [(2s 2 – ks2)2e–γz –4s 2γqe–qz ] sin(as1) sin(bs2) –i(xs1+ys2) e ds1 ds2, σzz(x, y , z ) = 2 π –∞ –∞ s1 s2 ∆(s ) where ∆(s ) = (2s 2 – ks2 )2 – 4s 2 γq is again the classical Rayleigh function. Calculation of integrals (6.12) is not a simple task. In the high-frequency range, where kp a 1, kp b 1, this requires too long computation time, even on modern powerful computers. In order to overcome the troubles of such sort Schoch (1941) proposed a method that allows, for a domain S of arbitrary shape, the reduction of the considered problem without any approximation to the calculation of a single integral (instead of the double one) over a finite-length interval — a fantastic result! His idea is as follows. Let us first study small-amplitude oscillations (linear theory) of a rigid plate on the free surface of the scalar acoustic half-space z = 0 (see Fig. 6.1).

r S

r

r1(j) j M' z M

Figure 6.1. On the calculation of the wave field generated by a plane oscillator If the amplitude of oscillations of the plane vibrator is v 0 ≡ const, (x, y ) ∈ S then the problem is reduced to the Helmholtz equation for acoustic pressure ∆p(x, y , z ) + k 2 p(x, y , z ) = 0,

z ≥ 0.

(6.13)

If the oscillator is fastened with a rigid screen occupying domain z = 0, (x, y ) ∈ R 2 \ S , then the boundary condition is iωρ0 v0 , (x, y ) ∈ S , ∂p = g (x, y ) = (6.14) z = 0: ∂z 0, (x, y ) ∈ R2 \ S .


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Let us denote, as in Chapter 2, the Green’s function for full 3D acoustic space by Φ(r ) =

eikr , 4πr

r=

p (x – x0 )2 + (y – y0 )2 + z 2 ,

(6.15)

where the point M with Cartesian coordinates (x0 , y0 , z ) is the observation point. Then, recalling the boundary property of the normal derivative of the single-layer potential (see Section 2.1), we can see that the exact solution of the posed problem can be expressed by the following formula: ZZ ikr ZZ ∞ e iωρ0 v0 , (6.16) p(M ) = –2 g (x, y ) Φ(r ) dx dy = – 2 π S r –∞

because the boundary value of the normal derivative of this expression is ZZ ∞ p ∂Φ(ρ, 0) dx dy + g (x0 , y0 ), ρ = (x – x0 )2 + (y – y0 )2 . (6.17) –2 g (x, y ) ∂z –∞ The last integral is equal to zero since ∂Φ(r )/∂z = ∂Φ(r )/∂r (r ⋅ n(x0 , y0 )) = 0 and because r ⊥ n when z = 0. Schoch’s idea works in the following manner. Let us place the origin of the cylindrical coordinate system at the vertical projection of the observation point, i.e., at the point M 0 (x0 , y0 , 0) (see Fig. 6.1). For simplicity, we consider here only the case when M 0 ∈ S . If M 0 lies outside the domain S all transformations can be repeated in a similar way. Now, formula (6.16) reads Z Z ikr Z 2π Z ρ1 (ϕ) ik√ρ2 +z2 e iωρ0 e iωρ0 √ 2 v0 =– v0 dϕ ρ dρ. (6.18) p(M ) = – 2π 2π ρ + z2 S r 0 0 Now we may apply the change of variable (see Fig. 6.1): r 2 = ρ2 + z 2 , dx dy = ρ dρ dϕ, r dr = ρ dρ. Then Eq. (6.18) becomes equal to Z 2π Z r1 (ϕ) Z 2π ωρ0 1 iωρ0 ikr ikz ikr1 (ϕ) v0 dϕ e dr = v0 e – e dϕ , (6.19) p(M ) = – 2π k 2π 0 0 z where r1 (ϕ) = (z 2 + ρ21 )1/2 . Our purpose is achieved: the 3D wave field is (absolutely exactly) expressed in terms of a one-dimensional integral over the interval of finite length. In some simple cases Eq. (6.19) permits analytical calculation of the full wave field. For example, if S is a disk of radius a, then ωρ0 ikz ik√a2 +z2 v0 e – e . (6.20) p(M ) = k

Helpful remarks We remarked above that the calculation of integrals of the type (6.12) for high frequencies is a hard problem, even when implemented on computers. In some cases this can be treated approximately. For instance, in the far field such integrals can be estimated asymptotically (the so-called Fraunhofer approximation). However, for small distances it is very difficult to obtain any result with acceptable accuracy by direct numerical treatment of the integral (6.12). This trouble arises even in the scalar acoustic problem, but in elastic case the integrand becomes more complex and the problem becomes more difficult.


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6.2. High-Frequency Wave Fields in Elastic Half-Space Schoch’s approach can be spread to more complex problems. First of all, let us expand it to the scalar problem when on the boundary plane surface of the acoustic half-space there is applied a uniformly distributed pressure p(x, y , 0) = p 0 (x, y ), (x, y ) ∈ S over the domain S , so that p(x, y , 0) = 0, (x, y ) ∈ R2 \ S . Then we can see that the exact solution here can be represented by the double-layer potential (see Section 2.1) ikr ZZ ZZ ∞ p0 ∂ e ∂Φ(r ) dx dy = dx dy , p(M ) = 2 p(x, y ) ∂z 2π r (6.21) S ∂z –∞ p 2 2 2 r = (x – x0 ) + (y – y0 ) + z ,

where M (x0 , y0 , z ) is the observation point. Correctness of the representation (6.21) is due to the boundary property of the potential of double layer (Section 2.1), which is ZZ ∞ p ∂Φ(ρ, 0) dx dy + p(x0 , y0 ), ρ = (x – x0 )2 + (y – y0 )2 . (6.22) 2 p(x, y ) ∂z –∞ The last integral is equal to zero since ∂Φ(r )/∂z = ∂Φ(r )/∂r (r ⋅ n(x, y )) = 0 and because r ⊥ n when z = 0. Equation (6.21) yields ZZ ikr Z 2π Z r1 (ϕ) e p0 ∂ p0 ∂ dx dy = dϕ eikr dr p(M ) = 2π ∂z r 2 π ∂z S 0 z (6.23) Z 2π Z 2π ikr1 (ϕ) 1 z e p0 ∂ ikz ikr1 (ϕ) ikz e dϕ = –p0 e – dϕ . e – = – ik ∂z 2π 0 2π 0 r1 (ϕ) The problem becomes much more complex if the medium cannot be considered as an acoustic one. Thus, the solution for the normal point force applied to the free surface of elastic half-space cannot be expressed in elementary form, but is expressed in terms of the Fourier–Bessel integral. In order to arrive at this representation, we can apply the same technique as in the previous section, in the case p 0 (x, y ) = p0 δ (x – x0 , y – y0 ). Then P0 (s1 , s2 ) = p0 ei(x0 s1 +y0 s2 ) , so for example, for the σzz stress component we have ZZ ∞ (2s 2 – ks2 )2 e–γz – 4s 2 γq e–qz –i[(x–x0 )s1 +(y–y0 )s2 ] p0 0 e ds1 ds2 . (6.24) σzz (x, y , z ) = 4π 2 ∆(s ) –∞ If we apply here the change of variables x – x0 = ρ cos ψ , x – x0 = ρ sin ψ , s1 = s cos µ, s2 = s sin µ, and pass from the Cartesian to a cylindrical coordinate system, then we can see that expressions under the double integral in the polar coordinate system become (x – x0 )s1 + (y – y0 )s2 = sρ cos(µ – ψ ). Since all other terms in the integrand depend only on s = (s12 + s22 )1/2 , it can easily be seen that the integration with respect to s leads to Z ∞ Z 2π (2s 2 – ks2 )2 e–γz – 4s 2 γq e–qz p0 0 s ds e–isρ cos(µ–ψ) dµ. (6.25) σzz (ρ, z ) = 4π 2 0 ∆(s ) 0 Since for arbitrary periodic integrand the integral may be taken over any interval of the length equal to the period of this function, it can be directly proved that expression (6.25) is free of the polar angle ψ : Z 2π Z 2π e–isρ cos(µ–ψ) dµ = e–isρ cos µ dµ = 2πJ0 (sρ) (6.26) 0

0


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(J0 is the Bessel function), so 0 σzz (ρ, z ) =

p0 2π

Z

∞ 0

(2s 2 – ks2 )2 e–γz – 4s 2 γq e–qz s J0 (sρ) ds . ∆(s )

(6.27)

Our basic idea is the following. We can estimate this integral asymptotically for high frequencies. The stationary phase method (see Section 1.4) can be applied to this integral, after the change of variable s = kpes (the first exponent in (6.27)) or s = kpe s (the second exponent), when we arrive at some exponentially oscillating function with a phase, containing a large frequency parameter. This also appears in the argument of the Bessel function, and an appropriate asymptotic formula should be taken to reduce the integral to the form containing only purely exponential oscillating functions. The principal2 contribution of the considered integral is given by the stationary point s = ρ/r = (x – x0 ) +(y – y0)2 1/2/ (x – x0 )2 +(y – y0)2 + z 2 1/2 , which permits the following asymptotic estimate: ρ i ρ ip0 z h iks r ikp r 0 + k e f , k e f σzz (ρ, z ) ∼ – s 2 p 1 2πr 2 r r (2s 2 – δ 2 )2 √ √ , β = kp /ks = cs /cp < 1, f1 (s ) = (6.28) (2s 2 – δ 2 )2 + 4s 2 1 – s 2 δ 2 – s 2 √ √ 4s 2 1 – s 2 β 2 – s 2 √ δ = 1/β . f2 (s ) = √ (2s 2 – 1)2 + 4s 2 1 – s 2 β 2 – s 2 If now the oscillator generates a uniformly distributed normal load p 0 applied to the free plane boundary surface over the domain S , then Eq. (6.28) allows us to operate with the Schoch method. Indeed, due to the linearity of the problem, the total wave field at an arbitrary point inside elastic half-space can be calculated as a superposition of the contributions from all point elementary forces of the type (6.28). This results in the following asymptotic formula: √ 2 2 √ 2 2 Z Z r1 (ϕ) dr r –z r –z ip0 z 2π ikp r iks r 0 dϕ kp e f1 + k s e f2 . σzz (ρ, z ) = – 2π 0 r r r z (6.29) In some sense this integral is an alternative representation of the integral given by Eq. (6.12). Unfortunately, in contrast with scalar acoustic case internal integral in (6.29) cannot be calculated analytically, so we would like to estimate it asymptotically. Such an approach is not equivalent to asymptotic approximation of integrals of the type (6.12), since here we have the only large dimensionless parameter k p r (or ks r ), which determines how many times the distance from the current applied point source is greater than the wavelength. By contrast, the asymptotic estimate of the integral (6.12) will encounter the two large dimensionless geometric parameters: the relative (with respect to the wavelength) distance of the observation point from the applied load, and relative (again compared with the wavelength) geometric size of the domain S . With the frequency increasing these parameters compete with each other. Domination of one of them relative to the other generates the far-field (Fraunhofer) approximation. In the opposite case we arrive at the so-called Fresnel approximation. Our approach, based upon brilliant ideas of Schoch, will lead us to a uniform high-frequency approximation. Besides, as we will see below, in the scalar acoustic case this approach gives exact formulas for the wave field. An efficient high-frequency representation can be obtained if we apply to the integral (6.29) a standard integration by parts. If we restrict ourselves by the leading asymptotic


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term only, then we arrive at the following asymptotic representation: " ! !#r1 (ϕ) √ √ Z 2π 2 – z2 2 – z2 dϕ r r p z 0 0 eikp r f1 + eiks r f2 σzz (ρ, z ) = – 2π 0 r r r ! !# z=r ) (6.30) ( √ √ Z 2π " 2 2 2 2 dϕ r –z r –z z eikp r f1 + eiks r f2 . = p0 eikp z – 2π 0 r r r1 (ϕ) The problem of efficient calculation of the high-frequency wave field in elastic half-space is thus reduced, by the Schoch method to a one-dimensional integral over finite integral. In some simple cases this again can be calculated analytically. For instance, for round disk oscillator of radius a we obtain from (6.30) √ √ a a z 2 2 2 2 eikp a +z f1 √ + eiks a +z f2 √ . σzz = p0 eikp z – √ a2 + z 2 a2 + z 2 a2 + z 2 (6.31) Let us note that in the far zone, as it follows from the structure of the functions f 1 and f2 , the last expressions (6.30) and (6.31) degenerate into the corresponding formulas for acoustic half-space. In the near-field they provide a high accuracy of the order of only a few percent, as was tested by numerous numerical simulations.

Helpful remarks Some examples on the application of the method proposed in this section are quoted in Sumbatyan (1988). The accuracy of our approach is around 1% of relative error, and this works in real time when implemented on PC. From our experience, direct numerical treatment of integrals of the type (6.12) requires approximately 10 2 to 103 times more computational time than the approach proposed here. A survey of existing methods of elastic wave fields calculation can be found in Achenbach (1973).

6.3. Asymptotic Nature of the Geometrical Diffraction Theory As in Kirchhoff’s physical diffraction theory you can find various formulations of the geometrical diffraction theory treated by different authors. A good survey can be found in Hönl et al. (1961, more physical treatment) and in Babich and Buldyrev (1989, more mathematical treatment). In two words, Kirchhoff’s theory gives a leading high-frequency asymptotic term of the solution on the boundary of the obstacle. Analogously, the geometrical diffraction theory gives a leading high-frequency asymptotic term of the solution at an arbitrary point of observation. The foundations of this theory were established by Keller (1962), but in this section we give an alternative derivation of some of his classical results, starting (more naturally) from Kirchhoff’s solution. To be more specific, let us consider a point source x0 of a (2D) acoustic wave which is incident on an acoustically hard obstacle D with the boundary line l. The Kirchhoff– Helmholtz integral formula gives an exact representation for the solution at arbitrary observation point x (see Section 2.2) Z i ∂Φ(kr ) dly , Φ(kr ) = H0(1) (kr ), r = |x – y |, x ∈ R2 \ D , (6.32) psc (x) = p(y ) ∂ny 4 l


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where psc (x) is a scattered wave field at the point x, and p(y ), y ∈ l, is a full wave pressure on the boundary contour l. According to Kirchhoff’s physical diffraction theory, the leading asymptotic term for p(y ), y ∈ l is given as follows (see Section 2.6): inc 2p (y ), y ∈ l+ , p(y ) = k → ∞. (6.33) 0, y ∈ l– , Since the incident wave, given by the same (point) Green’s function expressed again in terms of the Hankel function, has also an asymptotic representation, we may assume that eikr0 pinc (y ) = √ , r0

k → ∞,

r0 = |x0 – y |.

(6.34)

Now, taking into account that ∂Φ(kr ) = e–πi/4 2 ∂ny

r

k eikr √ cos(ny , r ), 2π r

k → ∞,

(6.35)

and collecting together Eqs. (6.32)–(6.35), we can see that at k → ∞ : r Z cos(ny , r ) ikS(y) k sc –πi/4 e dly , S (y ) = r0 + r = |x0 – y | + |x – y |. (6.36) p (x) = e √ 2π l+ rr0 THEOREM 1. Any point y0 ∈ l+ with nd y r = nd y r0 = θ supplies a stationary point for the phase function S (y ).

Proof. Let such a point y0 ∈ l+ exist, and we consider the current point y ∈ l + from a small neighborhood of the point y0 . Then we will prove that y0 is the only stationary point among all y from this neighborhood. To this end, we consider the distance |x 0 – y | as a side of the triangle formed by the three sides L0 = |x0 – y0 |, r = |x0 – y |, ∆s = |y – y0 |, and use a cosine theorem: |x0 – y |2 = L20 + (∆s )2 – 2L0 (∆s ) cos α,

α = Ld 0 ∆s .

(6.37)

Let us recall some classical results from differential geometry (see, for example, Pogorelov, 1978).

x x0 L

n L 0 q q a y0 b y Ds r r O

Figure 6.2. Single ray reflection from obstacle with a perfect boundary


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For the pair of the points y and y0 (see Fig. 6.2) situated on the smooth curve l + close to each other, there is an osculating circle whose radius ρ determines the curvature of the curve at the point y0 . Then geometry of the triangle Oyy0 determines the angle β as follows: β = π/2–arcsin[(∆s/2)/ρ] = π/2– ∆s/(2ρ)+ O [(∆s ) 3 ]. As follows from Fig. 6.2, α = π – θ – β = π/2 – θ + ∆s/(2ρ), so ∆s ∆s 3 + O [(∆s ) ] = sin θ – cos θ + O (∆s )2 . (6.38) cos α = sin θ – 2ρ 2ρ

The last relation allows us to express the distance |x 0 – y |2 in Eq. (6.37) in terms of (∆s ): ∆s cos θ – sin θ + O (∆s )3 . (6.39) |x0 – y |2 = L20 + (∆s )2 + 2L0 (∆s ) 2ρ By extracting the square root from both sides (6.39), we arrive at the following relation: (∆s )2 cos2 θ cos θ + + O (∆s )3 . (6.40) |x0 – y | = L0 – ∆s sin θ + 2 L0 ρ By analogy, we can find that a similar formula holds for the second term in the phase function (6.36): (∆s )2 cos2 θ cos θ + + O (∆s )3 , (6.41) |x – y | = L + ∆s sin θ + 2 L ρ so the phase function is 1 1 2 (∆s )2 2 cos θ + + + O (∆s )3 . (6.42) S (y ) = |x0 – y | + |x – y | = L0 + L + 2 L0 L ρ cos θ Note that linear term (∆s ) is absent here, so the point of specular reflection is indeed the stationary point of the phase function S (y ) (see Section 1.4). The theorem is proved. THEOREM 2. If there is the only point y0 ∈ l+ of the specular reflection on the contour l + and expression in the parentheses in Eq. (6.42) is not equal to zero, then the leading asymptotic term of the scattered wave psc (x) in Eq. (6.36) is psc (x) =

exp{i[k (L0 + L) + (π/4)(δ – 1)]} p , |L0 + L + 2L0 L/(ρ cos θ )|

δ = sign[L0 + L + 2L0 L/(ρ cos θ )]. (6.43)

The proof directly follows from the stationary phase method (see Section 1.4), since at the stationary point we have nd y r = nd y r0 = θ , r = L, r0 = L0 , S (y ) = r0 + r = L0 + L. The solution expressed by formula (6.43) is called a geometrical diffraction theory solution. It is interesting to note that the obtained results implicitly contain solutions not only for a point-source incident wave but also for a plane incident wave. This case can be obtained from our solution (6.43) by putting L0 L, so the scattered wave pressure is psc (x) =

exp{i[k (L0 + L) + (π/4)(δ – 1)]} , √ p L0 |1 + 2L/(ρ cos θ )|

δ = sign[1 + 2L/(ρ cos θ )].

(6.44)

If, in addition to the listed assumptions, we consider also a far-field scattered wave field with L0 , L ρ, then Eq. (6.43) reads exp{i[k (L0 + L) + (π/4)(δ – 1)]} p √ ρ cos θ, δ = sign(ρ). (6.45) psc (x) = 2L0 L We thus can conclude that the amplitude in the far-field reflected plane acoustic wave is proportional to the square root of the curvature radius.


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Helpful remarks 1◦ . The proved Theorem 1 asserts that the first reflection law of the linear wave theory, which states that the reflection angle is equal to the angle of incidence, well known for plane reflectors is also valid for arbitrary smooth curvilinear boundary. 2◦ . If there are several stationary points, i.e., several points of specular reflection on the reflecting boundary, then the total contribution to the scattered wave field is given as the sum of individual contributions of these (stationary) points. 3◦ . The leading asymptotic term for the scattered wave field derived in the present section is valid only if the stationary point is regular, i.e., if expression in the parentheses in Eq. (6.42) does not vanish and so the term with the small length increment of the second order (∆s ) 2 is not trivial. Otherwise, when the stationary point is degenerated and this factor is zero, expression in the denominator vanishes and more detailed analysis shows that in such cases the leading asymptotic term is of a higher order with k → ∞.

4◦ . It perhaps seems to be a little strange that the denominator of Eq. (6.43) may vanish in some cases. Note that the curvature radius ρ at the reflection point should be taken with its actual sign—positive for convex obstacles and negative for concave ones. In the case of nonconvex boundary with a negative curvature the considered expression may turn out zero. A bright example is given by illumination of a circle with radius a from a point source located in its center, when the observation point x and source point x 0 are taken at the same place. In this case we have L = L0 = a, ρ = –a, θ = 0, cos θ = 1, and hence L0 + L + 2L0 L/(ρ cos θ ) = 0. 5◦ . Results of this section can explain why Kirchhoff’s physical diffraction theory gives correct prediction only for convex obstacles. Indeed, from the point of view of geometrical diffraction theory the physical diffraction theory is related to a direct incidence of the incoming falling wave. If the boundary of the obstacle is convex, then this can guarantee that the only wave falling to a small neighborhood of the point y 0 ∈ l+ is a (directly) incident one. If the boundary is not convex, then in addition to a directly falling incident wave there may exist also waves re-reflected from other boundary points and arriving, after these re-reflections, to the same neighborhood of y 0 . 6◦ . Formula (6.43) admits a simple testing in the case of plane reflecting boundary. Heuristically, if a wave radiated from a point x0 and represented by formula (6.34) propagates in the space and then encounters any plane reflecting acoustically hard surface, then simple mirror ideas allow us to suppose that the amplitude of this wave behaves as in the case where there is no reflector and the amplitude of re-reflected wave depends on the total flight of the acoustic ray. We thus can predict from heuristic considerations that the solution must in this case be of the following form (compare with the form of the radiated pressure (3)): eik(L0 +L) . p(x) = √ L0 + L

(6.46)

This result directly follows from our formula (6.43), since in the case of plane reflector ρ = ∞, δ = sign(L0 + L) = 1.

7◦ . It should be noted that the results obtained in this section, in principle, can be obtained within the framework of Keller’s approach to the geometrical diffraction theory (see Keller, 1962; Borovikov and Kinber, 1994; McNamara et al., 1990). However, because of the number of re-reflections, it is too hard to trace the acoustic beams history. We believe that our approach leads to the final results more directly and more clearly.


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6.4. High-Frequency Diffraction with Re-Reflections Let us consider repeated (double) reflection from acoustically hard boundary (see Fig. 6.3).

x0 L0

n1 h x C A

q1 L1

n2 L2

y20

l1 y10 Ds1

D B

q2 q2

Ds2

+

q1

a +

l2

x

Figure 6.3. Repeated ray reflection from obstacle with a perfect boundary A scalar acoustic wave is emitted from the point source x 0 and then received at the point x after re-reflections from arcs l1+ and l2+ of obstacle’s boundary. According to Kirchhoff– Helmholtz integral formula, the received scattered pressure is given by Z ∂Φ(kr ) sc dl2 , r = |x – y2 |, (6.47) p (x) = p(y2 ) ∂n2 l+2 where p(y2 ) is the value of acoustic pressure on the arc l2+ . For high frequencies, according inc to physical diffraction theory, p(y2 ) = 2pinc 2 (y2 ), where p2 (y2 ) is a wave field falling to the points y2 ∈ l2+ of the contour l2+ after the first scattering happened on the contour l 1+ . The latter is determined again by the Kirchhoff–Helmholtz integral formula as follows: Z ∂Φ(kr1 ) inc dl1 , r1 = |y2 – y1 |, y2 ∈ l2+ , (6.48) p2 (y2 ) = p(y1 ) ∂n + 1 l1 where p(y1 ) is the value of acoustic pressure on the contour l1+ . According to the physical diffraction theory, p(y1 ) = 2pinc (y1 ), where pinc (y1 ) is a wave field incident on the contour l1+ . Taking into account that pinc (y1 ) is expressed for high frequencies again by the formula eikr0 pinc (y ) = √ , r0

k → ∞,

r0 = |x0 – y1 |,

(6.49)

and collecting together all relations (6.47)–(6.49), we can write out the following basic representation valid at k → ∞ : Z Z cos(n1 , r1 ) cos(n2 , r ) ikS(y1 ,y2 ) ik sc e dl1 dl2 , p (x) = – √ 2π l+1 l+2 r0 r1 r (6.50) S (y1 , y2 ) = r0 + r1 + r = |x0 – y1 | + |y2 – y1 | + |x – y2 |.

THEOREM. Any pair of points y10 ∈ l1+ and y20 ∈ l2+ with nd d d 1 r1 = n 1 r0 = θ1 and n 2r = nd 2 r1 = θ2 supplies a stationary value for the phase function S (y 1 , y2 ).

Proof. Let such a pair exist and for these values of y1 = y20 , y1 = y20 the quantities participating in the phase function take the following values: r 0 = L0 , r1 = L1 , r = L2 , S (y1 , y2 ) = L0 + L1 + L2 . Then we consider the current point A = y1 ∈ l1+ from a small


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neighborhood of the point y10 and analogous point B = y2 ∈ l2+ from a small neighborhood of the point y20 . Similarly to the derivation procedure of the previous section we can easily obtain (∆s1 )2 cos2 θ1 cos θ1 + + O (∆s1 )3 , |x0 – y1 | = L0 + ∆s1 sin θ1 + 2 L0 ρ1 2 (6.51) 2 cos θ2 cos θ2 (∆s2 ) 3 + + O (∆s2 ) . |x – y2 | = L2 – ∆s2 sin θ2 + 2 L2 ρ2 The key problem is to evaluate the third distance |y 1 – y2 |. To this end, we introduce the Cartesian coordinate system (ξ , η ) with the origin chosen at the point y 20 . As follows from Fig. 6.3 and the consideration in Section 6.3, AC = (∆s2 ) sin(π/2 – β )[1 + O (∆s2 )] (∆s2 )2 + O (∆s2 )3 , = (∆s2 ) sin{arcsin[∆s2 /(2ρ2 )]} [1 + O (∆s2 )] = (6.52) 2ρ2 2 (∆s1 ) + O (∆s1 )3 BD = 2ρ1 (cf. Fig. 6.2). Then in the introduced Cartesian coordinate system (ξ , η ) coordinates of the considered points can be expressed as follows: (∆s2 )2 + O (∆s2 )3 , ξ10 = L1 sin θ2 , η10 = L1 cos θ2 , ξA = –∆s2 , ηA = –AC = – 2ρ2 (∆s1 )2 (6.53) sin α + O (∆s1 )3 , ξB = ξ10 – ∆s1 cos α + BD sin α = ξ10 – ∆s1 cos α + 2ρ1 (∆s1 )2 cos α + O (∆s1 )3 . ηB = η10 – ∆s1 sin α – BD cos α = η10 – ∆s1 sin α – 2ρ1 Therefore, ( 2 (∆s1 )2 2 2 1/2 sin α + ∆s2 |y1 – y2 | = (ξB – ξA ) + (ηB – ηA ) = ξ10 – ∆s1 cos α + 2ρ1 2 )1/2 (∆s2 )2 (∆s1 )2 0 cos α + + O (∆s )3 , + η1 – ∆s1 sin α – 2ρ1 2ρ2 (∆s )2 = min (∆s1 )2 , (∆s2 )2 , (∆s1 ) (∆s2 ) . (6.54) Raising to the square in the braces here and keeping all powers up to the second one only, we can obtain |y1 – y2 | = L21 – 2∆s1 ξ10 cos α + η10 sin α + (∆s1 )2 + (∆s2 )2 – 2 (∆s1 ) (∆s2 ) cos α (6.55) 2 1/2 (∆s1 )2 0 0 0 0 (∆s2 ) 3 ξ1 sin α – η1 cos α + 2ξ1 ∆s2 + η1 + O (∆s ) . + ρ1 ρ2 Note that α = π – θ1 – θ2 ; hence ξ10 cos α + η10 sin α = L1 sin(α + θ2 ) = L1 sin θ1 , (6.56) ξ10 sin α – η10 cos α = –L1 cos(α + θ2 ) = L1 cos θ1 , cos α = – cos(θ1 + θ2 ).


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As a result, expression (6.55) can be written as follows: L1 ∆s2 (∆s1 )2 ∆s1 1 + sin θ1 + 2 sin θ2 + cos θ |y1 – y2 | = |AB | = L1 1 – 2 1 L1 L1 L21 ρ1 (6.57) 1/2 L1 (∆s1 ) (∆s2 ) (∆s2 )2 cos θ2 + 2 cos(θ1 + θ2 ) + O (∆s )3 . 1+ + L21 ρ2 L21

Now, Taylor expansion of this root square expression in powers of the small quantities (∆s1 ) and (∆s2 ) leads to the following representation: ∆s1 ∆s2 (∆s1 )2 L1 |y1 – y2 | 2 =1– sin θ1 + sin θ2 + 1+ cos θ1 – sin θ1 L1 L1 L1 2L21 ρ1 L1 (∆s1 ) (∆s2 ) (∆s2 )2 (6.58) 2 1+ cos θ2 – sin θ2 + + [cos(θ1 + θ2 ) + sin θ1 sin θ2 ] 2 2 2L1 ρ2 L1 3 + O (∆s ) .

Further, taking together all results (6.51) and (6.58) we can conclude that the phase function S (y1 , y2 ) in Eq. (6.50) is 1 2 1 (∆s1 )2 2 + + cos θ1 S (y1 , y2 ) = L0 + L1 + L2 + 2 L0 L1 ρ1 cos θ1 (6.59) 1 2 (∆s1 ) (∆s2 ) 1 (∆s2 )2 cos2 θ2 + + + cos θ1 cos θ2 + O (∆s )3 . + 2 L2 L1 ρ2 cos θ2 L1

This expression is free of linear terms (∆s1 ) and (∆s2 ). This proves the theorem. From results of this theorem, by applying the two-dimensional stationary phase method (see Section 1.4), we can derive the leading asymptotic term for the amplitude of the doubly re-reflected acoustic ray. This is based on the following expressions for the second-order derivatives: 1 1 2 ∂2S cos θ1 cos θ2 ∂2S 2 = cos θ + + , = , 1 ∂ (∆s1 )2 L0 L1 ρ1 cos θ1 ∂ (∆s1 )∂ (∆s2 ) L1 (6.60) 1 1 2 ∂2S 2 = cos θ2 + + , ∂ (∆s2 )2 L2 L1 ρ2 cos θ2 and, according to Section 1.4, has the following form: p(x) =

exp{i[k (L0 + L1 + L2 ) + (π/4)(δ2 – 2)]} √ √ , L0 L1 L2 | det(D2 )|

k → ∞.

Here D2 is the Hessian (a certain factor is omitted in front of all its elements)   1 2 1 1 + +   L0 L1 ρ1 cos θ1 L1 , D2 =    1 1 2 1 + + L1 L2 L1 ρ2 cos θ2

(6.61)

(6.62)

and δ2 = sign(D2 ) = ν+ (D2 ) – ν– (D2 ) is a difference between the number of positive and negative eigenvalues of the matrix D2 .


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Let us consider the generic case when the number of reflections is arbitrary being equal to N . Let the acoustic wave be emitted from the point x 0 , then re-reflected N times, from 0 0 + the points ym , ym ∈ lm , m = 1, . . . , N , sequentially, and finally received at the receiving point x. All these reflecting points are situated so that the incident and reflected angles at 0 the reflection point ym are equal to each other being θm . Let us denote the distance between 0 0 |, m = 1, . . . , N – 1. Besides, we denote the – ym sequential reflecting points as Lm = |ym+1 distance between the source x0 and the first point of reflection y10 as L0 = |y10 – x0 |, and the 0 0 distance between the receiver x and the last reflection point y N as LN = |x – yN |. It follows from the previous consideration that the leading asymptotic term for the amplitude of N -times re-reflected acoustic wave can be constructed by a high-frequency estimate of the multiple integral N Q N/2 Z cos θm Z k m=1 s ··· eikS dl1 . . . dlN , p(x) ∼ e–(π/4)N i 2π + + N lN l1 Q Lm (6.63) m=0

S = S (y1 , . . . , yN ) = |y1 – x0 | +

N –1 X

|ym+1 – ym | + |x – yN |.

m=1

By analogy with the above case N = 2, we can conclude that 1 1 2 ∂2S 2 = cos θm + + , m = 1, . . . , N , ∂ (∆sm )2 Lm–1 Lm ρm cos θm cos θm cos θm+1 ∂2S (6.64) = , m = 1, . . . , N – 1, ∂ (∆sm )∂ (∆sm+1 ) Lm ∂2S = 0, n ≠ m, n ≠ m + 1. ∂ (∆sm )∂ (∆sn ) Hence, the multiple stationary phase method (Section 1.4) leads to the explicit formula for the (geometric optical) leading asymptotic term: N P exp i k Lm + (π/4)(δN – N ) m=0 s , k → ∞, (6.65) p(x) = N Q Lm |det(DN )| m=0

where

DN = (dnm ) (n, m = 1, . . . , N ), 1 2 1 + + (m = 1, . . . , N ), dmm = Lm–1 Lm ρm cos θm 1 (m = 1, . . . , N – 1), dm+1,m = dm,m+1 = Lm dnm = 0 (n ≠ {m, m + 1, m – 1}), δN = sign(DN ).

(6.66)

Helpful remarks 1◦ . As in the case of single reflection, singular cases can occur in (6.66), too. This happens every time when the denominator is equal to zero, det(D N ) = 0. In such cases the leading asymptotic term has a different qualitative behavior, and its dependence on the frequency parameter k is absolutely different.


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2◦ . Recall that in the present section we studied only the case of acoustically hard boundary. It can easily be proved that in the case when all reflecting lines are acoustically soft, the generic formula (6.65) should be multiplied by (–1) N . Heuristically, this directly follows from the evident observation that at each reflection the phase changes its value by 180 ◦ in jump. 3◦ . It is important to stress the case when δN = sign(DN ) can be calculated in a simple way. Let us assume that all ρm > 0, m = 1, . . . , N , i.e., all reflecting curves are convex. In this case the matrix DN can be represented as the sum of two positive definite matrices (for definition, see Horn and Johnson, 1986; and Section 4.1, which main results can be applied also to the case of finite-dimensional space): DN = A N + B N , 1 1 + amm = Lm–1 Lm

AN = (anm ),

BN = (bnm ),

(m = 1, . . . , N ),

1 (m = 1, . . . , N – 1), Lm anm = 0 (n ≠ {m, m + 1, m – 1}), 2 (m = 1, . . . , N ), bnm = 0 (n ≠ m). bmm = ρm cos θm am+1,m = am,m+1 =

(6.67)

Positive definiteness of the matrix BN is obvious, since it has a diagonal structure with diagonal entries all positive. Let us prove that the matrix A N is positive definite too. First of all, let us prove by induction that

IN = det(AN ) =

N P

m=0 N Q

Lm

.

(6.68)

Lm

m=0

For N = 1 we have I1 = det(A1 ) = 1/L0 + 1/L1 = (L0 + L1 )/(L0 L1 ), so in this case the correctness of Eq. (6.68) is evident. For N = 2 we have 1 1 1 L +L 1 1 1 L0 + L 1 + L 2 L 0 1 1 + + = , (6.69) I2 = det(A2 ) = = L0 L1 L2 1 1 L0 L1 L0 L2 L1 L2 1 + L1 L1 L2 which coincides again with formula (6.68). For arbitrary N > 2 we give a proof by induction. Let us assume that Eq. (6.68) is valid for all M < N , and we consider I N , N > 2. Then, expanding this determinant by elements of the first column, we obtain

IN =

1 1 + IN –1 – 2 IN –2 = LN –1 LN LN –1 1

1

1 + LN –1 LN

N –1 P

m=0 N Q–1

Lm

– Lm

1 L2N –1

m=0

(1/LN –1 + 1/LN ) =

N –1 P

Lm – (1/LN –1)

m=0 N Q–1

m=0

N –2 P

m=0

Lm

=

N –1 P

m=0 N Q–1

m=0

m=0 N Q–2

Lm Lm

m=0

1 + (1/LN )

Lm

N –2 P

Lm

Lm

=

N P

m=0 N Q

(6.70) Lm

. Lm

m=0


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Now, when relation (6.68) is strictly proved, we will show that the matrix A N is positive definite, by using the well known Sylvester criterion (see Horn and Johnson, 1986). It claims that any symmetric matrix is positive definite if all its principal determinants (i.e., the ones arranged by cancelling equal numbers in the last columns and rows) are positive. For the matrix AN under consideration its principal submatrices coincide with the same A M of smaller dimensions: M = 1, . . . , N . It now becomes clear that, since all principal determinants of the matrix AN are positive, this matrix is positive definite by virtue of the Sylvester criterion. Further, since the sum of positive definite matrices is always again positive definite, we can conclude that DN = AN + BN is positive definite, so ν– (DN ) = 0, ν+ (DN ) + N and therefore in Eq. (6.65) δN = sign(DN ) = ν+ (DN ) – ν– (DN ) = N . 4◦ . It should be noted that the matrix AN is related to the problem with all ρm = ∞, m = 1, . . . , N , i.e., to the case when each reflector is locally a plane mirror. With such a system of plane reflectors, by substituting equality (6.68) into Eq. (6.65) with D N = AN , δN = N , we arrive at the following expression for the amplitude of N times re-reflected acoustic wave: N s N P P Lm , (6.71) p(x) = exp ik Lm m=0

m=0

which is in complete agreement with clear optical ideas that presence of plane reflectors do not modify the value of the incident amplitude.

6.5. Application: Examples of High-Frequency Multiple Diffraction As an example of the application of the theory let us consider symmetric back far-field reflection from a couple of equal acoustically soft circles of radius R with a distance between their centers of 3R (see Fig. 6.4). There are two rays of the usual reflection with N = 1 at point A and B , and two rays with N = 2 (CDEF and backwards). Due to ρn = R > 0, σN = N , N = 1, 2, the full reflection coefficient predicted by the ray theory is √ √ √ √ exp[(5 – 2 2)kRi] , L = | DE | = 3 – 2 R. p = – πkR + 2πk – q 1 √ 2 L1 1/L1 + 2 2/R – 1/L21 (6.72) Here, and in all the examples considered below, the reflected pressure is taken with such a scale √ that for back scattering by a disk of radius R the reflection coefficient is equal to πkR/2. Figure 6.4 demonstrates the dependence of the reflected pressure upon the parameter R/λ (λ = 2π/k is the wavelength). The exact solution was calculated using the boundary element technique. For all further examples the obstacle boundary was chosen as acoustically soft because of the convenience for exact calculations. Symmetric reflection from a semicircle. The problem of far-field symmetric back reflection from the concave side of a semicircle is very interesting, important, and not trivial, so it is considered separately. Figure 6.5 shows that for every N > 1 there are two rays (such as ABCDE and backward), for which L0 = LN = L/2 + ∞, Ln = L = 2R cos θN , n = 1, . . . , N – 1, ρn = –R, θN = 12 π (N – 1)/N . For N = 1, only the ray of direct mirror reflection exists. Hence in formula (6.65), det(DN ) = L–N det(GN ), with the entries of the matrix GN being: g11 = gN N = –3, gnn = –2, n = 2, . . . , N – 1, gn+1,n = gn,n+1 = 1, n = 1, . . . , N – 1, with the others gnm = 0. It can be proved by induction that det(GN ) = (–1)N 4N . The most difficult task is to determine


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|p| 14 C

12

F

A DE B R R R

10 8 6 4

1

3

2

R/l

4

Figure 6.4. Reflection from a couple of acoustically soft disks: solid line represents the exact solution and dashed line represents the ray solution

|p| A

E qN

30

B

D C

20

10

40

42

46

44

48

2R/l

Figure 6.5. Symmetric reflection from an acoustically soft semicircle: solid line represents the exact solution and dashed line represents the ray solution the value of δN . For this purpose localization of eigenvalues for the matrix G N can be achieved using the well-known Gerˇsgorin disks. It is shown that the full set of eigenvalues lies in the complex plane inside of a union of the disks |z – gnn |
0,

(y1 + y2 )2 + 4p2 (y1 – y2 ).

(6.79) (6.80)

It can be shown that every pair y1 , y2 for which y1 , y2 = –p2 gives a stationary value for ϕ, this proves the nondisconnection of the stationary point. To evaluate an asymptotic behavior of the integral (6.79), let us introduce the new variables: p z 2 = y1 + y2 , (6.81) z1 = (y1 + y2 )2 + 4p2 – (y1 – y2 ), with the Jacobian J ≡ –1/2. Then

4p ϕ = 12p2 – z12 .

(6.82)

Detailed consideration gives the result that the two-dimensional region of integration in (6.79) is such that r Z p, p2 Z b(z2 ) 2 e–ikz1 /4p dz1 ik p 1 3kpi qp , (6.83) 2e dz2 A=– π 2 2 2 2 2 2 0 a(z2 ) ( p + z2 – z1 /2) + z1 /4 p √ a(z2 ) = z2 – 2 2 p + z22 + 4p2 ,

b(z2 ) = –z2 – 2 +

p z22 + 4p2 .

(6.84)

The value z1 = 0 with any z2 is stationary. This confirms once more that the stationary point is not isolate. First of all, the internal integral should be estimated separately. The one-dimensional method of stationary phase can be applied for this. As could be √ expected a priori, its stationary value z1 = 0 exists not for arbitrary z2 , but only if 0 < z2 < p 2. This means that the arcs l mentioned above are located on p/4 < x 1 , x2 < p/2. Estimation of the integral in (6.83) gives r Z πp e–(π/4)i dz1 –ikz12 /4p q p p ∼ 2 . (6.85) e 2 2 + z2 k 4 p 2 2 2 2 4p + z2 – z1 /2 + z1/4 © 2005 by CRC Press LLC

Page 183

Therefore, A=–

r

2k (π/4)i 3kpi pe e π

Z

p, p2 0

dz2

p =– z22 + 4p2

r

k p e(π/4)i ln 2e3kpi . 2π

(6.86)

In order to write out the final result, the contribution of the ray with N = 1 should be recalled. Besides, the expression (6.86) should be taken twice because for every ray considered above, an inverse ray exists too. Taking into account a special scale, the final result can be represented as follows: p (6.87) A = –kp ln 2ei(3kp+π/4) + i πkp/2 e2kpi . It should be noted that the first power of k appears for the first time in this paper: compare, for instance, with (6.75). Thus the parabolic reflector is the most powerful among all known reflectors. Comparison between the exact solution and the solution given by (6.87) is shown in Fig. 6.6(b). The leading term (of the order of O (kp)) in (6.87) would give a straight line in Fig. 6.6(b) that passes through the average values of the oscillating graph.

Helpful remarks It is interesting to note that two very similar geometries represented respectively in Figs. 6.5 and 6.6 yield absolutely different reflected amplitude. In the case of semi-circle the amplitude is quasi-constant, and the calculation of this quantity requires summation of a slowly convergent series. In the case of parabolic reflector the reflected amplitude grows as the first power of the frequency, and this amplitude is expressed by a very simple formula. The marked different qualitative and quantitative reflecting properties of so close in shape reflectors demonstrates that high-frequency analysis should be very refined to estimate correctly the reflected amplitude.

6.6. Application: Physical Diffraction Theory for Nonconvex Obstacles The ideas discussed in the previous section can be applied to extend Kirchhoff’s physical diffraction theory to the case of nonconvex obstacles, where multiple ray re-reflections may take place. As we remember from Chapter 2, the diffraction problem in the generic case can be completely studied and numerically solved for arbitrarily shaped obstacle by the boundary element methods. Here we can mark the cases when the BIE method at high frequencies, within the framework of the physical diffraction theory, can be reduced to calculation of some integrals. As we will see below, generally this is expressed by some multiple integrals, and in a certain sense such an approach is indeed a generalization of the classical physical diffraction theory (see Section 2.6). For simplicity, let us restrict our consideration by the scalar problem only and only by the case of a back scattering, when the point source and the receiver coincide with each other. Let us study the diffraction problem for an acoustically hard obstacle, whose boundary contour l is an arbitrary smooth closed contour of finite length. Let us first emit an extremely dense set of irradiated rays from the source x0 and denote by l + the part of the boundary contour l containing the points of first reflections for any emitted ray. The contour l + is thus an illuminated part of l. As follows from the Kirchhoff’s theory, boundary value of acoustic pressure vanishes in the shadow, i.e., outside l + : p(y ) = 0, y ∈ l \ l + . Now, according to the


Page 184

Kirchhoff–Helmholtz integral formula (see Chapter 2), exact value of the diffracted wave amplitude is Z ∂Φ sc dly , (6.88) p (x0 ) = p(y ) ∂n y l+

where the basic question is: how the value of the boundary pressure over illuminated part l + can be determined in Eq. (6.88). The considered illuminated part l + may contain both convex and nonconvex zones. Let us denote again as l1+ and l2+ , respectively, the convex and nonconvex parts of the boundary, on the illuminated side: l + = l1+ ∪ l2+ . As indicated above, the solution of the basic BIE on l1+ is calculated in the same way as predicted by classical Kirchhoff’s theory, because no re-reflected ray can attain these parts: eik|x0 –y| . p(y )|l+1 = 2p0 = 2 √ |x0 – y |

(6.89)

If the considered contour l would be convex, then this value (6.89) could be directly substituted as a known boundary value of acoustic pressure to the Kirchhoff–Helmholtz integral, to calculate wave amplitude at any observation point. However it is not so easy to calculate a priori the pressure on the existing nonconvex parts l 2+ of the boundary, since there are multiple re-reflections on these parts. Our approach to simplify this calculation is based on the following evident observation. The geometrical diffraction theory in the case of multiple re-reflections operates with asymptotic estimates of some multiple integrals by the stationary phase method. If we calculate these multiple integrals exactly, not by applying any asymptotic estimates, then we arrive at a certain analogue to Kirchhoff’s physical diffraction theory for nonconvex obstacles admitting re-reflections of the incident wave by their boundaries. Following this idea, we should first trace the paths of all rays that depart from and come back to the same point x0 , with the calculation of the maximum number N for any ray during its re-reflections over the arc l 2+ . If N = 1, then for every ray, only simple reflections without re-reflections can take place. In this case the Kirchhoff–Helmholtz integral has the same form as in the case of convex obstacle, that is, it is given by a single integral over illuminated side of the boundary, since we have p(y ) = 2p0 , both on l1+ and l2+ , all over the light zone: Z 2ik|x0 –y| Z e ∂Φ 3πi/4 k sc dly = –e cos(ny , x0 – y ) dly . (6.90) p (x0 ) = 2p0 ∂ny 2π l+ |x0 – y | l+

For N > 1 we introduce the characteristic function χ(y 1 , y2 ), y1 , y2 ∈ l+ by the following definition: χ(y1 , y2 ) = 1 if the piece y1 –y2 does not intersect the contour (i.e., these two points are mutually visible), and χ(y1 , y2 ) = 0 in the opposite case. Suppose, for example, N = 2. This means that every ray irradiated from the point x 0 and returned back to this point does reflect from the obstacle boundary l no more than twice its during full flight. Hence, before its last hit at the boundary, which indeed defines the boundary value of the function p(y ) in Eq. (6.88), each ray did undergo no more than only one reflection. So, when calculating p(y ), y ∈ l + , we may add to the result of a direct hit 2p0 also a single Kirchhoff–Helmholtz integral that is responsible for the influence of one-time repeated reflections: Z ∂Φ χ(y , y1 ) dl1 , y ∈ l+ . (6.91) p(y ) = 2p0 (y ) + 2p0 (y1 ) ∂n + 1 l Here we have taken into account that the point y 1 can influence the boundary value of p(y ) at the point y only in the case when the latter is visible from y 1 . As a result, for N = 2 the


Page 185

physical diffraction theory is reduced to calculation of the double integral which is obtained by the substitution of relation (6.91) into Eq. (6.88). It is clear that for arbitrary N ≥ 3 the physical diffraction theory is reduced to a certain N -fold integral. The corresponding formulas are omitted here for the sake of brevity. If for a certain contour the quantity N is known approximately from any a priori understanding, then it may be taken on the safe side, i.e., a little higher by a few units. Correctness of this statement follows from the evident observation: if contribution of any re-reflected wave to the boundary value p(y ), y ∈ l + is absent, then the corresponding multiple integral will be asymptotically small compared to the result of direct hit. It should be noted however that for N ≥ 3 the efficiency of the thus generalized Kirchhoff’s theory decreases considerably because the number of required arithmetic operations grows as M N , where M is the number of nodes in the corresponding quadrature formulas when being implemented on a computer. At the same time direct numerical treatment of the BIE by a collocation technique (see Section 1.5) requires around M 3 arithmetic operations since this is typically reduced to an M × M linear algebraic system (some helpful simple and clear ideas on numerical treatment of linear algebraic systems, with a FORTRAN code, can be found in Forsythe et al., 1977).

Helpful remarks It should be noted that, as in geometrical diffraction theory, for acoustically soft boundary the final result differs from that for acoustically hard obstacle by a factor (–1) N , which goes in front of the corresponding formula for the receiving acoustic pressure p(x 0 ). This leads to a very important physical conclusion. As shown in Section 2.6, for convex obstacles, when there are present only single integrals with N = 1, classical Kirchhoff’s theory predicts the real amplitude of acoustic pressure in the reflected wave to be identical for acoustically hard and soft boundaries, because the complex amplitudes differ only by their (opposite) signs. By contrast, for nonconvex obstacles with N > 1 we face with the summation of contributions of rays reflected from the boundary at different times. For all that, the contribution of the rays with even N for both types of boundaries are equal to each other, but with odd N contribute in anti-phase. Therefore, the result of summation of such contributions in the generic case will be different. This means that two geometrically identical objects, one of them being acoustically hard and another acoustically soft, even at extremely short-wave regime may possess absolutely different back-scattered patterns. Nevertheless, there can be indicated relatively wide class of nonconvex obstacles, for which re-reflections are impossible, and so for them we may always put N = 1. For obstacles of this class, reflection from acoustically hard and acoustically soft boundaries will be the same in absolute value. It is proved in the author’s work (Druzhinina and Sumbatyan, 1992) that if on a concave arc of the boundary the angle between two arbitrary normals is less than the right angle, then ray re-reflections are surely impossible on this part of the boundary contour. We thus can extract a class of nonconvex domains with weakly concave boundaries. This class is covered by obstacles with the boundaries, for which on any concave part the angle between an arbitrary pair of normals is less than 90 ◦ . For the obstacles of this class the physical diffraction theory, like for convex domains, is described by a usual single Kirchhoff–Helmholtz integral.

6.7. Short-Wave Integral Operator in Diffraction by a Flaw in Elastic Medium This problem is of great interest as regards its application to ultrasonic NDT. The general approach developed above is suitable for the elastic case. Let p and s refer to the longitudinal (pressure) and transverse (shear) wave, respectively.


Page 186

At the beginning, consider p–p reflection. Let an oscillating unit force be located at the point x0 , and its projection into the x0 y direction be Q, y ∈ l. The value of Q depends upon the mutual disposition of x0 and y . For instance, if the initial unit force is perpendicular to the x0 y segment, then Q = 0. inc It can be shown that the longitudinal component of Cartesian displacements u inc 1 , u2 caused by the force Q and related to the tangent and the normal at the point y are defined as eikp |x0 –y| , uinc 1 (y ) = –Q sin θ √ |x0 – y |

(6.92)

eikp |x0 –y| √ . (6.93) ( y ) = – Q cos θ uinc 2 |x0 – y | The short-wave representation for the reflected longitudinal component of displacement is given by r 2 ikp |x–y| Z kp2 ks e 2kp –(π/4)i 2 dsy , e – sin 2θu1 (y ) + – 2 sin θ u2 (y ) √ ur (x) = 2 2 4ks π kp |x – y | l uψ (x) = 0, (6.94) where (r , ψ ) are polar coordinates related to the point y and (u 1 , u2 ) is the vector of full displacements over the contour l. In order to use Kirchhoff theory, those values for u1 (y ), u2 (y ) should be chosen in (6.94), which are defined by a solution of the diffraction problem for reflection of a plane wave from a half-plane cavity. As is well known, full displacements on the boundary are defined by the incident wave (6.92) and (6.93) and are equal to s ! kp2 ks (6.95) 1 – 2 sin2 θ Vp s uinc u1 (y ) = 1 + Vpp – 1 (y ), kp sin θ ks u2 (y ) = (1 – Vpp – tan θVps )uinc 2 (y ),

(6.96)

where Vpp and Vps are reflection coefficients for p–p and p–s reflection, respectively. After substituting (6.95), (6.96) and (6.92), (6.93) into (6.94), the expression in brackets of (6.94) becomes equal to s ( " # kp2 kp2 ks – sin 2θ – sin θ (1 + Vpp ) + 1 – 2 sin2 θ Vp s 2ks2 kp ks (6.97) ) 2 ks 2 – 2 sin θ [– cos θ (1 – Vpp ) + sin θVps ] = cos θVpp , + kp2 and here we give a proof for this identity. We start from Snell’s law (see Brekhovskikh, 1980) ks sin θ1 = kp sin θ

∼

sin θ ks = , kp sin θ1

(6.98)

where θ1 is the reflection angle of the transverse when θ is an angle between the incident longitudinal wave and the normal. We should also recall that (see Brekhovskikh, 1980) Vpp =

4 cot θ cot θ1 – (1 – cot2 θ1 )2 , 4 cot θ cot θ1 + (1 – cot2 θ1 )2

Vps =

4 cot θ (1 – cot2 θ1 ) . 4 cot θ cot θ1 + (1 – cot2 θ1 )2

(6.99)


Page 187

Then expression in the braces in Eq. (6.97) is ks2 ks 2 2 sin θ cos θ (1 + Vpp) – 2 sin θ cos θ cos θ1 Vps + 2 sin θ – 2 cos θ (1 – Vpp) kp kp 2 2 kp ks2 ks2 sin θ kp2 ks 2 2 – 2 sin θ sin θV = cos θ V + 4 sin θ – + + ps pp kp2 kp2 ks2 kp2 cos θ ks2 sin2 θ k2 sin θ (6.100) cos θ cos θ1 + 2 – 2 sin2 θ Vps = s2 cos θ Vpp × –2 sin θ1 sin θ1 kp sin θ sin2 θ sin2 θ1 2 2 –2 cos θ cos θ1 + 2 – 2 sin θ Vps + (4 sin θ1 – 1) + tan θ sin2 θ sin θ1 sin θ1 2 k = s2 cos θ Vpp + (4 sin2 θ1 – 1) + [–2 sin θ1 cos θ1 + (1 – 2 sin2 θ1) tan θ ]Vps . kp 2

Now, by the substitution of relation (6.99) as Vps into (6.100) the braces in (6.100) can be rewritten in the following form: Vpp +

sin2 θ1 4 cot θ cot θ1 + 1 – cot2 θ1

2 2 2 2 (3 – cot θ1 ) 4 cot θ cot θ1 + (1 – cot θ1 )

– 2 cot θ1 + (1 – cot2 θ1 ) tan θ 4 cot θ (1 – cot2 θ1 ) 2 4 cot θ cot θ1 – 1 – cot2 θ1 = Vpp + 2 = Vpp + Vpp = 2Vpp . 4 cot θ cot θ1 + 1 – cot2 θ1

(6.101)

It becomes finally clear that expression (6.100) is equal to 2(k s2 /kp2 )Vpp cos θ , and so expression (6.97) is kp2 ks2 2 Vpp cos θ = Vpp cos θ , (6.102) 2ks2 kp2 which was to be proved. Therefore, formula (6.94) for ur reads r Z kp cos θ –(π/4)i √ Vpp eikr (|x0 –y|+|x–y|) ds . ur (x) = Qe 2π L0 L l

(6.103)

Asymptotic estimate of the last integral is the same as for (6.97), and gives the final result QVpp exp{i[kp (L0 + L) + (π/4)(δ – 1)]} p . ur (x) = √ L0 L 1/Lo + 1/L + 2/ρ cos θ

(6.104)

As can be shown by analogy, for diffraction with an arbitrary number of reflection points p–p– · · · –p–p the general result coincides with a corresponding expressions for the scalar case multiplied by reflection coefficients at every boundary point. Deduction of a formula for other cases is more complicated. For one-time p–s reflection one can derive instead of (6.103) the following expression for a shear component of displacement: r Z ks cos γ –(π/4)i ks √ Vp s ei(kp |x0 –y|+ks |x–y|) ds , ur (x) = 0, (6.105) uψ (x) = Qe kp 2π L0 L l


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in polar coordinates, where γ is a reflection angle of the shear wave. This leads to the following result: r cos γ exp{i[kp L0 + ks L + (π/4)(δ – 1)]} ks ks p , uψ (x) = QVp s kp L0 L kp cos2 θ/L0 + ks cos2 γ/L + kp cos θ/ρ + ks cos γ/ρ kp cos2 θ ks cos2 γ kp cos θ ks cos γ + + + . δ = sign L0 L ρ ρ (6.106) Further, every type of multiple reflection can be described by a corresponding formula. For example, p–s –p double reflection is given by ur (x3 ) = Q

s

ks Vp s (θ1 )Vsp θ2 cos γ1 cos θ2 kp

exp i[kp L0 + ks L1 + kp L2 + (π/4)(δ2 – 2)] √ √ , × L0 L1 L2 | det(D2 )| where

(6.107) δ2 = sign D2 ,

D2 = (d2 )ij , i, j = 1, 2; cos2 θ1 ks cos2 γ1 cos θ1 ks cos γ1 + + + , (d2 )11 = L0 kp L1 ρ1 kp ρ1 cos γ1 cos γ2 cos γ1 cos γ2 , (d2 )21 = , (d2 )12 = L1 L1 ks cos2 γ2 cos2 θ2 ks cos γ2 cos θ2 + + + . (d2 )22 = kp L1 L2 kp ρ2 ρ2

(6.108)

Here, θ1 and γ1 are incident and reflected angles at the first point of reflection and γ 2 and θ2 are incident and reflected angles at the second point.

Helpful remarks Obviously, due to presence of reflection coefficient the reflected wave field from the obstacle in a scalar acoustic and elastic media are absolutely different. This is demonstrated by the following example. Let us consider symmetric √ reflection of a plane longitudinal wave from a concave side of a cosine-type curve: y = 2 a[1 – cos(x/a)], √|x| ≤ (π/3)a. It can be shown rays that besides the single direct reflection at x = 0 with ρ = –a 2, two mutually inverse √ p–p–p with N = 2 exist for which θ1 = θ2 = π/4 (at x = ±π/4), ρ1 = ρ2 = –a 8, δ2 = 0. Other types of reflection are absent. Thus the ray asymptotics has the following form: A=

p

πkp ai

! 2 √ Vpp (π/4) ikp a 1 e (π/2) + 2 – 2 2 –p . 321/4 1 – π/4

(6.109)

It was taken into account here that Vpp (θ = 0) = –1. Comparison of this ray solution with an exact one and with a scalar case is shown in Fig. 6.7 for an elastic aluminum medium. Note that Vpp (π/4) = –0.567 here. Scalar asymptotics would differ from (6.109) by the absence of the reflection coefficient Vpp . As can be seen, this leads to a totally different numerical result compared with the elastic case.


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30 |A|

20 3 2

10

1 0 10

akp /p 11

12

13

14

15

Figure 6.7. Symmetric reflection from a cosine-type crack in elastic medium: 1, exact solution; 2, ray solution; 3, prediction by scalar theory

6.8. High-Frequency Asymptotics of Integral Operator in a Three-Dimensional Diffraction Theory Here we extend some results of the previous sections, obtained in the 2D case, to the threedimensional theory. We consider diffraction of a high-frequency (i.e., short-wave) acoustic wave, generated by a point source, from arbitrary acoustically hard smooth obstacle Γ, which provides the boundary condition ∂p/∂n Γ = 0. As in the previous sections, our approach is based on an asymptotic estimate of Kirchhoff’s diffraction integral by a multiple stationary phase method. This allows us to obtain the amplitude of the reflected high-frequency wave in an explicit form. A similar result can be obtained starting from Keller’s geometric diffraction approach (Keller, 1962), and this can be found in more detail, for example, in McNamara et al. (1990). Let a point source x0 generate a spherical acoustic wave, which is incident onto a surface Γ. During interaction of the incident wave with the surface Γ on its convex parts there are only points of simple specular (mirror) reflection. Reflection process from concave parts of the surface is more complex; this is connected with possible re-reflections of acoustic rays (compare with Section 6.4), and is not considered here. The pressure in reflected wave, in the high-frequency limit, is defined by the direction of incident wave and by a small neighborhood of the point of specular reflection. Therefore, with the frequency increasing, the amplitude of the reflected wave can be obtained by a multiple stationary phase method. If the ray x0 – y – x is reflected from the surface only once, then the Kirchhoff–Helmholtz integral formula reads ZZ eikr ∂Φ(kr ) sc dsy , Φ(kr ) = , r = |x – y |, x ∈ R3 \ D , (6.110) p (x) = 2pinc (y ) ∂n 4 πr y Γ+ where we have taken into account that the leading asymptotic term of the pressure p(y ), y ∈ Γ+ is given by Kirchhoff’s physical diffraction theory as p(y ) = 2p inc (y ), y ∈ Γ+ (cf. Section 2.6). As usual, here pinc (y ) is the value of pressure in the incident wave on the boundary surface Γ, Φ(kr ) is the Green’s function, n y is the outer normal to the surface Γ at the point y of a small neighborhood of the specular reflection point y 0 . Note that asymptotically at k → ∞ we may put pinc (y ) =

eikr0 , r0

r0 = |x0 – y |,

∂Φ(kr ) eikr = ik cos γ 1 + O (k –1 ) , ∂ny 4πr

(6.111)


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x0 ny0

x

L0

L

y0

y

Figure 6.8. Reflection of acoustic wave from a smooth surface where γ is the angle between the normal ny and the direction of propagation of the incident ray r0 = x0 – y : γ = nd y r0 . Since the incident and reflected rays lie on the same plane with the normal to the surface at the reflecting point ny0 , formulas (6.110)–(6.111) lead to the following basic integral representation (after taking nonoscillating terms outside of the integral, in the considered short-wave approximation: ZZ ik cos γ sc eikS(y) dsy , S (y ) = r0 + r = |x0 – y | + |x – y |, p (x) = 4πL0 L (6.112) Γ+ L0 = |x0 – y0 |, L = |x – y0 |. We will obtain the leading (geometrical diffraction) asymptotic term by applying a multiple stationary phase method (see Fedorjuk, 1962) to the integral (6.112). The estimate procedure should take into account all points y from a small neighborhood of the point y0 . Let us link this small neighborhood to a right Cartesian coordinate system, which is defined by the normal ny0 and two lines of principal curvatures of the surface at the point y0 . Then y from the neighborhood of y 0 has the an arbitrary point following coordinates: y = ∆s1 , ∆s2 , – 21 k1 (∆s1 )2 + k2 (∆s2 )2 , where ∆s1 and ∆s2 are arc differentials along the lines of principal curvatures, k 1 = R1–1 and k2 = R2–1 are the principal curvatures, R1 and R2 are the principal curvature radii of the surface S at the point y0 , and k1 (∆s1 )2 + k2 (∆s2 )2 is the second quadratic form of the surface at the point y0 related to the principal curvature lines (see, for example, Pogorelov, 1961). Let us apply cosine theorem to the triangles x0 y0 y and xy0 y (see Fig. 6.4): |x0 – y |2 = L20 + |∆s|2 – 2L0 |∆s| cos(x0 y0 , y0 y ),

(6.113) |x – y |2 = L20 + |∆s|2 – 2L|∆s| cos(xy0 , y0 y ), where vector ∆s is: ∆s = ∆s1 , ∆s2 , – 21 k1 (∆s1 )2 + k2 (∆s2 )2 . Let us denote the unit vector along direction (y0 x0 ) as {cos α, cos β , cos γ }, then {– cos α, – cos β , cos γ } is the unit vector along direction (y 0 x). Now, by taking the scalar products of these vectors with the vector ∆s, we can obtain |∆s| cos(x0 y0 , y0 y ) = (∆s1 ) cos α + (∆s2 ) cos β – 21 k1 (∆s1 )2 + k2 (∆s2 )2 cos γ , |∆s| cos(xy0 , y0 y ) = –(∆s1 ) cos α – (∆s2 ) cos β – 21 k1 (∆s1 )2 + k2 (∆s2 )2 cos γ ,

(6.114)


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hence

2 2 |x0 – y | = L0 – (∆s1 ) cos α – (∆s2 ) cos β + 21 L–1 0 sin α + k1 cos γ (∆s1 ) 2 2 –1 3 + 12 L–1 0 sin β + k2 cos γ (∆s2 ) – L0 cos α cos β (∆s1 )(∆s2 ) + O (|∆s| ), (6.115) |x – y | = L + (∆s1 ) cos α + (∆s2 ) cos β + 21 L–1 sin2 α + k1 cos γ (∆s1 )2 + 12 L–1 sin2 β + k2 cos γ (∆s2 )2 – L–1 cos α cos β (∆s1 )(∆s2 ) + O (|∆s|3 ).

From these relations we can extract the behavior of the phase function S (y ) defined in Eq. (6.112), in the following form: S (y ) = L0 + L + 21 d11 (∆s1 )2 + d12 (∆s1 )(∆s2 ) + 21 d22 (∆s2 )2 + O (|∆s|3 ), 2 2 –1 –1 d11 = L–1 sin α + 2k1 cos γ , d22 = L–1 sin β + 2k2 cos γ , 0 +L 0 +L –1 –1 d12 = – L0 + L cos α cos β .

(6.116)

Absence of the first powers (∆s1 ) and (∆s2 ) here confirms that the point y0 of a (mirror) specular ray reflection provides a stationary value for the phase function S (y ). Thus, the multiple stationary phase method (see Section 1.4) implies that the leading asymptotic term of the integral (6.112) is determined by coefficients in front of (∆s 1 )2 , (∆s1 )(∆s2 ), (∆s2 )2 . This finally yields exp i k (L0 + L) + (π/4)(δ2 – 2) √ , (6.117) p(x) = L0 L |det(D2 )| where D2 is the Hessian of a symmetric structure: D2 = {dmj }, dmj = djm (m, j = 1, 2), and δ2 = sign(D2 ) = ν+ – ν– is the difference between positive and negative eigenvalues of the matrix D2 . Taking into account that d21 = d12 , the final result can be obtained in the following form: exp i k (L0 + L) + (π/4)(δ2 – 2) . (6.118) p(x) = 1/2 |(L0 + L)2 + 2L0 L(L0 + L) (k2 sin2 α + k1 sin2 β ) (cos γ )–1 + 4L20 L2 K|

Here K = k1 k2 is the Gaussian curvature of the surface at the point y 0 . Note also that the vector {– cos α, – cos β , – cos γ } determines direction of incidence of the ray x 0 y0 in the chosen coordinate system.

Helpful remarks 1◦ . Certainly, in the cases when expression in the denominator of Eq. (6.118) is trivial, our result (6.118) is not valid. In this case we are faced with an irregular stationary point, and the dependence of the reflected amplitude upon the frequency is more complex. 2◦ . As in the 2D problem, formula (6.118) was derived for the case when a high-frequency wave is incident on a convex surface. If the wave is incident on a concave surface, then the principal curvatures k1 and k2 should be taken negative. 3◦ . Let us stress two limiting cases when formula (6.118) can be considerably simplified. First of all, in the case when the reflecting surface is locally plane, we have k 1 = k2 = 0. It can be easily seen that in this case expression (6.118) reduces to p(x) =

eik(L0 +L) , L0 + L

(6.119)

which conforms to clear heuristic ideas.


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Another simple case is related to a far-field scattering. In this case L 0 , L R1 , R2 , and the last term in the denominator predominates over two other ones. Then formula (6.118) reduces to √ R1 R2 exp i k (L0 + L) + (π/4)(δ2 – 2) , (6.120) p(x) = 2L0 L which coincides with Shenderov (1972). As you can see, the amplitude of the far-field scattered wave is proportional to the √ square√root of the quantity inverse to the Gaussian √ curvature: p(x) ∼ R1 R2 ∼ 1/ K = 1/ k1 k2 .


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Chapter 7

Inverse Problems of the Short-Wave Diffraction 7.1. Some Basic Results in a Local Differential Geometry of Smooth Convex Surfaces Since almost the whole of this chapter will be devoted to the inverse (reconstruction) problem for smooth convex obstacles, we first give a short survey of some classical results from differential geometry devoted to the concept of the Minkowski support function, which is essentially used in our study. Let us consider a smooth convex closed surface S . We introduce the function P (q) related to the distance p(q) from the origin O to the tangential plane with the normal q, as follows (see Blaschke, 1930; Pogorelov, 1973, 1978; Ramm, 1986): P (α1 , α2 , α3 ) = r p(α1 /r , α2 /r , α3 /r ),

n = (α1 , α2 , α3 ),

r = (α21 + α22 + α23 )1/2 , (7.1)

where n = –q is a unit normal vector to the surface. Note that the Cartesian coordinates of the boundary surface can be expressed in terms of the function P (α 1 , α2 , α3 ) as xi = ∂P/∂αi . First, for some mathematical transformations we consider the function P (α 1 , α2 , α3 ) as a function of three variables (α1 , α2 , α3 ), each of them varying from –∞ to +∞. And then, in final treatment we should take into account that the real surface corresponds to the value r = 1. The introduced function P (α1 , α2 , α3 ) is called Minkowski support function. Let us derive a partial differential equation that gives the relation between the support function and the corresponding Gaussian curvature. The function P (α1 , α2 , α3 ) is a homogeneous function of the first degree, with respect to its arguments: P (µα1 , µα2 , µα3 ) = µP (α1 , α2 , α3 ),

µ > 0.

(7.2)

According to Euler theorem, for the homogeneous function P the following identity holds: α1 P1 + α 2 P2 + α 3 P3 = P ,

Pi =

∂P . ∂αi

(7.3)

The tangential plane to the surface can be written in the form α1 x1 + α 2 x2 + α 3 x3 = P ,

(7.4)

which is a classical result of analytic geometry. Let us apply to Eq. (7.4) the partial derivatives with respect to αi , i = 1, 2, 3. Then one comes to Cartesian coordinates of the points of tangency in terms of the derivatives Pi : xi = Pi (α1 , α2 , α3 ).

(7.5)


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Let us express principal curvatures of the surface in terms of the support function, following to Rodrigues relations (see, for example, Pogorelov, 1961). On the surface, along the curvature lines, these formulas take the form dxi + R dξi = 0.

(7.6)

It follows from (7.5) that dxi =

3 X

Pik dαk ,

Pik =

k=1

∂2P . ∂αi ∂αk

(7.7)

For real surface, when r = 1, the quantities αi become the coordinates of the unit normal ξi : αi = ξi . Then the substitution of (7.7) into (7.6) yields the following identity: 3 X

Pik dξk + R dξi = 0

(i = 1, 2, 3).

(7.8)

k=1

When moving along the curvature line all three dξ i cannot vanish simultaneously. Therefore, determinant of system (7.8) must be equal to zero: det(P ik + δik R) = 0, i, k = 1, 2, 3, where δik is Kronecker’s delta. The last determinant is R3 + (P11 + P22 + P33 ) R2 + (P11 P22 + P11 P33 + P22 P33

– P122 – P132 – P232 ) R + det(Pik ) = 0

(i, k = 1, 2, 3).

(7.9)

Let us prove that determinant in (7.9) is equal to zero. To this end, we apply to Eq. (7.3) differentiation with respect to αi , i = 1, 2, 3. As a result we arrive at the homogeneous P3 system k=1 Pik αk = 0, i = 1, 2, 3, where all three αi cannot be equal to zero simultaneously, so determinant of the system is trivial: det(Pik ) = 0. Hence, equation (7.9) is a quadratic equation, whose roots R 1 and R2 are principal curvature radii satisfying the following identities: P11 + P22 + P33 = –(R1 + R2 ),

(7.10)

P11 P22 + P11 P33 + P22 P33 – P122 – P132 – P232 = R1 R2 = K –1 (q),

(7.11)

according to the Vieta theorem (K is the Gaussian curvature). It is clear from Eqs. (7.10) and (7.11) that reconstruction of the surface S , through the Minkowski support function P , is possible in two cases: 1) if we know the average curvature, i.e., the function R1 + R2 is known; 2) if we know the Gaussian curvature K of the surface, i.e., the function R1 R2 is known. Below in this section we study in more detail the first problem, which is covered by the Christoffel theorem: THEOREM 7.1 (CHRISTOFFEL). If there is a one-to-one correspondence between the smooth convex closed surface S and a unit sphere arranged by the endpoints of outer unit normals to this surface, all beginning from the same fixed point of the 3D space, then the function R1 + R2 defines the surface uniquely. Proof. The proof is constructive and follows Blaschke (1930) and Pogorelov (1973). Let us introduce the spherical coordinate system   α1 = r sin θ cos ϕ, α2 = r sin θ sin ϕ, (7.12)  α3 = r cos θ . © 2005 by CRC Press LLC

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Further treatment is based upon some classical results on series in spherical functions (see, for example, Hobson, 1955; Tikhonov and Samarsky, 1977). The spherical functions Um (α1 , α2 , α3 ), m = 0, 1, 2, . . . , are homogeneous polynomials of the power m, being simple solutions of the Laplace equation (Um )11 + (Um )22 + (Um )33 = 0. In the spherical coordinate system these functions can be expressed as (j) sin(jϕ) m Um (α1 , α2 , α3 ) = r Pm (cos θ ) (j = 0, 1, . . . , m), (7.13) cos(jϕ)

(j) where the Pm (x) = (1 – x2 )j/2 dj Pm (x)/dxj are associated Legendre functions, and the Pm (x) are Legendre polynomials. It is proved that the system of spherical functions, considered as functions of (θ , ϕ), is complete and orthogonal on the spherical surface 0 ≤ θ ≤ π , 0 ≤ ϕ ≤ 2π . This predetermines the most natural and sufficient way to solve the considered Christoffel problem. Let us expand the support function, defined initially on a unit sphere, to a series in spherical functions:

P (α1 , α2 , α3 ) =

∞ X

r = (α21 + α22 + α23 )1/2 = 1,

cm Um (α1 , α2 , α3 ),

(7.14)

m=0

where cm are some unknown coefficients. Then, by taking into account that P (α 1 , α2 , α3 ) is a homogeneous function of degree 1, and Um (α1 , α2 , α3 ) is a homogeneous function of the power m, we can conclude that representation of the support function in the full 3D space is P (α1 , α2 , α3 ) =

∞ X m=0

cm

Um (α1 , α2 , α3 ) , r m–1

0 < r = (α21 + α22 + α23 )1/2 < ∞.

From this representation we can calculate ∆P = P11 + P22 + P33 : ( ) 3 ∞ X X ∂ 1 ∂Um 1 ∆Um +2 + Um ∆ . ∆P = cm r m–1 ∂αi r m–1 ∂αi r m–1 i=1 m=0

(7.15)

(7.16)

In order to calculate this expression, let us recall that ∆U m = 0, due to definition of spherical functions. Besides, since these functions are homogeneous with the power m, we have 3 3 X 1 ∂Um m – 1 X ∂Um m–1 ∂ = – m+1 = – m+1 mUm . αi m–1 ∂α r ∂α r ∂α r i i i i=1 i=1

At last, (m – 1)(m – 2) 1 = , ∆ m–1 r r m+1

∞ X

(7.17)

(m – 1)(m + 2) Um (α1 , α2 , α3 ). r m+1 m=0 (7.18) If we apply the last identity for r = 1, so as to consider the Christoffel equation (7.10) on the unit sphere, we arrive at the following representation: ∆P = –

∞ X

so ∆P = –

cm (m – 1)(m + 2) Um (α1 , α2 , α3 ),

cm

r = (α21 + α22 + α23 )1/2 = 1.

(7.19)

m=0


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Recall that the set of spherical functions form a complete and orthogonal system of functions. Therefore, let us expand the given function R 1 + R2 , as a function of the unit normal, i.e., as a function of the spherical coordinates (θ , ϕ) with r = 1, to a series in the spherical functions: ∞ X dm Um (α1 , α2 , α3 ), (7.20) R1 + R 2 = m=0

where we consider coefficients dm to be known. Then Eq. (7.10) involves (m – 1)(m + 2) cm = dm

∼

cm =

dm . (m – 1)(m + 2)

(7.21)

Therefore, the solution to the Christoffel problem (7.10) is given by the following equality: P =

∞ X m=0

dm Um . (m – 1)(m + 2)

(7.22)

You can notice that this formula defines a correct solution only if the coefficient with m = 1 is trivial: d1 = 0. Due to orthogonality of the spherical functions in relation (7.20) this is equivalent to Z (R1 + R2 ) U1 dΩ = 0, (7.23) r=1

where dΩ is an elementary space angle. Taking the concrete form of the spherical function U1 this in particular involves Z

r=1

(R1 + R2 ) αi dΩ = 0

∼

Z

(R1 + R2 ) n dΩ = 0.

(7.24)

r=1

The last condition is the necessary condition for the Christoffel theorem to be valid. Pogorelov (1973) proves that sufficient conditions are given by the set of the three relations, where the relation representing the necessary condition is only the first condition, among three other ones: Z 1) (R1 + R2 ) n dΩ = 0, 2) (R1 + R2 ) ≥ 0, 3) (R1 + R2 ) – (R1 + R2 )ss ≥ 0, (7.25) r=1

where differentiation is applied with respect to the arc length (s ) of any circle of a unit radius on the considered sphere.

Helpful remarks The method discussed in this section is not applicable for reconstruction of the convex obstacle shape from the known scattering diagram, since this problem is reduced to a Minkowski problem rather than a Christoffel problem. However this study demonstrates that the inverse problem for convex 3D convex surface, at least in the case of linear Christoffel formulation, can be efficiently resolved with the help of treatment in spherical coordinates.


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7.2. Reducing Inverse Problem of the Short-Wave Diffraction to Minkowski Problem As follows from Eq. (6.120), the far-field scattered amplitude in a direct diffraction problem for convex obstacles is expressed in terms of the Gaussian curvature. Therefore, in practical (inverse) reconstruction problems, when the scattered amplitude is usually known for a set of observation angles being received in a far zone, we arrive at the inverse problem which coincides with the classical Minkowski problem well known in differential geometry (see, for example, Blaschke, 1930; Pogorelov, 1973, 1978). This problem can be formulated in the following way. Let S be a sufficiently smooth closed convex surface with the given positive continuous Gaussian curvature known as a function of the outward normal: K = K (n). It is required to reconstruct the shape of the surface itself. Here we study the questions of existence and uniqueness. This problem was solved in part by H. Minkowski himself; then some important results were obtained by A.D. Aleksandrov and A.V. Pogorelov. Our study follows in the main Pogorelov (1973, 1978). All approaches are based on approximation of the given convex surface by a convex polyhedron. Let us first derive a necessary condition for the Minkowski problem to be solvable. We first start from the consideration of arbitrary convex polyhedron P N with N faces, with ni being the normal to the ith face and Si its area. Let e be an arbitrary fixed unit vector in the space, and we displace the polyhedron as a solid figure in direction of this unit vector, by the distance ε. Displacement of the ith plane involves the change of the polyhedron volume by Si (ni ⋅ e) ε. Since the total volume of the polyhedron does not change, we can conclude that the following equality must hold: N X

(ni ⋅ e) Si = 0.

(7.26)

i=1

Now, due to arbitrariness of the vector e, it follows from Eq. (7.26) the necessary condition of solvability, valid for arbitrary closed polyhedron N X

ni Si = 0.

(7.27)

i=1

Further, let us consider an arbitrary smooth closed convex surface S . We can arrange a sequence of convex polyhedra {PN }, N = 1, 2, . . . , which converges to S . Then for any polyhedron PN equality (7.27) is valid, so in the limit N → ∞ we arrive at the following condition, which must be valid for any closed convex surface S : Z n ds (n) = 0. (7.28) S

If K (n) designates Gaussian curvature of the surface at the point with normal n, then as known from differential geometry (see, for instance, Pogorelov, 1961), ds (n) = dω (n)/K (n), where dω is an elementary solid angle of space in direction n. Substitution of the last relation into Eq. (7.28) leads to the necessary condition, which the given function K (n) of our inverse problem must satisfy, in order to provide solvability of the inverse problem: Z n dω (n) = 0, (7.29) Ω K (n)

where Ω is the full solid angle of the space.


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We may pose quite naturally the question, whether this condition is also sufficient to guarantee the existence of a closed convex surface given by its Gaussian curvature. Unlike the Christoffel problem, the Minkowski problem cannot be solved by a constructive method, which permits any analytical representation for its solution. A real way to establish sufficient conditions is again to operate with approximation of the convex surface by inscribed polyhedra. The following results were stated by Alexandrov (see Alexandrov and Zalgaller, 1967). Let ni , i = 1, 2, . . . , N , be a system of noncoplanar unit vectors, and S i > 0, i = 1, 2, . . . , N , are some positive numbers, satisfying condition (7.27). Then there exists a unique closed convex polyhedron PN , whose faces have outward normals ni and respective areas Si . From this basic result of the theory of convex polyhedra Pogorelov (1973, 1978) proves the theorem, which declares: Suppose there is a unit sphere with center at the origin of a Cartesian coordinate system and suppose a continuous positive function K (n) > 0 determined for every point on the sphere with radius vector n is given. Let condition (7.29) be valid, where the integral is taken over this sphere. Then there exists a unique closed convex surface S , which for any point of the surface with the outward normal n has the Gaussian curvature K (n). The results of this section allow us, in principle, to uniquely reconstruct the shape of a convex closed obstacle from the data on the wave amplitude scattered by the present obstacle in a far field.

Helpful remarks 1◦ . The quoted classical results of differential geometry guarantee the existence and uniqueness of a solution to the Minkowski problem, but do not offer you any specific efficient way to construct such surface. Some concrete algorithms of the shape reconstruction of convex obstacle can be found below in the forthcoming sections. 2◦ . It should be noted that reconstruction of the convex obstacle shape, which is reduced in the high-frequency range to the Minkowski problem, allows us to come to some interesting and very important conclusions, related to uniqueness of the studied inverse problem. Let us assume that in the considered short-wave range there is known the back-scattered amplitude of some nonconvex obstacle as a function of the observation angle α: A(α). Obviously, this function is positive and continuous (some results on analytic properties of the scattered wave field can be found, for example, in Colton and Kress, 1983). As follows from Eq. (6.120), where we put L = L0 , the considered back-scattered far-field amplitude is √ eikL0 |pinc | R 1 R2 , pinc = = √ , (7.30) A(α) = 2 2L0 L0 2 K L0 which is evidently of the same dimension as the incident wave. Let us introduce the auxiliary function 2

K (α) =

|pinc | . 4 A2 (α) L20

(7.31)

The main result stated above in the present section declares that if condition (7.29) holds, then there is a unique convex obstacle whose boundary surface Gaussian curvature is given by Eq. (7.31). This implies a very important conclusion. Namely, let A(α) be a back-scattered far-field diagram for any (not necessarily convex) obstacle for a fixed high frequency. If Z A2 (α) n(α) dα = 0, (7.32) Ω


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then, in addition to the real obstacle generated by this diagram, there is always some convex obstacle with the scattered diagram A(α). An interesting conclusion follows from the last observation. If we consider any (nonconvex) obstacle symmetric with respect to the origin of the coordinate system, then it is obvious, due to natural symmetry, that relation (7.32) is always valid, because the contributions of the integrand with positive and negative values cancel each other out. The same fact takes place if the object is symmetric with respect to all Cartesian coordinate axes. So, in this case we can guarantee (at least at high frequencies) nonuniqueness of the solution of the considered inverse problem on reconstruction of the boundary surface of the obstacle in acoustic medium by its known back far-field scattering pattern.

7.3. Explicit Results for a Differential Operator of the 2D Inverse Problem Let us now study the two-dimensional problem. By analogy to the general case, the backscattered amplitude, in ray approximation, is proportional to |ρ(q)| 1/2 , where ρ is a radius of curvature at the point of specular reflection (see Chapter 6). The problem is again nonlinear, and it is wonderful that it can be reduced to a linear partial differential equation with respect to the Minkowski function. To derive that equation, let us start from the Rodrigues formula again, which can be written along the contour: dxi – ρ dαi = 0, i = 1, 2. Since xi = ∂P (α1 , α2 )/∂αi (xi = Pi ), P2 we have dxi = k=1 ∂ 2 P/∂αk ∂αi , so 2 X ∂2P dαk – ρ dαi = 0, (i = 1, 2). (7.33) ∂αk ∂αi k=1

The last homogeneous system has a nontrivial solution, so its determinant is equal to zero, with i , k = 1, 2: 2 ∂ P ∂2P ∂2P ∂2P – δik ρ = 0 ∼ ρ ρ – + = 0. (7.34) + det det ∂αk ∂αi ∂α21 ∂α22 ∂αk ∂αi It can be proved, by analogy with the general case, that a homogeneity of P involves the last determinant to vanish. Hence we obtain d2 p(θ ) ∂2P ∂2P + = ρ ∼ + p(θ ) = ρ(θ ), (7.35) ∂α21 ∂α22 dθ 2 where ρ(θ ) is a curvature radius of the obstacle boundary curve. The passage from the first to the second equation in formula (7.35) evidently follows the change of variables, from Cartesian to polar coordinates over the unit circle r = 1. It is well known that the solution of the inhomogeneous ordinary differential equation (7.35), for arbitrary function ρ(θ ), is p(θ ) = p0 (θ ) + p∗ (θ ), p0 (θ ) = C1 sin θ + C2 cos θ , Z θ Z θ (7.36) p∗ (θ ) = sin θ ρ(θ ) cos θ dθ – cos θ ρ(θ ) sin θ dθ , θ0

θ0

where p0 (θ ) is the general solution of the homogeneous equation, and p ∗ (θ ) is a particular solution of the inhomogeneous one. The latter is constructed by a variation method applied to the constants C1 and C2 . We have tested experimentally this theoretical method of reconstruction, which is suitable for 2D convex obstacles, on some artificially produced voids made in thick metallic rods. The results on measurements of the back-scattered far-field amplitude are presented in Fig. 7.1. As can be seen from this figure, measurements on the back-scattered diagram there are carried out with the angular step 30◦ .


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1.0

1.0

0.8

0.8

0.6

0.6

0.4

0.4

0.2

0.2

0°

90°

180°

270°

360°

0°

90°

180°

270°

360°

Figure 7.1. Back-scattered diagram for artificially made triangular and oval voids It should be noted that despite so poor experimental data (collected with the angular step 30◦ ) results of reconstructions show indeed that the proposed method is very stable. The results of reconstructions are demonstrated in Fig. 7.2, where real boundary is marked by a solid and the reconstructed contour by a dashed line.

Figure 7.2. Results of reconstruction of triangular and oval voids Another possible approach here is to directly solve the linear differential equation (7.35) by a finite-difference method, and our detailed analysis confirm efficiency of this method too. It is interesting to demonstrate our scheme on a simple obstacle, such as round disk. If the radius of the disk is equal to R, then in this case ρ(θ ) ≡ R, p(θ ) ≡ R if the origin of the coordinate system is taken in the center of the obstacle. Evidently, for these values of ρ(θ ) and p(θ ), Eq. (7.35) holds. Let us study the influence of the origin position inside the obstacle on the reconstruction process. The numerical investigation shows that when we use a symmetrical finite-difference representation for the derivatives: piθθ = (pi+1 – 2pi + pi–1 )/(∆θ )2 (i is the number of a node in the grid, ∆θ is the step of the discretization), then we come numerically to the solution ρ(θ ) ≡ R cited above. Moreover, symmetrical location of the origin is obtained for any symmetric obstacle, since we use the symmetric finite-difference form for the derivative. This result can be derived from the observation that any symmetry in the vector {pi } involves the same symmetry in ρ(θ ), by using uniqueness of the inversion considered in Sections 7.1, 7.2. For an arbitrary obstacle, which does not possess any symmetry, the reconstruction algorithm leads to an origin that can change itself with a change of the finite-difference scheme. Some more details concerning this problem can be found in the author’s work Sumbatyan et al. (1993).

Helpful remarks Note that the studied inverse problem is nonlinear both in the 3D and 2D cases. And it is wonderful that its formulation in terms of the Minkowski support function reduces to the linear ordinary differential equation with constant coefficients (7.35) admitting exact analytical solution in explicit form.


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7.4. Exact Explicit Inversion of the Basic Operator in the Case of Axial Symmetry Let an axially symmetric obstacle bounded by the surface S be spherically irradiated by a high-frequency acoustic wave; the propagation time of the reflected pulse t(q ) and the real amplitude of the reflected wave |A(q )| are known for any direction q . We denote the convex parts of the surface S by Sex and the nonconvex parts by Sin . We restrict our consideration to an obstacle with a simply connected surface S allowing no more than two points of single mirror reflection at any direction of irradiation in the echo mode. It is shown in Druzhinina and Sumbatyan (1992) that this condition is necessarily satisfied if the angle between the normals erected to the surface S at any two of its points within one nonconvex region is acute. All these facts will be used in the forthcoming study, but here let us restrict the consideration by axially symmetric convex obstacles. We consider the axially symmetric surface S bounding an obstacle in a Cartesian coordinate system OXY Z , where the OZ -axis coincides with the axis symmetry of S , since the surface S is axially symmetric, for its intersection with any axial plane. For the sake of definiteness, we consider the axial plane XOZ . In this plane, the unit vector q has the coordinates {sin θ , 0, cos θ }, and A = A(θ ). Here, θ is the angle between the axis of symmetry OZ and the direction q . First of all, by the known travel time of the reflected pulse t(θ ), we determine the distance p(θ ) to a tangent at the point of specular reflection, which, in the axial plane, is given by the equation x sin θ + z cos θ – p(θ ) = 0. (7.37) By this known function p(θ ), we construct a convex hull of the surface: x(θ ) = –p0 (θ ) cos θ + p(θ ) sin θ ,

z (θ ) = p0 (θ ) sin θ + p(θ ) cos θ ,

(7.38)

which determines the smallest convex contour S0 enclosing the given one (see Preparata and Shamos, 1985). This ill-posed procedure will be described in more detail in Section 8.9. Instead here we describe a stable (well-posed) approach of reconstruction on the basis of the known back-scattered amplitude, i.e., by using the function A(θ ). In this case, the back-scattered amplitude in a far-field zone (we omit here a certain insignificant coefficient) is expressed in terms of the Gaussian curvature. It will be shown in the next section that in the case of axial symmetry this is described by the following nonlinear ordinary differential equation of the second order: 2 dp dp +p cot θ + p = γ –1 (θ ), (7.39) dθ dθ

where γ (θ ) is the known Gaussian curvature. Note that the first factor of operator (7.39) of the axially symmetric inverse problem is the operator of the inverse problem for a plane case discussed in the previous section, the main equation of which is quoted here again d2 p(θ ) + p(θ ) = ρ(θ ), dθ 2

(7.40)

p(θ ) = p0 (θ ) + p∗ (θ ), p0 (θ ) = C1 sin(θ ) + C2 cos(θ ), Z θ Z θ ∗ p (θ ) = sin θ ρ(θ ) cos θ dθ – cos θ ρ(θ ) sin θ dθ .

(7.41)

with its solution

θ0

θ0


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The solution of the nonlinear equation for an axially symmetric case (7.39) can be obtained in an explicit form by reducing it to a sequential solution of two linear differential equations of the first order with variable coefficients. Here we should note that, by using a differential operator of the first order with variable coefficients, the first differential factor with constant coefficients on the left-hand side of (7.39) can be expressed in terms of the second factor (also with variable coefficients): d dp dp d2 p + p = tan θ cot θ + p + cot θ + p . (7.42) dθ 2 dθ dθ dθ This representation makes it possible by the substitution dp(θ ) cot θ + p(θ ) = U (θ ) dθ

(7.43)

to reduce the initial equation to a Bernoulli equation tan θ

dU (θ + U (θ ) = U –1 (θ )γ –1 (θ ). dθ

(7.44)

Following the general theory of differential equations, by the substitution U 2 (θ ) = f (θ ), equation (7.44) can be reduced to a linear equation for the function f (θ ), df (θ ) + 2 cot θf (θ ) = 2γ –1 (θ ) cot θ . dθ

(7.45)

We seek the solution to this equation in the form of the product of two functions f (θ ) = u(θ )v (θ ). We select the function v (θ ) as a particular solution of the homogeneous equation v (θ ) = sin–2 θ . Then, u(θ ) satisfies the equation du(θ ) = γ2–1 (θ ) sin 2θ , dθ

(7.46)

the general solution of which is u(θ ) =

Z

θ

γ –1 (θ ) sin 2θ dθ + C1 .

(7.47)

θ0

Thus, we determine the function U (θ ), which is the general solution to equation (7.44): U (θ ) = f 1/2 (θ ) =

Z

1 sin θ

θ

γ2–1 (θ ) sin 2θ dθ + C1

θ0

1/2

.

(7.48)

Once U (θ ) has been found, the right-hand side of equation (7.43) in P (θ ) becomes known. The solution to the linear differential equation (7.43) is obtained in the same way as the solution to equation (7.45). As a result, the function P (θ ) can be represented as (Z " ) Z 1/2 # θ

P (θ ) = cos θ

–2

θ

cos θ

θ0

γ2–1 θ sin 2θ dθ + C1

+ C2

.

(7.49)

θ0

The results presented here follow the author’s work (Boyev and Sumbatyan, 1999).


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Helpful remarks It should be noted that both the plane, Eq. (7.41), and axially symmetric, Eq. (7.49), general solutions to problems contain, in accordance with a classical theory of ordinary differential equations, a pair of arbitrary constants C1 and C2 . In both the plane and axial symmetric cases the function p(θ ) gives a plane boundary curve. The only difference is that in the case of axial symmetry the boundary curve represents a certain axial cut. In order to determine uniquely the function p(θ ) in Eq. (7.49) (as well as in Eq. (7.41) for the plane problem) for a concrete surface S , we need to assign two boundary (or initial) conditions. Usually such conditions are predetermined by a periodicity of the solution.

7.5. Nonlinear Differential Operator of the Three-Dimensional Inverse Problem In the generic 3D problem the problem cannot be resolved by any analytical method, since this reduces to the Minkowski problem, which can be represented by a strongly nonlinear equation (7.11). In Section 7.2 we established some general theoretical results on solvability and uniqueness of this problem, and remarked that there is no known algorithm in the literature with a concrete reconstruction method. So this is the principal goal of our investigation in the present section. Let us pass in Eq. (7.11) from Cartesian coordinates to the spherical ones (r , θ , ϕ). Since we consider obstacles with a closed boundary surface S , the end of the external normal vector traces over a sphere of radius r . When passing to spherical coordinates, in expressions for the derivatives Pi , Pij , let us recall the following properties of homogeneity (see Section 7.1): P ∂Pi ∂P = , = 0, i = 1, 2, 3. (7.50) ∂r r ∂r Eq. (7.50) implies that the functions P (q) and p(q) coincide to each other over the unit sphere r = 1. It is rather convenient to solve this equation on the unit sphere with respect to the unknown function p(θ , ϕ). Let us write out, in terms of the function p(θ , ϕ) on the unit sphere, the expressions of the first derivatives of P , ∂p ∂p cos θ cos ϕ – ∂θ ∂ϕ ∂p ∂p cos θ sin ϕ + x2 = P2 = p sin θ sin ϕ + ∂θ ∂ϕ ∂p sin θ , x3 = P3 = p cos θ – ∂θ x1 = P1 = p sin θ cos ϕ +

sin ϕ , sin θ cos ϕ , sin θ

and the second derivatives of P , 2 ∂p ∂ p 2 2 + p cos θ cos ϕ + cot θ + p sin2 ϕ P11 = ∂θ 2 ∂θ ∂p sin2 ϕ ∂ 2 p ∂ – p cot θ + , – 2 cot θ sin ϕ cos ϕ ∂ϕ ∂θ sin2 θ ∂ϕ2 2 ∂p ∂ p 2 2 + p cos θ sin ϕ + cot θ + p cos2 ϕ P22 = ∂θ 2 ∂θ ∂p cos2 ϕ ∂ 2 p ∂ – p cot θ + , + 2 cot θ sin ϕ cos ϕ ∂ϕ ∂θ sin2 θ ∂ϕ2

(7.51)

(7.52)

(7.53)


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∂2p + p sin2 θ , P33 = ∂θ 2 2 ∂p ∂ p 2 + p cos θ sin ϕ cos ϕ – cot θ + p sin ϕ cos ϕ P12 = ∂θ 2 ∂θ ∂ ∂p sin ϕ cos ϕ ∂ 2 p 2 2 , – p cot θ – + cot θ cos ϕ – sin ϕ ∂ϕ ∂θ sin2 θ ∂ϕ2 2 sin ϕ ∂ ∂p ∂ p + p cos θ cos ϕ – – p cot θ , P13 = – sin θ ∂θ 2 sin θ ∂ϕ ∂θ 2 cos ϕ ∂ ∂p ∂ p + p cos θ sin ϕ + – p cot θ . P23 = – sin θ ∂θ 2 sin θ ∂ϕ ∂θ

(7.54)

(7.55)

(7.56) (7.57)

In geographical coordinates (θ , ϕ) of the unit external normal, Eq. (7.11) can thus be rewritten for the function p(θ , ϕ) as follows: 2 ∂ ¯ ¯ ϕ (p) – 1 ¯ θ (p) cot θ L ¯ θ (p) + 1 Lϕ L = γ –1 (θ , ϕ), (7.58) L ( p ) Lθ L θ sin2 θ sin2 θ ∂ϕ where

∂p ¯ θ (p) = ∂p – p cot θ , + p cot θ , L ∂θ ∂θ (7.59) ∂p ¯ ϕ (p) = ∂p – p cot ϕ. + p cot ϕ, L Lϕ (p) = ∂ϕ ∂ϕ Such a representation of the main operator (7.11) of the considered inverse problem in the form (7.58)–(7.59) allows us to extract operators of the respective inverse problems for the two-dimensional case and the case with axial symmetry (see Sections 7.3 and 7.4). As we could see above, in both cases the main differential operator permits explicit inversion. In the considered generic 3D case we can arrange a mapping of the considered closed surface S to the unit sphere. Such a mapping is achieved if one draws unit vectors, with their origin being put at any fixed center, in direction of the unit normal to the corresponding point of the surface. Reconstruction of the function p(θ , ϕ) can be realized by a numerical step by step algorithm. In this numerical algorithm, on the unit sphere surface, central differences are used for the first (p0θ , p0ϕ ) and the second (p00θθ , p00θϕ , p00ϕϕ ) derivatives over the mesh which consists of parallels θ = θi (0 ≤ θi ≤ π ) and meridians ϕ = ϕj (0 ≤ ϕj < 2π ), with a uniform mesh interval, which is assumed to be the same along θ and ϕ: h θ = π/N , hϕ = 2π/(2N ), hθ = hϕ = h. We have at node (i, j ) Lθ (p) =

∂p pi,j+1 – pi,j–1 ∂p pi+1,j – pi–1,j = , = , ∂θ 2h ∂ϕ 2h ∂2p pi,j+1 – 2pi,j + pi,j–1 ∂ 2 p pi+1,j – 2pi,j + pi–1,j = , = , ∂θ 2 h2 ∂ϕ2 h2 pi+1,j+1 – pi–1,j+1 – pi+1,j–1 – pi–1,j–1 ∂2p = . ∂θ∂ϕ 4h2

(7.60)

If we substitute relations (7.60) into Eq. (7.58), then we come to a nonlinear algebraic system, which can be written in the operator form Ap = γ2–1 ,

p = {p1 , p2 , . . . , pM },

γ –1 = {γ –1 (θ1 , ϕ1 ), . . . , γ –1 (θM , ϕM )},

(7.61)


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where M is the total number of the mesh nodes. In order to solve this nonlinear operator equation, we apply iterations of the Newton–Kantorovich method (see Kantorovich and Akilov, 1982), to solve this system pn+1 = pn + zn ,

(7.62)

with zn being defined from the linear algebraic system Jzn = γ –1 – Apn .

(7.63)

Here J = A0 is a Jacobian of system (7.61). As regards a concrete implementation of this method, note that the Newton–Kantorovich method requires in our problem a linear array of unknown elements p k . In order to arrange a one-dimensional vector of unknowns pk on the two-dimensional mesh, it is necessary to enumerate the nodes (θi , ϕj ) so as to obtain a linear array. We propose the following strategy. For each ϕj sequential enumeration corresponds to increasing of θ i : h ≤ θi ≤ π – h. Then the process jumps to the next value ϕj+1 , by increasing of the index j . The last two nodes are two poles. As a result, dimension of the nonlinear algebraic system (7.61) is M = 2N (N – 1) + 2. Let us give explicit expressions for the elements of the Jacobian. i Pϕϕ 2 2 i i i i Pθθ + P + 1 – 2 Pθ cot θi + +P , Ji,j = – 1 – 2 2 h sin θi h sin2 θi i 1 Pϕϕ cot θi i i i i Pθθ + P + 2 Pθ cot θi + +P , Ji,j±1 = – ± 2h h sin2 θi cos θi 1 i i (7.64) Pθθ + Pi ± Pϕi cot θi – Pθϕ , Ji,j±N = – – 2 2 3 h sin θi h sin θi 2 i Pϕi cot θi – Pθϕ , Ji,j±(N +1) = 2 2 h sin θi 2 i i P cot θi – Pθϕ . Ji,j±(N –1) = – 2 2 h sin θi ϕ

Figure 7.3 demonstrates the reconstruction of a three-dimensional obstacle, which is a circular cylinder with smoothed edges, when the back-scattered amplitude is known with the step of π/8 in both φ and θ . The third iteration is practically indistinguishable from exact surface. In order to produce data on the back-scattered diagram, we have solved the direct diffraction problem by the BIE method (see Chapter 2).

Figure 7.3. Reconstruction of a circular cylinder with smoothed edges


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Some more details on the discussed reconstruction method can be found in the author’s work (Sumbatyan and Troyan, 1992).

Helpful remarks It should be noted that the Jacobian is a sparse matrix, with the most part of its nontrivial elements being situated near the main diagonal and its two neighbors. However, generally, the structure of the matrix is not banded. Consequently, the method to solve this system should take into account that its matrix is very sparse. A number of efficient methods for solving sparse linear algebraic systems can be found, for instance, in Pissanetski (1988).

7.6. Reconstruction of Nonconvex Obstacles in the High-Frequency Range: 2D Case We assume that the acoustic medium contains a nonconvex obstacle (Fig. 7.4) with piecewise smooth positive-curvature boundary L with M corner points. For this contour, we construct a convex hull L0 (i.e., the smallest convex contour containing L 0 ). Let SM Lm ( m–1 Lm = L ∩ L0 ) be the common segments of contours L and L0 , and lm be the SM segments that belongs only to L ( m=1 lm = L \ L0 ). We assume that, when L is echo sounded at an arbitrary angle α, the number of single scattering points is, at most, two. It has been demonstrated that, at boundary segments l m , multiple scattering, which induces a back-scattering ray, must be absent if two normals n m1 and nm2 erected at two arbitrary points of lm make an acute angle βm (see Druzhinina and Sumbatyan, 1992). In view of all the limitations imposed on L, each segment lm may only be constituted by two intersecting convex arcs lm1 and lm2 , with the angle between their normals at the point of intersection being acute.

n1

y

n2 l11 P1 l12

P2

L1

a x

O l21

l22

L2 n21

n22 b2

Figure 7.4. Diffraction of a plane acoustic wave by a 2D nonconvex obstacle When the obstacle is echo-sounded at an arbitrary angle α, there always exists at least one point of specular reflection that lies on one of the boundary segments L m . In the


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general case, near lm , the diffraction pattern is more complicated. However, the above assumptions imply that, for a given angle of incidence, there exists one point of simple singular reflection lying on one of the arcs lm1 and lm2 of boundary segment lm . The reconstruction algorithm rests on the following consideration. We assume that, in circular echo-sounding, the known parameters are the arrival time of the reflected pulse t(α) and the real-valued amplitude of the reflected wave |A(α)|, 0 ≤ α < 2π . The knowledge of t(α) completely defines the convex envelope of the object boundary, because it determines the support function p(α) defining the distance from some selected center O inside the SM region to the convex boundary m=1 Lm at the point with normal n(α) (see Fig. 7.4). The convex hull L0 of contour L is the envelope of the family of these tangents. Its Cartesian coordinates are defined by x = –p0α (α) sin α + p(α) cos α,

y = p0α (α) cos α + p(α) sin α.

(7.65)

Once the convex envelope L0 (7.65) is constructed using the known function t(α), the arcs lm1 and lm2 may be determined for boundary segments lm using the complex amplitude of the reflected wave A(α). In the range 0 ≤ α < 2π , where there exists only reflection from convex parts of Lm , the amplitude of the reflected wave is A∼

ei(2kR–π/4) r , √ 2 πk R + R ρ1 cos ϑ

(7.66)

where R is the distance between the point of observation and the boundary contour L, cos ϑ = 1, because in the case under consideration, ϑ = 0, k is the wave number, and ρ 1 is the contour curvature radius at a point of specular reflection on a convex boundary segment. In the far field (R → ∞ ), the last formula may be approximately written as p (7.67) A(α) ∼ ρ1 (α) e–2ip1 (α)k .

In other intervals of α, waves reflected from one point of L m (2) and one point on lm overlap. In this case, the amplitude of the reflected wave reduces to the form p p (7.68) A(α) = ρ1 (α) e–2ip1 (α)k + ρ2 (α) e–2ip2 (α)k .

In formulas (7.66)–(7.68), ρ1 (α) and ρ2 (α) are the curvature radii at points of segments Lm and lm respectively; p1 (α) and p2 (α) are the respective support functions. The complex back-scattering amplitude, defined by these expressions, has the form independent of the distance between the obstacle and the point of observation. Then, for the intensity of the reflected wave, we have p (7.69) F (α) = F (α, k ) = |A(α)|2 = ρ1 (α) + ρ2 (α) + 2 ρ1 (α)ρ2 (α) cos[2k (p1 – p2 )].

We note that formula (7.68) allows for the contributions from the points of ray reflection into the amplitude of the scattering wave. In the lit zone, as k → ∞, the contribution from other points of the smooth boundary segments is a quantity of a higher order of smallness than k –1 (see Section 1.4). However the boundary has corner points. In the considered case free from multiple reflections, their contribution into A(α) is of the order of k –1/2 , and as k → ∞, it becomes a small quantity. In classical acoustic holography, it is assumed that the main information is contained in the oscillating component of formula (7.69). In this section we suggest an alternative treatment that rests on the following idea. We note that, in formula (7.69), only the last term


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depends on k . At high frequencies, it is a strongly oscillating function of a. Consequently, the first two terms are slowly oscillating functions of α only if the contour has no segments with sharp variations of curvature. Using filtering (e.g., digital), we retain only the slowly variable component in F0 (α); then F0 (α) = ρ1 (α) + ρ2 (α).

(7.70)

Now, on the convex part of Lm , the curvature radius ρ1 (α) may be deemed known after the convex envelope of the contour has been constructed. This construction defines the curvature of arcs lm1 and lm2 on lm as ρ2 (α) = F0 (α) – ρ1 (α).

(7.71)

As noted above, for the selected class of boundary contours L, each segment l m consists of two convex intersecting arcs lm1 and lm2 having a common tangent with the convex envelope L0 (see Fig. 7.4). In order to reconstruct each of these arcs, we can use the approach developed in Section 7.3, which reduces the reconstruction of the convex parts of the boundary contour to the solution of an ordinary differential equation of the second kind with constant coefficients, Section 7.3, whose exact analytical solution is given explicitly. However, an alternative efficient approach can be found upon the natural equation of the plane curve (see, for example, Pogorelov, 1961): sα0 (α) = ρ2 (α),

(7.72)

where s is the arc length counted from the point of tangency. In the Cartesian coordinate system, this arc is described by Z s Z s x(s ) = – sin α(s ) ds + x0 , y (s ) = cos α(s ) ds + y0 , (7.73) 0

0

where

s = s (α) =

Z

α

[F0 (α) – ρ1 (α)] dα,

(7.74)

α0

and x0 and y0 are the Cartesian coordinates of the intersection of the arc and the tangent to the convex envelope L0 . Equations (7.73) take into account the fact that, at any point of the contour, the tangent is directed at π/2 to the normal. We note also that the angle α 0 is defined by the normal to the point of tangency; this normal is common for adjacent segments Lm and Lmj , j = 1, 2, of the contour L. Indeed, this point is a point of tangency of a curvilinear arc belonging to the convex boundary segment and the above constructed convex envelope of the object. Since the boundary contour is smooth, the normal n(α 0 ) is uniquely defined at the point of tangency. An appropriate choice of the upper bound of α plays an important role in using formulas (7.73) and (7.74). This bound must coincide with the visible segment of arc l mj ; this is equivalent to the fact that formula (7.74) is valid only for those α which permit specular reflection from arc lmj . Evidently, these intervals of α are associated with the presence of two points of specular reflection (one on Lm , the other on lmj ). Thus, the considered intervals refer to α, for which the back-scattered amplitude |A(α)| is a strongly oscillating function in the high frequency range (see formula (7.69)). For a better insight into this problem, Figure 7.5 shows the pattern of back-scattering from an acoustically √ soft obstacle configured as two equal circles of radius R with a centerto-center distance 2 R (R = 3λ, and λ = 2π/k is the wavelength). The back-scattering pattern A(α) was determined by solving the direct diffraction problem by the BIE method (see Chapter 2).


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|A| 8

q

6 4 2

0

90

180

270

q, degrees

Figure 7.5. Amplitude of scattering by a pair of intersecting circles The function |A| was defined by solving a direct scattering problem by BIE. Note that the angle between normals to arbitrary pair of points over any internal part S in is less than π/2. The only case, when a limiting value π/2 is reached, is related to the points of intersection between the circles. Therefore, as mentioned above, only points of single specular (mirror) reflection exist in the region Sin , for arbitrary direction of incidence, and multiple ray re-reflections are impossible. It is clear from Fig. 7.5, when passing over the segment 0 ≤ θ < 2π , that there are both intervals with slow variance of the amplitude |A(θ |) and intervals with high oscillations of this function. Approximate values of θ 1 and θ2 are just chosen as points of jump from slow to high oscillations. In practical implementations of the numerical algorithm, the upper integration limit, α 1 , in (7.74) is determined approximately because we cannot exactly determine the boundary between segments of quick and slow variation of the known scattering amplitude. In all considered tests, the arcs of each pair lm1 , lm2 go somewhat beyond the point of intersection. Therefore, in the reconstruction of the simply connected domain, one must select only those parts of these arcs, which lie from the point of tangency with the convex envelope L 0 to the point of their intersection. The proposed reconstruction algorithm involves a natural method for extracting the low-frequency component F0 (α) of F (α) = |A(α)|2 . It is based on the fact that, upon substituting the experimentally determined F (α) in (7.74), the integration in this formula, Z α [F (α) – ρ1 (α)] dα, (7.75) s= α0

provides the filtering of the low-frequency component. For well-defined computation of the integral (7.75), we should define correctly the value of the small integration step. The convex envelope defines the characteristic dimension d of the obstacle. If we assume that ten points per wavelength suffice for well-defined computation of this integral, then the number of nodal points in the integration may be estimated as 10 d/λ. Formula (7.75) takes into account that the main term of the asymptotic expansion, obtained from the integral of the last oscillating term in (7.69) by the stationary phase method (see Section 1.4) as k → ∞, is of the order of O (k –1 ). Numerical experiments indicate that this natural method of filtering provides a more accurate reconstruction than many known procedures. We can thus state that the calculation of


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Figure 7.6. Reconstruction of an obstacle configured as a pair of intersecting circles integrals themselves in these problems provide an efficient filtering of the highly oscillating integrands if the step of integration is fixed and does not depend on the wavelength. A reconstruction of a pair of unit circles (Fig. 7.6 demonstrates the real circular backscattered amplitude |A|) was performed. The accuracy of the reconstruction is not worse than 2–3 % uniformly all over directions of incidence. Note that solid lines here represent the true contour, and dashed lines represent the contour reconstructed by the outlined method. The results presented here follow the author’s work (Boyev et al., 1997).

Helpful remarks This algorithm was tested by reconstructing the geometries, which do not belong to the considered class. One of them is presented in Fig. 7.7, where the angles between the normals to arcs lm1 and lm2 may be obtuse and even close to π . The region consists of two tangent circles with R = 3λ. In this case, segments l m1 and lm2 admit multiple reflections of rays. However, the accuracy of the reconstruction remains acceptable and, seemingly, the validity range of this method is really wider than the class of contours specified in this paper.

Figure 7.7. Reconstruction of an obstacle in the form of a pair of tangent circles This conclusion is confirmed also by the reconstruction of a smooth contour in the form of a three-leaf rose, ρ(ϕ) = a(2 + cos 3ϕ), 0 ≤ ϕ < 2π , (7.76) shown in Fig. 7.8. This contour does not belong to the considered class because it contains segments with negative curvature. Dashed lines correspond to a = 2λ (curve 1), a = 4λ (curve 2), a = 6λ (curve 3). The characteristic size of the obstacle is 10λ, 20λ, and 30λ, respectively. It is apparent that the reconstruction accuracy increases with frequency. This contour also does not belong to the class, considered at the present work, because it contains some parts with negative curvature. Dashed lines correspond there to three different values of the parameter a, with λ being the wavelength. For all that, the average size of the obstacle is equal to 10λ, 20λ, 30λ, respectively. As can be seen from Fig. 7.8, the accuracy of the reconstruction improves as frequency increases. Therefore, in practice, the proposed method is applicable to a wider class of boundary surfaces than those assumed.


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1 2

3

Figure 7.8. Reconstruction of a three-leaf rose: solid line, exact boundary; dashed line, reconstructed boundary; 1, a = 2λ, 2, a = 4λ, 3, a = 6λ

7.7. Reconstruction of Nonconvex Obstacles in the High-Frequency Range: 3D Case In this section we expand some results of the previous sections to the 3D case. Let there be provided a circular irradiation by high-frequency harmonic waves, for arbitrary angle of incidence, with respect to an obstacle detected in an acoustic medium, under a scanning around the obstacle (S denotes the closed boundary shape of the obstacle). Such an opportunity can be provided in practice, for instance, under immersion scanning of an elastic body placed into a liquid. With this irradiation the echo-impulse time of flight t(q) and the real back-scattered (reflected) amplitude |A(q)| are assumed to be known, for arbitrary direction of incidence q. Let us start first from the reconstruction of convex part of the boundary surface. Let S ex denote convex (external) parts of the surface S , and S in is its nonconvex (internal) parts. Let us restrict the consideration by the obstacles which 1) have simply connected boundary surface S with smooth convex parts Sex and smooth junction between convex Sex and nonconvex Sin ones; 2) admit no more than two points of single mirror (specular) reflection in echo regime, for arbitrary direction q. It is proved in Druzhinina and Sumbatyan (1992) that the latter is automatically satisfied if the angle between two arbitrary external normals over the same nonconvex part Sin is acute. Figure 7.9 demonstrates a typical example of obstacle from the considered class of surfaces. An acute angle between two normals is shown on the left side of Fig. 7.10, for axial cut of an obstacle of axial symmetry. Let us relate the boundary surface S to a Cartesian coordinate system OX 1 X2 X3 , with the origin O being taken inside the surface. The direction of incidence q and the unit external normal n to the surface S may be related to a geographical coordinate system θ , ϕ (θ is the latitude and ϕ is the longitude) or, alternatively, may be given by the direction cosines (cos α, cos β , cos γ ). Under irradiation in echo regime, along arbitrary direction q = {cos α, cos β , cos γ } there is always a certain point of specular reflection on the convex part S ex . Diffraction by the regions Sin is generally more complex. However, assumptions 1) and 2) guarantee that the surface Sin has only points of simple mirror reflections, and there are no multiple ray re-reflections. The proposed reconstruction algorithm is the following. First, starting from the known


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Figure 7.9. A typical shape of the obstacle under consideration

n1

x3

q1 Sex

n2 q2

Sin

n2 q

x1

Figure 7.10. Convex hull of the nonconvex obstacle function t(q) a convex hull S0 of the boundary surface S can be completely defined, i.e., a minimum convex surface containing S inside (see Fig. 7.10). Indeed, t(q) determines the function p(q) that defines the distance from the origin O to the plane tangential to the convex boundary, whose equation is given as follows: Q(x1 , x2 , x3 , α, β ) = x1 cos α + x2 cos β + x3 cos γ – p(α, β ) = 0.

(7.77)

The convex hull S0 is an envelope of a two-parameter family of these tangential planes. We have taken in (7.77), as a pair of independent parameters, the angles α and β , since the


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third direction angle γ is connected with them, by the trivial identity cos2 α + cos2 β + cos2 γ = 1.

(7.78)

Following a classical theory (see, for example, Pogorelov, 1961), the boundary points of the convex envelope have to be defined from the system  Q(x1 , x2 , x3 , α, β ) = 0,      ∂Q (x1 , x2 , x3 , α, β ) = 0, (7.79) ∂α    ∂Q   (x1 , x2 , x3 , α, β ) = 0. ∂β

The latter is equivalent to   x1 cos α + x2 cos β + x3 cos γ = p(α, β ), – x1 sin α cos γ + x3 sin α cos α = p0α cos γ ,  – x2 sin β cos γ + x3 sin β cos β = p0β cos γ ,

(7.80)

since γα0 sin γ = – sin α cos α/cos γ , γβ0 sin γ = – sin β cos β/cos γ , which directly follows from Eqs. (7.77), (7.78). The solution of the last linear algebraic system can be obtained, for instance, by using Cramer’s rule. Therefore, the Cartesian coordinates of strictly convex parts of the convex hull S0 are given by the following equations: x1 = [p(α, β ) + p0β (α, β ) cot β ] cos α – p0α (α, β ) sin α, x2 = [p(α, β ) + p0α (α, β ) cot α] cos β – p0β (α, β ) sin β , x3 = [p(α, β ) +

p0α (α, β ) cot α

+

(7.81)

p0β (α, β ) cot β ] cos γ .

Now we proceed to reconstruction of nonconvex regions of the boundary. As soon as the convex hull S0 is constructed, on the basis of the known time of flight function t(q), for reconstruction of internal parts of the boundary surface S the measured values of the echo-impulse amplitude |A(q)| can be used. All directions of incidence q are divided into two types, which is schematically shown in Fig. 7.10 as unit vectors q 1 and q2 . The first set contains those q, where only one point of specular reflection exists on the convex part S ex , in echo regime of the scanning. For such directions q the back-scattered far-field amplitude, up to some factor, is (see the previous sections of this chapter) –1/2

A(q) = γ1

(q) exp[–2ikp1 (q)].

(7.82)

For other directions q, for which there are points of specular reflections from the parts S in as well, the echo amplitude is given as follows: –1/2

A(q) = γ1

–1/2

(q) exp[–2ikp1 (q)] + γ2

(q) exp[–2ikp2 (q)].

(7.83)

Here γi = (R1(i) R2(i) )–1 , i = 1, 2 are the Gaussian curvatures; R1(1) , R2(1) and R1(2) , R2(2) are the principal curvature radii at the points of reflection from S ex and Sin , respectively; p1 and p2 are the distances from the origin to the tangential planes at the respective reflection points; and k is the wave number. The main information for reconstruction of the regions S in is contained in the real amplitude of the reflected wave: F (q) = |A(q)|2 = γ1–1 (q) + γ2–1 (q) + 2[γ1 (q)γ2 (q)]–1/2 cos{2k [p1 (q) – p2 (q)]},

(7.84)


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which is the sum of slowly varying terms γ1–1 , γ2–1 and a highly oscillating function of (q), under high-frequency irradiation by plane acoustic waves. Let us apply any filtration technique, to extract a slowly varying component (at large k ) from the function F (q) (see Cappellini et al., 1978): F0 (q) = γ1–1 (q) + γ2–1 (q). (7.85) So far as the function γ1 (q) is known from the convex hull S0 , the Gaussian curvature γ2 (q) at the points of the internal parts Sin of the surface S can be defined as γ2 (q) = [F0 (q) – γ1–1 (q)]–1 .

(7.86)

Shape reconstruction of the nonconvex parts Sin of the surface S can be performed on the basis of the Minkowski function, starting from the known function γ 2 (q), Eq. (7.86). Further study can be completed in more detail, for example, in the case of axially symmetric obstacle. Here the problem to reconstruct the function p(θ ) was studied in Section 7.4 and we could see that this admits an explicit analytical solution expressed through two arbitrary constants C1 and C2 , which can be uniquely determined in the following way. The surface S0 determines common points of the convex Sex and internal Sin parts of the real boundary surface (see Fig. 7.10, Sex and Sin junction). For every region Sin such points generate a pair of circles, disposed on parallel planes, with their centers being on the axis of symmetry OZ . As regards the points of these circles of the same axial cross-section, the normals to the surface are parallel to each other and, being related to the unit sphere, represent the same normal with certain coordinates (θ 0 , ϕ0 ). For both circles the ends of the unit normals form, in geographical coordinates, a parallel θ = θ 0 . This parallel is disposed in a spherical belt θ1 < θ < θ2 between the parallels θ = θ1 and θ = θ2 , which define a visible zone of the region Sin . Now the calculation of p(θ ) can be performed by using the double integral (7.49) over θ ∈ [θ1 , θ2 ]. The general solution (7.49) contains a pair of the unknown constants C 1 and C2 . The required initial conditions, to determine these constants, can be formulated on the intervals [θ1 , θ0 ] and [θ0 , θ2 ] as follows. Since the convex hull S0 is already constructed, this means that at the junction points between regions S in and Sex , i.e., at θ = θ0 , there are known p(θ0 ) = p0 and dp/dθ (θ = θ0 ) = p00 . It thus leads to a sequential definition of the constants in (7.49): C1 = sin2 θ0 (p00 cot θ0 + p0 )2 ,

C2 = p0 / cos θ0 .

(7.87)

Similarly, the unknown constants (7.41) can be found in the two-dimensional problem: C1 = p0 sin θ0 + p00 cos θ0 ,

C2 = –p0 cos θ0 + p00 sin θ0 .

(7.88)

It should be noted that the quantity (dp/dθ ) at θ = θ 0 is naturally treated, over every segment of integration, as a one-sided limit at θ → θ 0 , when approaching to the limiting point from the corresponding convex part Sex of the convex hull S0 . Hence, the calculation of the function p(θ ), which describes the internal region S in , is reduced to a numerical treatment of some single integral (two-dimensional case) or to a double integral (the case with axial symmetry), over the intervals [θ1 , θ0 ] and [θ0 , θ2 ]. For all that, the corresponding parts of the curves (in the two-dimensional problem) and the analogous parts of the curves in the axial cross-section (for the case of axial symmetry) can be reconstructed up to the point of their intersection. In practice, realization of the proposed method allows one to define the values θ 1 and θ2 with some error. That is because these values are indicated on the graph of the real


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amplitude |A(θ )| of the back-scattered wave as the points, where a slowly varying part of the diagram turns into a highly oscillating one, for corresponding part S in . This can be clarified again by the two-dimensional example shown in the previous section in Fig. 7.5, which presents a diagram of the real amplitude back-scattered from an acoustically soft obstacle, which √ is a junction of a pair of equal circles with radius R and distance between their centers 2 R (R = 3λ, where λ = 2π/k is the wavelength). Let us return back to investigation of the general three-dimensional inverse problem without axial symmetry. We can arrange a mapping of the considered closed surface S to the unit sphere. Such a mapping is achieved if one draws unit vectors, with their origin being put at any fixed center, in direction of the unit normal to the corresponding point of the surface. In this case at any internal part Sin of the surface S its visible zone generates a certain complex region on this unit sphere. Since there can be only points of single back specular reflections on the internal parts Sin , these parts can be a union of some convex pieces only. This guarantees that the region drawn by the ends of external unit normals is simply connected, so it can be defined by the inequalities ϕ 1 < ϕ < ϕ2 , θ1 (ϕ) < θ < θ2 (ϕ) (see Fig. 7.11). Particularly, this region may appear to be a spherical belt 0 ≤ ϕ < 2π , θ1 (ϕ) < θ < θ2 (ϕ).

q = q2(j)

q

P0 P1 P2 P3

q = q0(j)

q = q1(j) j1

j j2

Figure 7.11. A visible region of an internal part of the boundary surface A general sketch of the proposed reconstruction algorithm will be given for the case, when the internal boundary region Sin contains a curve of intersection of its composing convex surface parts (see Fig. 7.11). There is a curve θ = θ 0 (ϕ) that determines, for every ϕ, a unit normal of the convex hull to lines of junction between the given part S in and the neighbor parts Sex . Within the framework of the developed approach some numerical experiments were undertaken on reconstruction of a number of concrete nonconvex obstacles. For all examples below we first reconstruct a convex hull of the defect on the basis of measured echo–impulse time of flight, as described at the beginning of the paper. Note that both in the two-dimensional case and in the case of axial symmetry the Cartesian coordinates of the convex hull are given as x(θ ) = –p0 (θ ) cos θ + p(θ ) sin θ , z (θ ) = p0 (θ ) sin θ + p(θ ) cos θ ,

(7.89)

where θ is the angle between external normal and the axis OX 3 .


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The following two examples were considered, to test efficiency of the proposed algorithm. The first one (see Fig. 7.12) is related to a reconstruction, in the plane X 1 OX3 , of an axial cut of the axially symmetric obstacle, bounded by a surface join of two ellipsoids (both with the symmetry axis OX3 ): x21 x22 x23 x21 x22 x23 + + = λ2 , + + = λ2 . (7.90) 4 4 16 36 36 4 Shape reconstruction of these obstacles was performed on the basis of solution (7.49). x3 4l

2l

x1 -6l

-2l

2l

6l

-2l

-4l Figure 7.12. Reconstruction of an axially symmetric obstacle: solid line, exact boundary; dashed line, reconstructed boundary Another example was related to a three-dimensional obstacle without axial symmetry, bounded by a surface join of the two ellipsoids of rotation, around the axes OX 3 and OX2 (see Fig. 7.9): x21 x22 x23 x2 + + = 1. (7.91) x21 + x22 + 3 = 1, 9 4 9 4 In this case reconstruction of the boundary surface is performed on the basis of the proposed numerical step-by-step method. Note that the differential equation (7.58) should be treated here numerically in a spherical belt 0 ≤ ϕ < 2π , θ1 (ϕ) < θ < θ2 (ϕ) of the unit sphere. In both examples just considered, the accuracy of the convex boundary parts reconstruction is around 5–7 %, and that of the reconstruction of the internal parts is no worse than 10 %. The results presented in this chapter follow the author’s work (Vorovich et al., 2001).

Helpful remarks In order to implement the numerical algorithm developed for the considered inverse problem, we need to have an array of values of the real wave amplitude |A(θ )|, back-scattered from the obstacle along all possible directions. We constructed these values by a numerical treatment of the direct diffraction problem, with the help of the BIE method. Note that a typical size of the obstacles was nearly 8–12 wavelengths, which requires rather detailed mesh with a small mesh step.


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Chapter 8

Ill-Posed Equations of Inverse Diffraction Problems for Arbitrary Boundary 8.1. Ill-Posed Problems for Operator Equations of the First Kind: General Properties The theory of ill-posed (i.e., incorrectly posed) operator equations of the first kind was founded by A.N. Tikhonov (see, for example, Tikhonov and Arsenin, 1977) and had become a further natural development of some ideas of Hadamard and other classicists. Hadamard first formulated the concept of a well-posed (or correctly posed) problem for the operator equation Au = f , u ∈ U, f ∈ F, (8.1) where A is a linear operator acting from a Banach space U to a Banach space F . This includes the three points: 1) Equation (8.1) has a solution for any f ∈ F . 2) The solution is unique. 3) The solution is stable with respect to small perturbations of the right-hand side. In other words: small perturbation of f (in the F -metrics) involves small perturbation of u (in the U -metrics). Later on, it was recognized that for a correct treatment the last condition is of most importance. Thus, boundary value problems are well posed for elliptic systems, and initial (Cauchy) problems for hyperbolic systems, but not vice versa. Obviously, the existence of a bounded operator A –1 is sufficient for condition 3) to be valid. However, as shown in Section 1.6, in the case when A is compact the inverse operator A–1 does not exist. In the meantime, operator equations of the first kind (8.1) with a compact operator A are widespread in applied mathematics. A typical example is given by the Fredholm integral equation of the first kind Z b (Au)(x) = K (x, ξ )u(ξ ) dξ = f (x), a ≤ x ≤ b, (8.2) a

with a continuous (in both its arguments) kernel K (x, ξ ), a ≤ x, ξ ≤ b. In Section 1.5 it was proved that such an operator Au is compact in C (a, b), i.e., when U = F = C (a, b). It is also well known (see, for example, Kantorovich and Akilov, 1982) that operator (8.2) is compact on the Hilbert space U = F = L2 (a, b). In order to understand more clearly why the solution of Eq. (8.2) is unstable with respect to small perturbations of the right-hand side, let us suppose that Au 0 = f0 , u0 ∈ C (a, b), f0 ∈ C (a, b), and we take u1 = u0 + cos λt. It is obvious that in the case a = 0, b = 2π , ||u1 – u0 || = 1. If we denote f1 = Au1 , then Z 2π ||f1 – f0 || = ||A(u1 – u0 )|| = max K (x, t) cos λt dt x∈(0,2π) 0 (8.3) Z 2π iλt –iλt 1 K (x, t) e + e dt . = 2 max x∈(0,2π)

0


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Let the kernel K (x, t) be differentiable with respect to t. Then, according to results discussed in Section 1.4, ||f1 – f0 || = O (1/λ), λ → ∞. Therefore, for sufficiently large λ small perturbation of the right-hand side corresponds to a finite deviation of the solution. This example and some general theoretical ideas indicate the possible approach where we can guarantee stability with respect to small variations of f . Actually, the considered set of functions {u0 + sin(λt)} do not satisfy Arzela–Ascoli criterion (see Section 1.5) of compactness in C (a, b) since these functions are not equicontinuous when λ → ∞. Classical results of functional analysis (see, for instance, Kantorovich and Akilov, 1982) show that if we seek a solution to equation (8.1) on any compact set V ⊂ U , then the solution depends continuously upon f , which is provided by the following theorem. THEOREM. If a continuous operator A : U → F determines a one-to-one mapping of a compact set V ⊂ U onto a set AV = G ⊂ F , then the inverse operator A –1 is also continuous on G. This theorem gives you a quite natural way to construct a stable solution, which is called the method of solution by inspection. It involves the minimization of the discrepancy over some compact set V ⊂ U : u0 is taken as an approximate solution to equation (8.1) if min ||Au – f || = ||Au0 – f ||.

u ⊂V

(8.4)

It is clear that if u0 ∈ V , then the solution is stable with respect to small change of f , since A–1 is continuous on G = AV . These ideas allow us to operate with the concept of correctness in the sense of Tikhonov, instead of correctness in the sense of Hadamard. DEFINITION. Problem (8.1) is called well posed in the sense of Tikhonov if: 1) For some f0 there exists a solution u0 to this equation belonging to a compact set V : u0 ∈ V , Au0 = f0 . 2) This solution u0 is unique. It is clear that in this case operator A–1 is continuous on the set G = AV . So, if the right-hand side f of equation (8.1) is given with some small perturbation (which typically occurs when it is known with some small error) and f ∈ G, then the element u = A –1 f can be considered as a correct approximation to the exact solution u 0 = A–1 f0 , due to continuity of the operator A. The considered compact set V ⊂ U is called a class of correctness for the operator A. There can be listed evident cases when the above ideas can be applied efficiently. 1◦ . If we seek a solution inside a finite-dimensional subset V of the Banach space U . Actually, any finite-dimensional subset is compact in the Banach space. For example, let us consider the integral equation (8.1) on an interval symmetric with respect to the origin and of the length of periodicity of trigonometric functions Z π (Au)(x) = K (x, ξ ) u(ξ ) dξ = f (x), |x| ≤ π . (8.5) –π

If the kernel is, for example, even with respect to both its arguments x, ξ , and also the right-hand side f (x) is even with respect to x, then the solution of equation (8.5) is also even, and we may seek an approximate solution in the form un (x) =

n X

am cos(mx),

(8.6)

m=0


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where a few terms with n = 2–5 are usually sufficient to approximate well the exact solution u0 (x) (if the latter exists and is unique). In practice, you should minimize the discrepancy ρn = ||Aun – f ||. This can be achieved either by the inspection method or by a strict solution of the corresponding variational problem. In the latter case it is more convenient to consider equation (8.6) in the Hilbert space L2 (–π , π ). Then we arrive at a problem of minimization of the quadratic functional with respect to the unknown coefficients a m , m = 0, 1, . . . , n min ρn , ρn = ||Aun – f ||2 = (Aun – f , Aun – f ) {am } n n X X (8.7) = am aj (Aϕm , Aϕj ) – 2 am (Aϕm , f ) + (f , f ), ϕm (x) = cos(mx), m,j=0

m=0

whose solution can be found by using the well-known criterion of minimum: ∂ρ n /∂am = 0, m = 1, 2, . . . , n. This technique leads to the following linear algebraic system, to define coefficients am , m = 0, 1, . . . , n: n X (Aϕj , Aϕm ) aj = (f , Aϕm ), m = 0, 1, . . . , n. (8.8) j=0

It is interesting to note that this system coincides with the one which could be constructed by a certain type of Galerkin method (see Mikhlin, 1964) if you take a small number of basis functions ϕm (x). Another appropriate structure of the approximate solution is n X un (x) = am xm , (8.9) m=0

which in many cases leads to a more stable method than the one with trigonometric basis. The structure (8.9) represents a linear combination of monotonic functions, and the following paragraph explains why operation with monotonic functions usually leads to stable calculations. 2◦ . In many problems we possess some a priori information about qualitative structure of the solution. For example, in Section 3.6 we could see that in Arctic seas variation of the wave speed with depth represents typically a positive monotonically increasing function. One of the basic results in functional analysis states that any set of monotonic function is compact in L2 (a, b). The proof can be found, for example, in Tikhonov and Arsenin (1977). So, if we solve the problem on reconstruction of the wave-speed function in a stratified ocean from some measured data on acoustic wave scattering, then we can seek the unknown function by inspection in the class of monotonically increasing (or decreasing) functions. As an appropriate method we can quote here representation of the solution in the form (8.9) with unknown but positive (or negative) coefficients a m , m = 1, 2, . . . , n. The set of such functions consists of monotonic functions only. Note that stable calculations here are provided with arbitrary, even very large, number n, since we do not need to check here that the constructed set is finite-dimensional (in practice, with some small number n).

Helpful remarks When following the general strategy described above to construct a correct (i.e., stable) solution of the ill-posed operator problem of the first kind, we may encounter the case when the exact solution u0 related to the exact right-hand side f0 belongs to a compact set V ⊂ U : u0 ∈ V , Au0 = f0 ∈ G = AV , but a perturbed right-hand side f is outside of the set G = AV : f ∈ / G. In such cases equation (8.1) may not have solution in the classical sense. To this end, there is introduced the concept of a quasi-solution (see Tikhonov and Arsenin, 1977).


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DEFINITION. Let V be a compact set of the Banach space U . Then an element u ∗ ∈ V that provides a minimum for the functional min u∈V ρ(u) = ρ(u∗ ), ρ(u) = ||Au – f ||, is called a quasi-solution of equation (8.1). It is obvious that since V is a compact set, a quasi-solution surely exists (its uniqueness is under question). Besides, we can indicate the case when the quasi-solution is simultaneously a classical solution, namely when f ∈ G = AV . Some interesting and important results on quasi-solutions can be found in Tikhonov and Arsenin (1977).

8.2. Regularization of Ill-Posed Problems with the Help of Smoothing Functional As we could see from the previous section, a solution of the operator equation Au = f , where a compact linear operator A acts from one Banach space U to another Banach space F , i.e., A : U → F , is surely stable with respect to small perturbations of the right-hand side f if this solution is sought on a certain compact subset V ⊂ U and if f ∈ AV . However, in some cases the perturbed right-hand side fδ is outside of the set AV : fδ ∈ / AV , being at the same time very close to the exact function f : ||f – f δ || < δ , with some small positive quantity δ > 0. In such cases the results of Section 8.1 cannot guarantee stability of the solution, i.e., that the distance ||u – uδ || is small, where Auδ = fδ . This phenomenon seems to be a little strange. The equality Auδ = fδ implies that a solution of our operator equation exists in the space U , and in many cases it is even unique. However, small changes in f result in a large change in the solution, which is indeed a specific feature of equations of the first kind. In order to overcome this trouble and also for some other reasons Tikhonov (see Tikhonov and Arsenin, 1977) introduced the concept of regularization for ill-posed problems. The basic idea may be related to some fundamental ideas of Lavrentiev (1959, 1960). Let for simplicity operator A be self-adjoint and positive in some Hilbert space, A : H → H . If it is not so, then by multiplying the equation A u = f by A∗ from the left, we can achieve both required properties. In this case, as follows from results of Section 1.5, the operator A + αI , where I is a unit operator, is invertible with arbitrary positive ∀α > 0, i.e., (A + αI )–1 is continuous on H . Indeed, for positive operator A all its eigenvalues λ n are nonnegative, and so for the operator A + αI , α > 0, they all are positive. Therefore, the point λ = 0 is a regular value of the operator (A + αI – λI ). Then, according to the fundamental theorem of the operator theory (see Kantorovich and Akilov, 1982), we can conclude that (A + αI ) is invertible. Heuristically, if 0 < α 1, then the new operator A + αI is in some sense close to the main operator A. So we may hope that the thus constructed solution u α = (A + αI )–1 f can be accepted for a certain small α > 0 as an approximate solution to the main equation Au = f , stable with respect to small perturbations of f ∈ F , even when f is an arbitrary element in F . This idea was further developed by Tikhonov who introduced the concept of regularization by smoothing functional. The latter in the simplest treatment consists in the consideration of the variational problem of minimization of the functional Mα (u, f ) = ||Au – f ||2 + α ||u||2 ,

α > 0,

(8.10)

which is justified by the following theorem. THEOREM 1. If A is a one-to-one compact linear operator from a convex closed set V ⊂ H to the set AV = G ⊂ H and u ∈ V is a (unique) solution of the equation Au = f


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for a certain given f ∈ G, then there exists a unique solution u α ∈ V of the minimization problem uα : inf Mα (u, f ) = Mα (uα , f ). (8.11) u∈V

Proof. The functional (8.10) Mα (u, f ) = (Au – f , Au – f ) + α(u, u)

(8.12)

is a quadratic functional with respect to u. So it is infinitely differentiable and its Frechet differentials of the first (which is in fact its gradient) and the second order is easily calculated. Their structure can be obtained from a small variation of the functional as follows: δu Mα (u, f ) = 2(Au – f , Aδu) + 2α(u, δu) = 2(A∗ Au – A∗ f + αu, δu); hence, gradu Mα (u, f ) = 2(A∗ Au – A∗ f + αu).

(8.13)

Further, 2 δuu Mα (u, f ) = 2(A∗ Aδu + α δu, δu) = 2(Aδu, Aδu) + 2α(δu, δu) ≥ 2α ||δu|| 2 , (8.14)

so the functional Mα (u, f ) is strongly convex. It follows from some classical results of functional analysis (see, for example, Mikhlin, 1965; Groetsch, 1984; Morozov 1984) that any strongly convex functional attains its (unique) minimum value on any convex closed set V ∈ H . The theorem is proved.

THEOREM 2. If the solution of the variational problem (8.11), u α ∈ V , is an internal point of the set V , then it can be found from the uniquely solvable equation A∗ Au + α u = A∗ f .

(8.15)

Proof. It is known (see Tikhonov and Arsenin, 1977; Groetsch, 1984; Mikhlin, 1965) that under the conditions of this theorem the problem of minimization of the strongly convex functional by reducing to Euler’s equation leads to the necessary and sufficient condition: grad Mα = 0, which by taking into account Eq. (8.13) is equivalent to Eq. (8.15). The theorem is proved. THEOREM 3. Under the conditions of Theorems 1 and 2, if u ∗ is a (unique) exact solution of the equation Au = f , then ||uα – u∗ ||,

α → 0,

if

δ2 →0 α

and ||fδ – f || ≤ δ .

(8.16)

The proof to this theorem can be found in Tikhonov and Goncharsky (1987), and for a more general case, in Tikhonov and Arsenin (1977). This theorem gives an efficient instrument to choose an appropriate value of the regularization parameter α. Indeed, if the error on the right-hand side f is small enough being of the order of δ , then the regularization parameter α cannot be chosen too small. The optimal choice is such that α ∼ δ 2–ε , ε > 0. In some more details this important question on optimal choice of the parameter α will be discussed below. The basic idea of Tikhonov’s regularization justified by Theorems 1–3 is related to the fact that the smoothing functional Mα (u, f ) is strongly convex, so that it attains a minimum at the point u = uα . This allowed Tikhonov to propose an alternative concept of regularization by more smooth stabilizers. One of possible approaches is constructed by introducing the so-called stabilizing functional Ω(u), which is positive: Ω(u) ≥ 0, ∀u ∈ V , and continuous over V , so that the exact solution u ∗ of the equation Au = f belongs to the


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set V : u∗ ∈ V . Tikhonov proves that if the functional Ω(u, f ) is such that the set Ω(u) ≤ d is a compact subset in V for ∀d > 0, then the listed conditions are sufficient for the smoothing functional Mα (u, f ) = ||Au – f ||2 + αΩ(u), α > 0, (8.17) to have a unique minimum value inf u∈V Mα (u, f ) for arbitrary positive α, and for all that uα ∈ V . Moreover, under these conditions Theorems 1 and 3 are valid (see Tikhonov and Arsenin, 1977). Therefore, Eq. (8.15) supplies you with a more general instrument to construct a stable approximate solution to the operator equation of the first kind Au = f , due to a general form of the functional Ω(u). Besides, these results remain valid if the initial space U is not Hilbert but only metric. It should be noted that the analogue to Theorem 2 in this more general case can be obtained by reducing the minimization problem to the Euler equation, which in this problem can be written out in the following form: A∗ Au + α Ω0 (u) = A∗ f ,

(8.18)

where Ω0 (u) is the Frechet derivative of the functional, which is the same as its gradient. The approach described at the last part of the previous discussion turns out very fruitful in numerical treatment of the Fredholm integral equations of the first kind Z b Au = f ∼ K (x, ξ )u(ξ ) dξ = f (x), x ∈ (a, b). (8.19) a

Let us consider this operator equation in the Banach space of continuous functions: K (x, ξ ) ∈ C [(a, b) × (a, b)], f (x) ∈ C (a, b), A : C (a, b) → C (a, b). In the ambit of Tikhonov’s regularization we can put V = W21 (a, b) ⊂ C (a, b) = U . In this case the norm in W21 (a, b) can be accepted as the stabilizing functional Z b 2 q0 u2 (x) + q1 u0 2 (x) dx, q0 , q1 > 0, (8.20) Ω(u) = ||u||2,1 = a

since the set Ω(u) ≤ d is here a sphere in W21 (a, b), which is known to be compact in C (a, b) (see, for instance, Mikhlin, 1964; Tikhonov and Arsenin, 1977). Here q 0 , q1 are some positive constants. It now becomes clear that the functional (8.20) is a generalization of Ω(u) = ||u||22 in the space L2 (a, b) since in the case q1 = 0 the former is reduced to the latter. Within such a treatment the Euler equation for the variational minimization of the smoothing functional can be explicitly constructed since Z b Z b 0 0 δΩ(u) = 2 q0 uδu + q1 u δu dx = 2 q0 u – q1 u00 δu dx, (8.21) a

a

if we fix the values of the unknown function u(x) at the endpoints of the interval (a, b). After these preliminary transformations, a stable solution to equation (8.19) can be found from Euler’s equation: Z b K1 (x, ξ )u(ξ ) dξ + α q0 u(x) – q1 u00 (x) = f1 (x), (8.22) a

where

K1 (x, ξ ) =

Z

a

b

K (t, x)K (t, ξ ) dt,

f1 (x) =

Z

b

K (t, x)f (t) dt.

(8.23)

a


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It should be noted that here f1 = A∗ f , and the kernel in the integral operator in (8.23) is evidently related to the operator A∗ A. Now some words about the choice of the small parameter of regularization α. There are known several approaches to this problem (see Tikhonov and Arsenin, 1977). The following two of them seem to be quite natural and very efficient. 1. If the error of the measurement in the known data-in function is δ , and this quantity is a priori known from some considerations, then this can helpfully serve to estimate the value α, due to the estimate ||f – fδ || ≤ δ (cf. above). Indeed, if we start to solve equation (8.22) for relatively large α, then the solution u α will be very stable (as a solution of an operator equation with invertible operator), but too far from the exact solution u ∗ . In other words, the deviation ||uα – u∗ || is too large, and so is the discrepancy ||Auα – f ||. If we gradually decrease α, then, according to Theorem 3, u α → u∗ and so, by continuity of the operator A, the discrepancy ||Auα – f || → 0. However, it is not reasonable to arrange this discrepancy less than δ . So a good criterion to choose α is to decrease this parameter until ||Auα – f || ∼ δ . 2. Numerous implementations of model examples and problems showed that the choice for which

duα

(8.24) α : inf α α>0 dα

appears to be very efficient. Heuristic justification to this statement can be clarified as follows. As the parameter α decreases, while it is not too small, the solution u α weakly depends upon a small change of the value of α, so the quantity du α /dα is almost constant and the value of α duα /dα seems to be relatively large for such large α. With further decrease of α, when approaching its extremely small values, u α becomes very sensitive to small variation of α, and so duα /dα becomes very large. The optimal choice of α should thus be made for the smallest value of the product α du α /dα. It should be noted that the quantity, for which we need to estimate the norm (8.24), admits direct estimate. To clarify this idea, we demonstrate this technique on the simplest case of Eq. (8.15). Let us differentiate Eq. (8.15), where the right-hand side f δ contains some error δ , with respect to α: duα duα duα ∗ ∗ duα + uα + α = 0, A A α +α α = –αuα . (8.25) A A dα dα dα dα But from Eq. (8.15) we have –αuα = A∗ Auα – A∗ fδ ,

(8.26)

so Eq. (8.25) is:

duα A A α dα ∗

duα +α α dα

= A∗ Auα – A∗ fδ .

(8.27)

Thus, the quantity α duα /dα can be found from the same integral equation (8.15) with a certain different right-hand side.

Helpful remarks 1◦ . Approach applied for investigation of Eq. (8.19) remains very efficient even in the extremely ill-posed case when the interval of variation of the variable x ∈ (c, d) is different from (a, b) (in practice, the former is a subinterval of the latter), the case rather typical in


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many inverse problems with very limited input data. You can apply the discussed approach also in this case, which will lead to an equation which is very similar to Eq. (8.22): Z b K1 (x, ξ )u(ξ ) dξ + α q0 u(x) – q1 u00 (x) = f1 (x), a ≤ x ≤ b, a (8.28) Z d Z d K1 (x, ξ ) =

K (t, x)K (t, ξ ) dt,

c

f1 (x) =

K (t, x)f (t) dt.

c

2◦ . Operation with the Sobolev Hilbert space W21 (a, b) seems to be quite natural from the viewpoint of compactness in the Banach space C (a, b). It is absolutely clear that any set of continuous functions whose derivatives are uniformly bounded by a certain constant satisfy the Arzela–Ascoli theorem (see Section 1.5).

8.3. Iterative Methods for Operator Equations of the First Kind As follows from results of Chapter 1 (see Sections 1.6 and 1.7), generally there are no results on solvability in the classical sense of operator equations of the first kind, in contrast to the second-kind equations, where the classical Fredholm theory guarantees that in the case of regular values of the parameter λ the operator (I – λG) is invertible and so the equation u – λGu = f is solvable in this case, at least for compact operator G. Some results on solvability of the first-kind integral equations are discussed in Section 1.7, but only for convolution kernels (i.e., when the kernel depends only on the difference of its arguments) and only about a generalized rather than classical solution. Uniqueness of such equations is also a poorly studied question. The only strong (and quite clear) result here is that if A is compact in any Banach functional space, which is not finite-dimensional, then the inverse operator A–1 does not exist (i.e., it is not continuous). From this point of view, it is absolutely unclear how equation (8.1) can be solved in practice by any traditional numerical method, and a powerful indirect approach is based upon the Tikhonov’s regularization scheme discussed in the previous section. It is rather unexpected that some classical iterative techniques well known in the literature on numerical methods automatically provide convergence of iterations to an exact solution, in the case when the latter exists, perhaps not uniquely. A natural idea to construct such an approach is based on the following consideration. Let us assume that we know a certain approximation u n to the exact solution u∗ of the equation Au = f , where A : H → H is a linear continuous operator in the Hilbert space H . Let us calculate the discrepancy ϕ(u) = ||Au – f ||2 at the element un : ϕ(un ) = ||Aun – f ||2 = (Aun – f , Aun – f ),

(8.29)

and try to find a direction where this quadratic functional decreases with the maximum possible rate. This is known to be determined by the gradient to this functional (see Kantorovich and Akilov, 1982; and also Section 8.2 here). The latter can be defined from its elementary variation δϕ: δϕ(u) = δ (Au – f , Au – f ) = 2(Au – f , Aδu) = 2(A∗ (Au – f ), δu); hence, grad ϕ(u) = 2A∗ (Au – f ),

(8.30)

where A∗ designates the adjoint of the operator A. This implies that the next step of iteration can be found along the direction of the calculated gradient: un+1 = un – γ grad ϕ(un ) = un – γA∗ (Aun – f ) ∼ ξn+1 = ξn – γA∗ Aξn ,

ξn = un – u∗ , (8.31)


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where u∗ is the exact solution to the studied equation: Au∗ = f .

(8.32)

It is obvious that ξn designates the deviation of the approximate solution from the exact one. Let us calculate its norm (ξn+1 , ξn+1 ) = (ξn , ξn ) – 2γ (ξn , A∗ Aξn ) + γ 2 (A∗ Aξn , A∗ Aξn ) ∼ ||ξn+1 ||2 = ||ξn ||2 – 2γ ||Aξn ||2 + γ 2 ||A∗ Aξn ||2 ,

(8.33)

and try to achieve its minimum value, as much as possible. This is reduced to a simple minimization of the quadratic functional, which involves γ = ||Aξ n ||2/||A∗ Aξn ||2 . Therefore, the iteration scheme (8.31) has the optimal form un+1 = un –

||Aξn ||2 A∗ (Aun – f ), ||A∗ Aξn ||2

(8.34)

or equivalently, ξn+1 = ξn –

||Aξn ||2 A∗ Aξn , ||A∗ Aξn ||2

||ξn+1 ||2 = ||ξn ||2 –

||Aξn ||4 , ||A∗ Aξn ||2

(8.35)

which is called the Steepest Descent Method (SDM). The following theorem was first proved by Fridman (1962). THEOREM 1. Iterative process (8.34) ∼ (8.35) converges monotonically to the exact solution u∗ of the equation Au = f .

Proof. Monotonic behavior of the convergence follows from Eq. (8.35), which shows that the distance between the approximate and exact solutions decrease at each step of the iterative process. In order to give the proof of convergence of the process, let us recall some classical results about geometrical properties of Hilbert spaces (see, for example, Kantorovich and Akilov, 1982). We denote by {ej }, j = 1, 2, . . . , an ortho-normalized basis of the space H , constructed from eigenfunctions of the self-adjoint and positive operator A ∗ A: A∗ Aei = βi ei , where βi are its eigenvalues, which we arrange in decreasing order. Then any element x ∈ H can be expanded into a series in elements of this basis: x=

∞ X j=1

xj ej ∼ ||x|| = 2

∞ X

x2j ,

A Ax = ∗

j=1

∞ X

βj xj ej .

(8.36)

j=1

Let us rewrite Eq. (8.35) in the following form: ξn+1 = ξn –

1 ∗ A Aξn , λn

λn =

||A∗ Aξn ||2 , ||Aξn ||2

(8.37)

and note that 0 ≤ m ≤ λn ≤ ||A∗ || = ||A|| = M . For further consideration it is important whether m > 0 or m = 0. We give the proof if the former condition is valid. The latter case is proved in a similar way and can be found in Fridman (1962). So, let us assume that 0 < m ≤ λn ≤ M . First of all, we note that Eq. (8.37) implies (ξn+1 , ξn+1 ) = (ξn , ξn ) –

(Aξn , Aξn ) . λn

(8.38)


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It follows from the last relation that a) ||ξn ||2 monotonically decreases with n → ∞ (which was already outlined), and b) that ||Aξn ||2 → 0, λn

n → ∞.

(8.39)

Since λn ≤ M , this implies that ||AξnP || → 0, n → ∞. ∞ Further, from the expansion ξn = j=1 ξnj ej , ∀n = 1, 2, . . . it follows that ∞ ∞ X βj 1 X j j ξ βj ej = ξn 1 – ej ξn+1 = ξn – λn j=1 n λn j=1 n Y βj βj j j j = ξ0 1– . ∼ ξn+1 = ξn 1 – λn λi i=0

(8.40)

It follows from the last relation that ||ξn || = 2

∞ X

(ξnj )2

=

j=1

∞ X

(ξ0j )2

j=1

n–1 Y i=0

βj 1– λi

2

.

(8.41)

P∞ Since the convergence of the series j=1 βj2 implies that βj → 0 with j → ∞, so there exists the number N such that βj ≥ m for j = 1, 2, . . . , N , and βj ≤ m for ∀j ≥ N + 1. Then we decompose the sum in Eq. (8.41) to the two ones:

||ξn || = S1 + S2 , 2

S1 =

N X

(ξ0j )2

j=1

n–1 Y i=0

βj 1– λi

2

,

S2 =

∞ X

(ξ0j )2

j=N +1

n–1 Y i=0

βj 1– λi

2

, (8.42)

and consider these sums separately. We thus have S2 ≤

∞ X

(ξ0j )2 ,

(8.43)

j=N +1

since for these values of j 0≤

βj ≤ 1, λi

and consequently 0 ≤ 1 –

βj ≤ 1. λi

(8.44)

The quantity (8.43) can be made less than arbitrary small value ε/2, for sufficiently large N : S2 < ε/2. When estimating the first sum S1 , we return to the initial representation (8.41): S1 =

N X j=1

(ξnj )2

N N X 1 X 1 j 2 βj (ξn ) ≤ βj (ξnj )2 = β m j j=1 j=1

1 ε 1 = (A∗ Aξn , ξn ) = (Aξn , Aξn ) < , m m 2

(8.45)

where the last expression can be made less than the small quantity ε/2, because ||Aε n || → 0, n → ∞. The theorem is proved.


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In the case when the operator A is self-adjoint and positive, i.e., for ∀u, v ∈ H (Au, v ) = (u, Av ) and (Au, u) > 0, the result of the theorem remains valid when the iterative formula is applied in the less complex form: un+1 = un =

||Aξn ||2 Aun , (Aξn , A2 ξn )

ξ n = un – u∗ ,

Aξn = Aun – f ,

(8.46)

which is equivalent to a formulation of Kantorovich (see Kantorovich and Akilov, 1982). The most difficult problem here is to give the estimate for the rate of the convergence, i.e., to estimate what is the real behavior of ||ξn || = ||un – u∗ || as n → ∞. Generally, the only known estimate here is that 1 2 , n → ∞, (8.47) ϕ(un ) = ||Aun – f || = O n

and no known estimate for ||un – u∗ ||.

In the case of nonlinear operator equations of the first kind, the theory of regularization is not so well developed, but some interesting results can be obtained also in this case. Here our discussion follows Tikhonov and Goncharsky (1987). First of all, it is clear that if we construct an approximating sequence, then for ill-posed problems the number of iteration steps should be coupled with the in-put error, i.e., the value n(δ ) must be related with the error δ . Thus n(δ ) would act as a regularization parameter. It is clear that the solution of an operator equation like (8.1) also in the nonlinear case can be reduced to the minimization of a discrepancy functional. That is why we consider here the iterative methods simultaneously to solve the first-kind operator equations and to minimize functionals. The most essential difficulties arise because classical iteration algorithms cannot be directly applied to ill-posed problems. Let f (u) be a rigorously convex differentiable functional defined on a closed convex set U of a Hilbert space H . Then we apply the basic ideas of the method of gradient projection used above, i.e., un+1 = PU un – γn f 0 (un ) , (8.48) where PU is the projection operator on U ⊂ H , γn is a constant depending on the properties of f . In this formulation classical algorithms like (8.48) typically provide only a weak convergence of the approximating sequence un . This is true for any classical method. Note that the initial nonlinear problem can be re-formulated in the form of a variational inequality. It is known that the minimization of a convex functional on a closed convex set U ⊂ H of a Hilbert space can be reduced to the solution of the following variational inequality with respect to u: f 0 (u), (u – z ) ≤ 0, ∀z ∈ U . (8.49)

Here f 0 (u) means the usual gradient if we assume that the functional is smooth. If f (u) is a convex functional, then f 0 (u) is a monotonic operator, i.e., f 0 (u2 ) – f 0 (u1 ), (u2 – u1 ) ≥ 0, ∀u 1 , u 2 ∈ U . (8.50) Let us formulate the problem of solving the inequality F (u), (u – z ) ≤ 0, ∀z ∈ U ,

(8.51)

where F (u) is a monotonic operator. In the case U = H , the formulation includes the solutions of a nonlinear operator equation with a monotonic operator F . A standard iterative process cannot be applied, as we already noted, to solve this problem.


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Let M (u) be a strictly monotonic operator on U ⊂ H , i.e., M (u1 ) – M (u2 ), (u1 – u2 ) ≥ CM ||u1 – u2 ||2 ,

and we consider, instead of (8.51), the auxiliary inequality F (u) + αM (u), (u – z ) ≤ 0, ∀ z ∈ Q,

CM > 0.

(8.52)

α > 0.

(8.53)

Since the operator F (x) + αM (x) is strictly monotonic (for any α > 0), the inequality in (8.53) can be solved by any ordinary classical iterative method. On the other hand, the inequality (8.53) approximates (8.51) when α → 0 (which directly follows from Brauder– Tikhonov lemma). So, let us apply a standard iterative method to solve problem (8.53) for a fixed α > 0: un+1 = R(α, un , M ). (8.54) The iteration algorithm

un+1 = R(αn , un , M )

(8.55)

can solve (8.51) for a certain choice of the sequence α n → 0. In order to clarify this statement, let us apply the following theorem, which illustrates the iterative scheme. THEOREM 2. Let f (u) be a doubly differentiable functional such that the domain of definition D (f 0 ) = H , M (u) = u, and ||f 00 (u)|| ≤ N , ∀u ∈ U . Let γn > 0 and αn > 0 satisfy the following conditions: (a) lim αn = 0, (b) αn ≤ αn–1 , (c) n→∞

Then the iterative process

∞ X n=1

γn αn = +∞, (d) lim (1+ αn )γn < 2/N . (8.56) n→∞

un+1 = PU un – γn (f 0 (un ) + αn un )

(8.57)

converges from any starting point u0 ∈ H to a solution of the variational problem (8.51), which has the minimal norm in H . The proof can be found in Tikhonov and Goncharsky (1987). Until now we considered only ill-posed problems with exactly known initial data. Iterative methods such as (8.57) can also be used to solve problems with perturbed input data. This is because the constructed algorithm, with a proper choice of the number of applied iterations linked with the error δ of the input data, yields some regularizing algorithm for (8.51). The following theorem demonstrates how we can determine the applied number of iterations n as n(δ ), to be used in the iterative regularization (see Tikhonov and Goncharsky, 1987) THEOREM 3. Let us assume that the gradient of the perturbed right-hand side f˜ 0 belongs to the same class as the exact f 0 of Theorem 2, i.e., ||f 00 (u)|| ≤ N , ∀u ∈ U , and ||f 0 (u) – f˜0 (u)|| ≤ δ ||u||. Then the process defined in (8.57) generates a regularizing algorithm for (8.51) if, under the conditions of Theorem 2, n(δ ) is such that lim δ/αn(δ) = 0.

n→∞

(8.58)

The theorems we have represented illustrate the power of iterative regularization. However, the discussed ideas can also be applied to directly solve linear ill-posed operator equations such as (8.1), which can be constructed without using (8.55) or (8.57). For


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example, a solution of (8.1) with approximate data f δ and an exact operator A can be found by using the following iterative process (compare with results of Theorem 1): un+1 = un + µ(A∗ fδ – A∗ Aun ),

u0 = 0,

0 < µ < 2/||A∗ A||.

(8.59)

Let b > 1 be a fixed constant. Then the regularization parameter can be chosen according to the discrepancy principle in the following way. For n(δ ) we take the first n such that ||Aun – fδ || ≤ bδ .

(8.60)

The proof to the following theorem can be found in Tikhonov and Goncharsky (1987). THEOREM 4. Let A be a continuous linear operator from a Hilbert space U to a Hilbert space F , with the iteration sequence un defined by (8.59). If n(δ ) is chosen as given in (8.60), then ||un(δ) – u∗ || → 0 with δ → 0, where u∗ is the normal solution to Eq. (8.1) for δ = 0, i.e., the solution with a minimal norm in U .

Helpful remarks 1◦ . The discussed iterative method given by Theorem 1 represents an approach, which is an alternative to Tikhonov’s smoothing functional construction. A role analogous to the small parameter of regularization 0 < α 1 here is played by the number of iteration n 1. In order to justify this statement, let us suppose that u 0 is a solution to the main equation: Au0 = f , and fδ is an approximate (i.e., perturbed) right-hand side with ||f – f δ || < δ , where δ > 0 is a small quantity. Then the explicit form of the SDM, Eq. (8.34) ∼ (8.35), which expresses the (n + 1)st iteration through the previous nth iteration by a finite combination of the direct operator A, shows that for a finite number of the iteration n the difference ||uδn – un || can be made arbitrarily small when ||f – f ||δ → 0, i.e., the iterative process with finite number n guarantees stability with perturbations of the right-hand side f . 2◦ . Between Tikhonov’s and iterative schemes there is a significant difference. If we need to solve the first-kind operator equation (8.1) with exact (i.e., unperturbed) right-hand side f , then the iterative SDM can give us an approximation to the exact solution with arbitrary accuracy, by taking more and more iterations. In contrast with this, Tikhonov’s scheme based on the help of a small smoothing parameter cannot provide the approximation to the exact solution with arbitrarily small error, since the regularized equation (A ∗ A + αI )u = f cannot be solved for too small α.

8.4. Comparison of Various Methods for Reconstruction of the Scatterer Geometry Over the last years, increasing interest has been shown in inverse scattering problems. A good survey of the state of the art can be found in Colton and Kress (1992), where mainly rigorous and abstract mathematical theory is developed. In practical aspect, this interest can be explained by applications to radio-location, ocean acoustics, medical and technical ultrasonics, seismic problems, etc. The current section presents an analysis of the main approaches to the problem of reconstruction of obstacle shape. For all that, we consider only the papers where: 1) the authors do not use any simplified theories (like Kirchhoff’s hypothesis) that were a subject of the previous Chapter 7; 2) some examples of reconstruction are presented; 3) proposed algorithms are sufficiently wide to be applicable to more complex problems. It should also be noted that we are only interested in those algorithms which can directly be applied in practice.


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We start from the remark that the studied problems are ill-posed and nonlinear simultaneously. Apparently, the approach proposed by Imbriale and Mittra (1970) was the first one where the authors did not accept any simplified assumptions. If we solve the 2D Helmholtz equation in the exterior of a certain acoustically soft obstacle, then the authors’ principal idea is as follows. Let the scattered pattern F (ϕ) be known in the far zone. Then the Fourier coefficients am in the series expansion of this function ∞ X F (ϕ) = am eimϕ (8.61) m=–∞

uniquely determine the scattered wave field in the polar coordinate system in the following form: ∞ X (1) psc (ρ, ϕ) = am im Hm (kρ) eimϕ , (8.62) m=–∞

where the series converges at least outside a minimum disk C 0 containing the unknown obstacle. Since the full pressure is trivial on the boundary contour l (due to the boundary conditions), so with decreasing ρ the first point for which p inc + psc = 0 is a point belonging to the boundary of the object. If we use different origins of coordinate systems, then in principle we can restore a set of boundary points. This method can be easily spread to the case of acoustically hard body, since it is obvious that for smooth boundary curves ∂p/∂n = ∂p/∂ρ. Despite its simplicity and attractiveness, this method possesses at least the following two disadvantages: 1) This cannot be applied directly to reconstruction of nonconvex objects. In order to overcome this restriction, the authors apply a method of sequential continuation of the scattered wave field. Such an approach, which is contiguous to analytic continuation in the theory of complex-valued analytic functions, consists in multiple re-expansion of the Fourier series in wider and wider domains. Just as the problem of analytic continuation, this problem is one of the most difficult ill-posed problems. 2) Convergence of the series (8.62) in the exterior of the disk C 0 does not mean that it is so easy to calculate its sum in practice. In order to realize in practice the discussed method, we need to calculate the sum (8.62) for points very close to the boundary of its convergence, and it is not clear what is the rate of its convergence in such cases. This question is closely connected with the classical Rayleigh hypothesis if it is possible to continue the series (8.62) inside C0 , and if so, for how long of a distance. For instance, there is a known result (see also Section 8.7) that for elliptic obstacle in the coordinate system, related with its axes, the boundary of the convergence of the series (8.62) passes its foci, situated on the principal axis. Therefore, for elongated ellipses the point at which pinc + psc = 0 is situated very close to the boundary of convergence. However, this feature is a common disadvantage of almost all existing methods of reconstruction: they lose their efficiency for narrow obstacles. In Colton and Monk (1985) the authors propose an interesting approach, which is based on the concept of a Herglotz function. A Herglotz function is an entire function (in the sense of complex-valued analysis) that satisfies a Helmholtz equation. Then the authors introduce a certain nonnegative functional that vanishes at the actual solution of the problem. Since the considered inverse problem is ill-posed, the proposed numerical algorithm consists in the minimization of the functional on some compact set. In order to apply the proposed algorithm, one needs to know the scattered far-field amplitude on some interval of the wave number, k ∈ (k1 , k2 ), which is too problematic in practice. In the next works of the authors (see Colton and Monk, 1986) this restriction was overcome.


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In Kirsch et al. (1988) there is proposed a method where the scattered wave field is sought as a single-layer potential (cf. Section 2.1) over some auxiliary contour (compare with the auxiliary sources method in Section 8.7): Z sc p (x) = H0(1) (k |x – y |)ψ (y ) dsy , (8.63) C

over a certain contour C situated inside the unknown obstacle, whose boundary is under reconstruction. Equation (8.63) implies the following expression for the far-field scattered amplitude: Z e–ik(y1 cos ϕ+y2 sin ϕ) ψ (y ) dsy ,

F (ϕ) =

0 ≤ ϕ ≤ 2π .

(8.64)

C

If the scattered diagram in the far zone F (ϕ) is known, then Eq. (8.64) is a linear integral equation of the first kind (which is ill-posed) with respect to the function ψ (y ), y ∈ C . The second step of the proposed method consists in the determination of the sought boundary contour l for which pinc + psc = 0, where psc is defined by Eq. (8.63). It is clear that efficiency of such an approach strongly depends upon the choice of the contour C . Completing the survey of the first three methods let us notice that they all possess the essential restriction: since they use in one or another way analytical representation of the wave field these approaches cannot be directly spread to the case when we know not the full far-field wave amplitude F (ϕ) but only its modulus |F (ϕ)|. In the meantime, the problems with given F (ϕ) are of most importance in practice. So, let us proceed to discussion of methods free of the marked disadvantage. The method proposed by Angell with co-authors (1989a,b) was tested on some examples. The wave field outside the sought domain is approximated by a finite number of the elementary waves sc

p (ρ, ϕ) =

M X

cm i

m

(1) Hm (kρ)

m=–M

cos(mϕ) . sin(mϕ)

(8.65)

The chosen approximate representation of the solution a priori satisfies the Helmholtz equation and radiation condition. The only remaining conditions to be satisfied are the boundary condition over the contour l: (pinc + psc )|l = 0, and the requirement that the scattered wave field must coincide with the known function F (ϕ): F (ϕ) =

M X

m=–M

cm i

–m

cos(mϕ) , sin(mϕ)

0 ≤ ϕ ≤ 2π .

Further, by introducing the two discrepancy functionals 2 Z 2π M X cos( mϕ ) Ψ1 = cm i–m F (ϕ) – dϕ, sin( mϕ ) 0 m=–M 2 Z X M cos( mϕ ) (1) Ψ2 = cm im Hm (kρ) + pinc dϕ, sin(mϕ) l m=–M

(8.66)

(8.67)

the problem is reduced to a minimization of the functional Ψ = Ψ 1 + σΨ2 with a certain positive parameter σ > 0. Since the studied problem is ill-posed, the minimum of Ψ is being sought on a compact set.


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By the first view this approach seems to be very similar to the method of Imbriale and Mittra (1970). However, here the question of convergence of any series is of no significance. The difference is quite the same as the difference between the question about convergence P∞ of the power expansion of any function f (x) = m=0 am xm on the interval (0, L) and the question about approximation of this function on the same interval by a polynomial: PM f (x) ≈ m=0 am xm . It should be noted that the discussed method has the evident merit, because this does not require under the minimization of the functional Ψ multiple repeated solving of the direct scattering problem. It is interesting to note that this merit of the method is at the same time its disadvantage for the following reason. Let for a problem, where the unknown object is irradiated by an incident wave, there be chosen M terms in Eq. (8.65) and equal number of terms in expansion of the contour ρ(ϕ). Then the dimension of the corresponding linear algebraic system, which is solved at each iteration step by the Levenberg–Marquard (see, for example, Gill et al., 1981) method, is 2M . If there are N incident waves in the problem then the total number of coefficients cm is equal to M N , and dimension of the corresponding system is (N = 1) × M . For example, for M = 5 with one incident wave there is a need to solve a 10 × 10 system, while for 24 incident waves (a circular scanning with a 15◦ step), a 125 × 125 system (compare with the method discussed below in Section 8.4). At the same time, a direct problem would need in this situation to solve a 5 × 5 matrix system with 24 right-hand sides, which is much easier. We should agree that in an arbitrary nonlinear problem the most natural way is to apply the Newton method. This approach was realized by Roger (1981), and chronologically that was the second published work, after Imbriale and Mittra (1970), where the author quoted examples on reconstruction of the unknown boundary. In some more detail, the method proposed by Roger is as follows. For a known far-field scattered diagram F (ϕ), in the case of acoustically soft obstacle, the author solves the system of two integral equations Z 2π F (ϕ) = g (θ )e–ikρ(θ) cos(ϕ–θ) dθ , 0 ≤ ϕ ≤ 2π , 0 Z 2π (8.68) g (θ )H0(1) [kr (ϕ, θ )] dθ = eikρ(ϕ) cos ϕ , 0 ≤ ϕ ≤ 2π , 0

1/2 r (ϕ, θ ) = ρ2 (ϕ) + ρ2 (θ ) – 2ρ(ϕ)ρ(θ ) cos(ϕ – θ ) ,

with respect to the two unknown functions — the function g (ϕ), related to the normal component of the velocity over the boundary contour, and the function ρ(ϕ), which describes the boundary of the unknown star-like domain. Here the incident wave is assumed to be plane. For numerical implementation there is used a step-by-step method where at each step a linearized system is being solved. These linear systems, similarly to the full problem, are also ill-posed, and so are studied with the help of Tikhonov’s regularization. Such an approach is very fast in the sense of computation time. However, this algorithm possesses an intrinsic disadvantage—the question on convergence of the Newton method in ill-posed problems is poorly studied in the literature. In particular, it is not so evident what is the rate of convergence of this method. In this connection we only note here that all known estimates about convergence of the Newton method applied to nonlinear systems operate with the norm of the operator inverse to the Jacobian of the considered operator equation. However, in the ill-posed problems Jacobian is irreversible. In the work of Tobocman (1989) the Newton method is extended to the case of acoustically hard boundary. It is also admitted irradiation of the obstacle by several sources with different frequencies. The basic distinction from the method of Roger is that for


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solving linear systems at each iteration step the author first applies, instead of Tikhonov’s regularization, a singular decomposition of the corresponding matrices and then neglects its small eigenvalues. Kristensson and Vogel (1986) apply a quasi-Newton method to minimization of a functional that corresponds to Tikhonov’s smoothing functional under the process of discrepancy minimization for the initial nonlinear system. For all that, as an operator of the corresponding direct problem, the authors use the so-called T-matrix approach (see Boström, 1986). It should be noted that a good survey of various methods to solve inverse problems in diffraction theory can be found in Burov (1986). Besides, some interesting approaches to inverse problems of diffraction were also proposed in Colton (1984),

Helpful remarks 1◦ . Let us emphasize the two restrictions inherent in all above discussed approaches: 1) Their possibilities, as a rule, are demonstrated in practice only for the case of a 2D problem. The passage from the model 2D to a real 3D case increases the number of unknowns at least by an order of magnitude, which leads to a significant increase in the computational time. 2) They are efficient only for long or moderate waves. The passage to the real frequency range, where representative size of obstacles is very often by an order larger when compared with the wavelength, implies significant increase in the number of nodes on the grid. Therefore, with the frequency increasing, the dimension of the corresponding finitedimensional linear (or nonlinear) algebraic systems becomes too large even for modern computers. If we take into account both remarks 1) and 2), we can conclude that for real practical requirements the total number of the grid nodes increases at least by a factor of M = 50 to 100. Since all the above discussed methods reduce the studied inverse diffraction problem to some (linear or nonlinear) system of algebraic equations, this involves an increase in the computation time by a factor of M 3 ∼ 105 to 106 .

2◦ . Another important disadvantage inherent in all discussed methods (perhaps, excluding the work of Imbriale and Mittra, 1970) is the fact that they restore the boundary of only star-like domains. Such curves can be described by a one-to-one function ρ = ρ(ϕ) in the polar coordinate system. Operation with arbitrary domains is impossible with such approaches, since it is not clear a priori what are the intervals of the polar angle variance for which the function ρ = ρ(ϕ) is not single-valued. In order to overcome this difficulty absolutely different approaches should apparently be developed.

8.5. General Inverse Diffraction Problem: Combination of Iterations and Smoothing As we could see from the previous Chapter 7, the inverse problem on the reconstruction of the shape of unknown obstacles from the known scattering diagram within the framework of geometrical diffraction theory can be reduced to a Minkowski problem, which is quite classical in differential geometry. At least for the class of convex obstacles this can be numerically studied as a well-posed boundary value problem since this is reduced to a second-order partial differential equation with respect to the support function, the problem known to be very stable with respect to small perturbations of the given data. In contrast with this, the considered inverse problem in exact formulation is reduced (see below) to a system of nonlinear integral equations, with the principal of them being a


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Fredholm equation of the first kind with very smooth (and hence “very bad”) kernel. The first three sections of the present chapter dealt with such a type of ill-posed problems, and here we will demonstrate in which way some techniques discussed above can be applied to the studied inverse problem on reconstruction of obstacles’ shape in exact formulation. Let us note once again that inverse problems of mathematical physics in general were studied by many authors (a good state of the art, with applications to inverse diffraction problems, is presented in the monograph of Colton and Kress, 1992). Let a certain specified wave p0 be incident on a body situated in an acoustic medium. We will restrict ourselves by the two-dimensional case for the sake of simplicity. As usual, the incident wave is assumed to be plane and for simplicity propagating along the horizontal axis, so that p0 = eikx (k is the wave number). We limit ourselves to the class of star-like regions. Then a parametric representation in polar coordinates, ρ = ρ(φ), 0 ≤ φ < 2π , is possible for the contour l, bounds the body under consideration. To be more specific, we assume that contour l is acoustically soft. In this case, the boundary condition has the form p|l = 0. The problem is posed R as follows. The field scattered from the body at all angles ϕ, i.e., the function F (ϕ) = l exp[–ikρ(θ ) cos(ϕ – θ )]∂p/∂n|l dl, 0 ≤ ϕ < 2π is known (see Chapter 2 for an expression of the scattered wave field). It is required to determine the function ρ(ϕ). We will assume that contour l is smooth. Then we can apply the equality dl = (ρ2 + ρ02 )1/2 dθ to introduce the function g (θ ) = [ρ2 (θ ) + ρ02 ]1/2 dp(θ )/∂n|l , in terms of which the scattered wave field is expressed in the form Z 2π F (ϕ) = exp [–ikρ(θ ) cos(ϕ – θ )] g (θ ) dθ , 0 ≤ ϕ < 2π . (8.69) 0

There are two unknown functions in this problem: g (θ ) and ρ(θ ). It is therefore necessary to attach another relation to (8.69). For example, it might be the boundary integral equation on contour l (see Chapter 2): Z 2π g (θ )H0 [kr (ϕ, θ )] dθ = p0 |l = exp [ikρ(ϕ) cos ϕ], 0 ≤ ϕ < 2π , (8.70) 0 2 2 1/2 r (ϕ, θ ) = [ρ (ϕ) + ρ (θ ) – 2ρ(ϕ)ρ(θ ) cos(ϕ – θ )] .

Thus, it is possible to reduce the problem to a system of two nonlinear integral equations (8.69), (8.70) for the two unknown functions g (θ ) and ρ(θ ). Note that the problem posed is the simplest inverse diffraction problem. In more complex cases, it is possible to specify not F (φ) but only its absolute value |F (φ)|, and not for all 0 ≤ φ < 2π . Further, the direction of the incident field may be unfixed. It should be noted once again that the problem under consideration is ill-posed in the sense of Tikhonov. This is because the operator of the direct problem G : ρ(φ) → F (φ) is compact in the spaces naturally associated with the problem. Then it becomes obvious, as we already remarked above, that the inverse operator G –1 : F (φ) → ρ(φ) cannot be continuous. In addition, the problem is nonlinear. All these features complicate its solution since the most efficient numerical methods for ill-posed problems are developed for linear problems. One of the most universal numerical methods is the gradient descent method (see the previous sections). It has been shown in Section 8.3 that in the linear case (i.e., when the discrepancy functional is quadratic) this method yields a regularizing algorithm. Although there is no rigorous proof for the general case, it turns out that this property of the method also can be extended to nonlinear problems. This follows from the fact that any smooth functional is quadratic in the neighborhood of a local minimum. The numerical solution algorithm is constructed as follows. First Eqs. (8.69) and (8.70) are written in a finite dimensional form at the nodes φ i , θj (φi = θj ) using a simplest


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rectangle quadrature formula. The collocation nodes form a uniform grid on the segment (0, 2π ). The diagonal terms in (8.70) (at i = j ) must be evaluated as the integral over a small interval of a principal singular part of the kernel (the singularity is logarithmic). With a sufficiently dense grid of nodes, the nonlinear finite-dimensional system obtained in this way fi (g1 , . . . , gj , ρ1 , . . . , ρj ) = 0, i = 1, . . . , 2J , (8.71) has operator properties similar to those of the original fully continuous operator. For example, some of the eigenvalues of its Jacobian are very close to zero. In constructing the numerical algorithm, therefore, it is necessary to remember that the problem (8.71) is in practice ill-posed. PJ We write the discrepancy functional ψ (g , ρ) = i=1 |fi |2 . At the zero-th step, we find min ψ (g , ρ) on all possible circles. Suppose that the values g 0 , ρ0 correspond to a circle found in this way. The next step is to apply an iterative gradient-descent procedure. Here (see, for example, Gill et al., 1981) grad ψ = A∗ f

(8.72)

where A is the Jacobian of system (8.71), f is the vector of its left-hand side, and the asterisk is the symbol of the conjugate. The Jacobian A can be calculated in explicit form in the present problem. Thus, the direction of descent is given by the equality q = –A∗ f ,

q = (∆g , ∆ρ)T .

(8.73)

It is found that this method makes it possible to solve certain simple problems. For example, it is possible to obtain an ellipse of small eccentricity from a circle. However, the method converges very slowly for more complex problems, hence the convergence should be acceleration by any reasonable procedure. To provide such an acceleration, we note that any direction g = –B grad ψ will be a direction of descent if B is a positive definite matrix. A suitable matrix B must be chosen to accelerate the procedure. For this purpose, the functional ψ is approximated quadratically at the point z = (g , ρ) (see Gill et al., 1981): ψ (z + q ) = ψ (z )+(A ∗ f , q )+ 21 (q , Gq ). Minimization of ψ (z + q ) generates the direction (8.73) in linear approximation. Minimization of ψ (z + q ) in the quadratic approximation results in the relation Gq = –A ∗ f . It is shown in Gill et al. (1981) that the Hessian G ≈ A∗ A in the neighborhood of its minimum, so this relation is approximately A∗ Aq = –A∗ f . With (8.72), we might have taken (A∗ A)–1 as the matrix B . However, since the problem is ill-conditioned, the operator A ∗ A has no inverse (in the sense specified above), and for this reason we use an idea implicitly related to Tikhonov’s regularization. To the operator A∗ A we add a small positive operator, for example, αI , α > 0. We then obtain (A∗ A + αI )q = –A∗ f . (8.74) This relation determines a certain direction of descent. Thus, B = (A ∗ A + αI )–1 is correctly defined. It is obvious that (8.74) becomes (8.73) for α 1. It turns out that implementation of the above algorithm faces a difficulty, because it is not so easy to find a universal criterion for selection of the parameter α, so that this will be suitable for various classes of contours l. By our experience, the following approach turns out very efficient. Several values αm (m = 1, . . . , M ) that form a geometric progression (usually M = 5 to 8) are chosen in the range αmin = 10–4 < α < 101 , and the values αopt at which the functional ψ decreases most sharply on the corresponding direction q opt is identified among them. Naturally, the optimum α opt determined in this way will be different at different steps of the iterative procedure. We find that α opt takes large values in the first few steps and then gradually decreases with iterations. The limit α > α min , which


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is accepted here, preserves well-posedness for the evaluation of the matrix (A ∗ A + αI )–1 . This universal method of selection of the parameter α was used in all of the following examples of reconstruction of the boundary contour l. It should be noted that system (8.71) is overdetermined. Actually, it contains 3J real unknowns: J values of {Re gj , Im gj , ρj }. At the same time, the system has 4J real equations. The proposed approach permits solution of overdetermined systems. For example, the numbers of unknowns and equations would be the same if only the real amplitude |F (φ)| of the wave was known. Some examples of reconstruction for certain types of boundary contour l by the method described in this section can be found in the author’s work (Vorovich and Sumbatyan, 1990). Let a body with an unknown boundary l be irradiated successively by plane waves at various angles γn , n = 1, . . . , N : pn0 = exp[–ikρ(ϕ) cos(ϕ – γn )]. Suppose that the scatteringwave amplitude is known for several angles ϕni , i = 1, . . . , I for each nth incident wave. Then the problem in finite-dimensional form reduces to the nonlinear system 2 J X n n n fi = h exp[–ikρi cos(ϕi – θj )] gj – Fin = 0, (8.75) j=1 h

J X

H0 [kr (ρi , ρj )] gin = exp[–ikρi cos(ϕi – γn )]

(h = 2π/J ),

(8.76)

j=1

where we should remember that the term i = j in the second equality requires a careful treatment. System (8.75)–(8.76) can be solved by the method described above. Then the linear system (8.74) for determination of the descent direction is a system of (2N + 1)J real equations for the same number of real unknowns Re gjn , Im gjn , ρj . This approach requires a great deal of machine time when J and N are large. To decrease the dimension we make use of two properties of system (8.76), which we treat as a system for g jn : it is linear with respect to gjn ; only the right-hand side changes in this system when n changes. We therefore propose to reduce the problem to minimization of the functional ψ=

N X I X

|fin |2

(8.77)

n=1 i=1

with condition (8.76), i.e., to a problem of conditional optimization with constraints in the form of equalities (see Gill et al., 1981). As above, we solve this problem by quadratic approximation of functional (8.77) with respect to the increment vector: q = (∆g jn , ∆ρj )T = (xV , xU )T . Differentiation of (8.76) gives the relation ! v ... 0 VxV + UxU = 0, V = ... v ... . (8.78) 0 ... v Here the matrix V consists of N identical matrices v = (v ij ), vij = hH0 [kr (ρ1 , ρj )] situated on the diagonal of the matrix V . The matrix U can also be written in explicit but more complex form. We note that the matrix V corresponds to the direct diffraction problem, has a dominant diagonal by virtue of the logarithmic singularity of the kernel, and is therefore properly invertible.


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The conditional minimization of the quadratic functional with constraints in the form (8.78) gives the equation (see Gill et al., 1981): (Z ∗ A∗ AZ + αI )¯q = –Z ∗ A∗ f , (8.79) where the direction of descent is determined from the formula q = Z q¯ . Here the matrix Z has the form –1 –V U Z= , (8.80) I the vector f = (fin ), and A is the Jacobian of system (8.75). The optimum value of the parameter α is chosen at each step in the same way as in the case of unconditional minimization. The dimension of system (8.79) is equal to J . Figure 8.1 demonstrates some examples on reconstruction in the case when there is known the back-scattered amplitude for N = 24 irradiation directions uniformly distributed in the range of angles (0, 2π ). For all that, I = 1, φ n = γn , and J = 24. Solid lines designate actual boundary contour, and dashed ones are calculated on results of the proposed reconstruction method.

Figure 8.1. Reconstruction of the elliptic and three-leaf obstacles in acoustic medium A similar approach can be developed in the case of reconstruction of a void of unknown shape in elastic medium. Let in elastic plane (2D case) there be situated a void with unknown boundary l, which is free of load. The given incident wave, which to be more specific we assume to be plane, falls to this void from right to left, parallel to the x 1 axis: ◦ σ11 = σ0 exp(–kp x1 ), where kp , ks are wave numbers related to longitudinal and transverse waves, respectively. For simplicity, we assume that the contour l is star-like, so that it can be represented in the polar coordinate system by a single-valued function ρ = ρ(ϕ), 0 ≤ ϕ < 2π . The most important problem here is to reconstruct the shape of the void, i.e., the function ρ(ϕ) from the amplitude of the far-field scattered diagram. The direct problem here is to construct this diagram if the shape of the obstacle is known. It can be shown that this diagram is represented by the following integral: √ R σrr (R, ϕ) ∼ F (ϕ), R → ∞, Z 2π F (ϕ) = e–ikp ρ(θ) cos(θ–ϕ) u1 (θ )[2µ γ (θ , ϕ) cos ϕ + λ n1 (θ )] (8.81) 0 + u2 (θ )[2µ γ (θ , ϕ) sin ϕ + λ n2 (θ )] dθ , γ (θ , ϕ) = n1 (θ ) cos ϕ + n2 (θ ) sin ϕ, n1 (θ ) = ρ(θ ) cos θ + ρ0 (θ ) sin θ , n2 (θ ) = ρ(θ ) sin θ – ρ0 (θ ) cos θ .


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In the considered inverse problem function F (ϕ) is supposed to be known. The unknown quantities in Eq. (8.81) are (as in the above scalar acoustic problem) ρ(ϕ), which describes the boundary contour l, and the functions u1 , u2 which designate the components of the displacement vector over the boundary l in the Cartesian coordinate system x 1 , x2 . In order to write out a complete system of equations, we need to add to equations (8.81), for example, a system of boundary integral equations (BIE) on the contour l. In this section we use, both in scalar and elastic problems, the direct BIE method (cf. Section 2.2), where the mentioned system is Z u1 (x) – 2 [P1(1) (y , x)u1 (y ) + P2(1) (y , x)u2 (y )]dly = 2u01 (x), Zl (8.82) (2) (2) 0 u2 (x) – 2 [P1 (y , x)u1 (y ) + P2 (y , x)u2 (y )]dly = 2u2 (x), l

with ∂ [U (k) (y , x)] + λny ⋅ divy [U (k) (y , x)] + µ{ny × roty [U (k) (y , x)]}, ∂ny ∂2 i (1) (1) (1) 2 k δ H [ H ( k r ) – ( k r ) – H ( k r )] , Uj(k) (y , x) = p p s 0 s kj 0 4µks2 ∂yk ∂yj 0 p r = ρ2 (ϕ) + ρ2 (θ ) – 2ρ(ϕ)ρ(θ ) cos(θ – ϕ), x = x(ϕ), y = y (θ ), x ∈ l. P (k) = 2µ

(8.83)

Here, as usual, H0(1) is the Hankel function, and (u01 , u02 ) are the components of the displacement vector in the incident wave: u01 =

σ0 e–ikp x1 , (λ + 2µ) ikp

u02 = 0.

(8.84)

Therefore, the problem is reduced to a system of the three nonlinear integral equations (8.81)–(8.82) with respect to the three unknown functions ρ, u 1 , u2 . Since relation (8.81) for the scattered wave field is an integral equation of the first kind with a smooth kernel (which yields a compact integral operator of the first kind), the described problem, similarly to the scalar case, is really an ill-posed problem. That is why this requires a regularization. Here we use the same approach as that developed for the above scalar case. The algorithm consists again of minimization of the discrepancy functional for equations (8.81), (8.82) in the functional space L 2 (0, 2π ). The minimization is carried out again by an iterative steepest descent method, with the help of its quadratic approximation at each step. Recall that this method does coincide neither with the Lewenberg–Marquardt method nor with Tikhonov regularization scheme (see the previous sections), nor with any other known technique. When performing numerical realization, we may write equations (8.81), (8.82) in a finite-dimensional form, which contains the values of the unknown functions ρi , u1i , u2i , ϕi , θj (ϕi = θi ), i, j = 1, 2, . . . , N , at the nodes of a finite mesh. The principal difficulty is related to the calculation of the Jacobian of the obtained nonlinear finite-dimensional system. Expression for the Jacobian can be written in an explicit form, which reduces the time of computations by around an order compared with its calculation by using a finite-difference scheme. Figures 8.2–8.5 demonstrate examples on reconstruction of various flaws in elastic medium. Here we used cp /cs = 5.85/3.23, which corresponds more or less to a steel medium. The results are shown for the total number of nodes N = 72 and for the number of iterations near 8–10.


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x2 b a x1

O

Figure 8.2. Reconstruction of elliptic flaw with semi-axes ak p = 1.6, a/b = 3 in elastic medium

x2

x1

O -0.6

1

1.8

Figure 8.3. Reconstruction of the flaw kp ρ(ϕ) = 0.6 (2 + cos 3ϕ) in elastic medium

b

x2

a

-a

O

x1

-b Figure 8.4. Reconstruction of the vase-type flaw with ak p = 0.9, bkp = 1.5 in elastic medium


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x2 1

-1

1

x1

O

-1 Figure 8.5. Reconstruction of the mask-type flaw in elastic medium: k p x1 (t) = cos t + 0.65 (cos 2t – 1); kp x2 (t) = 1.5 sin t; 0 ≤ t ≤ 2π

Helpful remarks 1◦ . In the case of acoustically hard contour, when the boundary condition has the form ∂p/∂n|l = 0, the problem reduces to a system of two nonlinear integro-differential equations. However, this is not an obstacle to the application of the method discussed below also to this case. 2◦ . The methods proposed here differ from other known methods: the Gauss–Newton and Lowenberg–Marquardt methods (see Gill et al., 1981), the smoothing functional method (Tikhonov and Arsenin, 1977; and Section 8.2), and others. In the Lowenberg–Marquardt procedure, for example, a unit step along the direction of descent is always fixed. And in the smoothing functional method, α cannot depend on the iteration number (if an iterative procedure is used to minimize the smoothing functional). 3◦ . The method discussed in the present section at first sight seems to coincide with the method proposed by Roger (1981) (see our brief survey in Section 8.4), but this is not so. Roger uses the Newton–Kantorovich method to solve a nonlinear functional equation which, strictly speaking, does not converge in the case of the first-kind operator equation with compact operators (both linear and nonlinear). In contrast with this, our approach is based on the steepest descent method whose convergence is strictly proved (see Section 8.3), at least for quadratic discrepancy functionals, corresponding to linear operator equations. Locally each functional is quadratic, so although we cannot strictly prove the convergence of our method in the nonlinear case, there is a good chance, starting from the proved convergence for quadratic functional, that the proposed algorithm converges also in the more general nonlinear case considered here.

8.6. A Correct Treatment of Ill-Posed Boundary Equations in Acoustics of Closed Regions Let D be a simply connected two-dimensional finite domain with boundary l occupied by a linear isotropic acoustic medium. Then the wave pressure p(x 1 , x2 ) satisfies the Helmholtz equation ∆p + k 2 p = 0, (8.85) where k = ω/c is the wave number, ω is the circular frequency, and c is the wave speed.


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The wave process in D is determined by the boundary condition set on l, which may be a Dirichlet condition (p is given on l ), a Neumann condition (the normal derivative ∂p/∂n is given on l), or a mixed condition (p is given on a part of l; ∂p/∂n, on the rest of it). We apply the conventional procedure based on the application of the Kirchhoff– Helmholtz integral formula to Eq. (8.85). An auxiliary system of elementary solutions to the Helmholtz equation (8.85) is chosen as the set of plane waves exp[ik (η , x)], where η12 + η22 = 1, x = (x1 , x2 ), η = (η1 , η2 ), (η , x) = η1 , x1 + η2 x2 , instead of the Green’s function traditionally applied in the classical approach. In general, plane waves of any type (both homogeneous and inhomogeneous) are suitable for our analysis, but we restrict ourselves to homogeneous waves with real η1 and η2 , which propagate in the plane (x1 , x2 ) at arbitrary angles without damping. We integrate Eq. (8.85) times the chosen elementary solution over D , and apply the conventional procedure based on Green’s integral formula to obtain the integral relation Z ∂ p(x) – ik (η , n)p(x) exp [ik (η , x) dlx = 0, |η | = 1, (8.86) ∂n l where n = (n1 , n2 ) is the outward unit normal to l at x ∈ l. Subject to any boundary condition, Eq. (8.86) is an ill-posed integral equation of the first kind, since its kernel is dominated by an exponential factor and is an infinitely differentiable function. Another characteristic feature of this equation is that the inner variable runs over l, while the outer one runs over the unit circle η = 1, and these variables are defined in absolutely different geometrical domains. Let us reduce the basic equation (8.86) to an infinite algebraic system. In designing a numerical algorithm, we restrict ourselves to the class of star-shaped domains D with smooth boundaries. Then, we have the following unique representation in polar coordinates: x1 = ρ(ψ ) cos ψ , x2 = ρ(ψ ) sin ψ , 0 ≤ ψ < 2π , η1 = cos ϑ, η2 = sin ϑ, 0 ≤ ϑ < 2π ,

(8.87)

where ρ(ψ ) > 0. To be more specific, we consider the Dirichlet boundary value problem, with p(x) given on the boundary: p(x)|l = f (x). By using the elementary relations (η , n) = [p(ψ ) cos(ψ – ϑ) + ρ0 (ψ ) sin(ψ – ϑ)]/[ρ02 (ψ ) + ρ2 (ψ )]1/2 , (η , x) = ρ(ψ ) cos(ψ – ϑ),

dlx = [ρ02 (ψ ) + ρ2 (ψ )]1/2 dψ ,

(8.88)

the basic equation (8.86) is transformed into the following one for the unknown function g (ψ ) = (ρ02 + ρ2 )1/2 (∂p/∂n): Z 2π g (ψ ) exp[ikρ(ψ ) cos(ψ – ϑ)] dψ 0 Z 2π (8.89) = ik f (ψ )[ρ(ψ ) cos(ψ – ϑ) + ρ0 (ψ ) sin(ψ – ϑ)] 0

× exp[ikρ(ψ ) cos(ψ – ϑ)] dψ ,

0 ≤ ϑ < 2π .

Strictly speaking, this equation is ill-posed, because, in contrast to classical boundary equations based on the conventional application of Green’s formula, its kernel does not possess even a weak logarithmic singularity that would ensure the stability of the collocation method (Voronin and Tsetsokho, 1981). However, the right-hand side of Eq. (8.89) has a very special form resulting from the application of a completely continuous operator to a regular function and, hence, has the same degree of smoothness as the operator function on


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its left-hand side. A natural question arises: is it possible to construct a stable numerical method for Eq. (8.89) that would take into account this property of the right-hand side? The answer to this question is rather difficult. Collocation-type methods are unlikely to be stable. For this reason, we develop a special projection method based on the following idea. For a circular contour l, the Fourier series expansion of the unknown function g (ψ ) in Eq. (8.89) yields a known exact solution, which can easily be verified directly from (8.89). Let us represent g (ψ ), in the case of arbitrary contour l, also as a Fourier series in ψ ∈ [0, 2π ). For simplicity, we assume that the problem is symmetric about the x 1 -axis, which does not lead to a loss of generality. Then, ρ(ψ ), f (ψ ), and g (ψ ) are even functions of ψ , consequently, ∞ X g (ψ ) = gm cos(mψ ). (8.90) m=0

We substitute (8.90) into Eq. (8.89) and calculate the scalar product of the resulting relation with cos(nϑ) (n = 0, 1, . . . ), in order to obtain the infinite algebraic system ∞ X m=0

bnm gm = hn , Z

bnm = 2 Z hn = k

n = 0, 1, . . . ,

2π

0 2π

0 0

Jn [kρ(ψ )] cos(nψ ) cos(mψ ) dψ ,

(8.91)

f (ψ )hρ(ψ ) cos(nψ ){Jn–1 [kρ(ψ )] – Jn+1 [kρ(ψ )]}

+ ρ (ψ ) sin(nψ ){Jn–1 [kρ(ψ )] + Jn+1 [kρ(ψ )]}i dψ .

To derive the last equation, we have used here the following tabulated integrals: Z 2π exp [ikρ(ψ ) cos(ϑ – ψ ) cos(nϑ) dϑ = 2πi cos(nψ )Jn [kρ(ψ )], 0 Z 2π exp [ikρ(ψ ) cos(ϑ – ψ ) cos(ϑ – ψ ) cos(nϑ) dϑ 0

Z

= πin+1 cos(nψ ){Jn+1 [kρ(ψ )] – Jn–1 [kρ(ψ )]},

(8.92)

2π

exp [ikρ(ψ ) cos(ϑ – ψ ) sin(ϑ – ψ ) cos(nϑ) dϑ

0

= πin+1 sin(nψ ){Jn+1 [kρ(ψ )] + Jn–1 [kρ(ψ )]},

where Jn is the Bessel function of the first kind of order n. System (8.91) can be rewritten as an equivalent system of the second kind: Z 2π ∞ X –1 gn + anm gm = fn , n = 0, 1, . . . , fn = hn dn , dn = Jn [kρ(ψ )] dψ , m=0

anm =

bnm d–1 n,

n ≠ m,

ann =

d–1 n

Z

0

2π

(8.93)

Jn [kρ(ψ )] cos(2nψ ) dψ .

0

The properties of the solution to system (8.93) are determined by the asymptotic behavior of anm and fn as n, m → ∞. Using the well-known asymptotics of the Bessel function, 1 z n , n → ∞. (8.94) Jn (z ) ∼ n! 2 © 2005 by CRC Press LLC

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After some algebra, we obtain at n → ∞ Z 2π Z 2π Jn [kρ(ψ )] cos [(n + m)ψ ] dψ Jn [kρ(ψ )] cos [(n – m)ψ ] dψ + bnm = 0 0 Z 2π Z 2π kn kn n –i|n–m|ψ –i(n+m)ψ ρ (ψ )(e +e ) dψ , dn ∼ ρn (ψ ) dψ . ∼ n!2n 0 n!2n 0

(8.95)

The coefficients hn can be represented likewise. The asymptotics of integrals (8.93) and (8.95) can be found by the saddle-point method and depends on the analytic properties of ρ(ψ ), which determines the boundary contour l (see Fedorjuk, 1977). Explicit asymptotic representations of these integrals can be derived when ρ(ψ ) has the general form ρ(ψ ) =

∞ X

an cos(nψ ),

a0 > 0,

an ≥ 0,

n = 1, 2, . . . ,

0 ≤ ψ < 2π .

(8.96)

n=0

In this case, by introducing the new variable z = exp (iψ ), the coefficients b nm in (8.95) can be represented in the form Z ∞ 1 X ik nS1 (z) nS2 (z) { e + e } dz , ρ ( z ) = an z n , an = a–n , bnm ∼ – n!2n |z|=1 2 n=–∞ (8.97) n+m+1 |n – m| + 1 , µ2 = . Sj (z ) = ln ρ(z ) – µj ln z , j = 1, 2, µ1 = n n The expression for dn is analogous, with µj = 0. The general properties of phase functions similar to S 1,2 were examined in detail in Fedorjuk (1977). In particular, it was proved that the saddle point z = x 0 (µj ) is always positive and is determined by the equation zρ0 (z ) = µj ρ(z ).

(8.98)

However, a stronger result can be stated in our problem: x 0 (µj ) > 1. Indeed, it was proved in Fedorjuk (1977), Lemma 5.2, that "∞ #" ∞ #–1 X X zρ0 (z ) n –n n –n = nan (z – z an (z + z ) (8.99) ρ(z ) n=0 n=0 is a monotonically increasing function. It vanishes at z = 1 and tends to plus infinity as z → +∞. Consequently, for any µj > 0, the function in (8.99) equals µj at the single point z = x0 (µj ) > 1. Thus, the integral in (8.97) has the following asymptotic form (where a certain inessential factor is omitted): n Z ρ(x0 ) exp[nSj (x0 )] = n–1/2 , x0 = x0 (µj ) > 1, n → ∞. exp [nSj (z )]dz ∼ µj 1/2 n x |z|=1 0 (8.100) The crucial point in further analysis in the proof of the fact that S j (x0 ) = Sj [x0 (µj )] is a monotonically decreasing function of µj . Indeed, it is easy to see that ∂S dx0 dS = – ln x0 = – ln x0 < 0, dµj ∂x0 dµj

x0 > 1,

(8.101)

which follows from ∂S/∂x0 = 0, since x0 is a saddle point.


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This property implies that the coefficients dn , which correspond to µj = 0 in the integral (8.100) and arise in the denominators of anm and fn in (8.93), are exponentially large as compared to the numerators of the corresponding fractions, provided that µ j in these fractions is not infinitely small. More specifically, for n → ∞, the following holds: n{S[x0 (µ1 )]–S[x0 (0)]} anm = bnm d–1 , n ∼ e

ann ∼ e

n{S[x0 (2)]–S[x0 (0)]}

,

fn =

hn d–1 n

µ1 = |n – m|/n, ∼ e

x0 (0) = 1,

n{S[x0 (1)]–S[x0 (0)]}

n ≠ m,

.

(8.102)

Now, it is clear that monotonic decrease in S as a function of µ implies exponential decrease in the coefficients and absolute terms in (8.93) as n → ∞, provided that µ 1 = O (1) ∼ |n – m| = O (n), n → ∞. Consider the structure of the matrix (8.93) and the structure of its right-hand side for large n in more detail. Obviously, all matrix elements, which are not exponentially small, lie near the principal diagonal. In this sense, the matrix of system (8.93) is similar to a banded matrix. However, its bandwidth increases linearly with the row index n. Infinite algebraic systems with such matrices have not been analyzed previously. In particular, they cannot be referred to any well-known class, such as regular, quasiregular, or Poincar e´ – Koch normal systems (see, for example, Kantorovich and Krilov, 1958). Nevertheless, the truncation method for system (8.93) converges very rapidly. This is explained by the special form of the right-hand side in Eq. (8.89) and entails exponential decrease in the components of fn . Indeed, for the truncated system (8.93), gn = f n –

M X

anm gm ,

(8.103)

m=0

we easily prove by induction that gn = O (exp (–εn)), as n → ∞, for ε > 0. The proof is based on the following fact. When n is large, each summand in (8.103) is exponentially small, because so is anm with a small m (as an element situated far from the diagonal) or gm with a large m. Even though the above properties of the infinite system are proved only for ρ(ψ ) with positive Fourier coefficients (see (8.96)), some test results obtained by using the method in various examples have shown that the exponential decrease in g n with increasing n is a common property for smooth contours l with analytic boundary functions ρ(ψ ). In all test examples, a constant boundary function was used: f (ψ ) ≡ 1. In one example, an ellipse with semi-axes ka = 3 and kb = 1 was considered. In this case, the Fourier expansion (8.96) of kρ(ψ ) = ab[a2 sin2 ψ + b2 cos2 ψ ]–1/2 contains both positive and negative coefficients. Nevertheless, the method converges very rapidly, as illustrated by the curves shown in Fig. 8.6.

g(y) -6 -5 -4 0

90°

180°

270°

y

Figure 8.6. Numerical results for an elliptic domain Figures 8.7 and 8.8 show analogous curves of g (ψ ) obtained for contours l shaped as three-leaf roses with different eccentricities, kρ(ψ ) = 2 + a cos(3ψ ) with a = 21 and a = 1 corresponding to Figs. 8.7 and 8.8, respectively.


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g(y) -5 -4 -3 -2 0

120°

240°

y

Figure 8.7. Numerical results for a domain in the form of a three-leaf rose: a =

1 2

g(y) -4 -3 -2 -1 0

120°

240°

y

Figure 8.8. Numerical results for a domain in the form of a three-leaf rose: a = 1 Solid curves in all figures were calculated by the classical method of boundary integral equations, in which the unknown function g (ψ ) is determined by the following Fredholm integral equation of the first kind whose kernel has a logarithmic singularity at the origin: Z

2π

Z

2π

Y1 (kr ) [(ξ1 – x1 )(ρ0 sin ψ + ρ cos ψ ) r 0 0 (8.104) + (ξ2 – x2 )(ρ sin ψ + ρ0 cos ψ )] dψ , ϕ ∈ [0, 2π ), r = [(ξ1 – x1 )2 + (ξ2 – x2 )2 ]1/2 , x1 = ρ(ϕ) cos ϕ, x2 = ρ(ϕ) sin ϕ, ξ1 = ρ(ψ ) cos ψ , ξ2 = ρ(ψ ) sin ψ , g (ψ )Y0 (kr ) dψ = –2f (ϕ) – k

f (ψ )

where Y1 (x) is a Bessel function of the second kind. Equation (8.104) was solved numerically by the collocation method with the number of nodes N = 180 in ψ ∈ [0, 2π ). Dot-and-dash and dashed curves in Figs. 8.6–8.8 corresponded to M = 10 and M = 25, respectively, where M is the number of equations retained in the truncated system (8.103). The results displayed in Figs. 8.6–8.8 demonstrate that the algorithm proposed here is highly stable and accurate. The presented results are discussed in the author’s paper (Sumbatyan, 2001).

Helpful remarks 1◦ . Note that the run time of the algorithm is comparable to that of the conventional boundary integral equation method in the isotropic case. Indeed, a greater portion of CPU time in the former case is required to evaluate the integrals in the matrix of the infinite algebraic system. However, test results have shown that dimension of the system in that case can be reduced by an order of magnitude, as compared to the classical method of boundary integral equations, in which the elements of the algebraic system are expressed in explicit form.


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2◦ . The method discussed here seems to be a little alien to the considered direct boundary value problem, when compared for example with the BIE method discussed in Chapter 2. However, the proposed approach looks more efficient in the cases when the Green’s function cannot be constructed in a finite form. Indeed, efficiency of the classical BIE approach, based on the Kirchhoff–Helmholtz integral formula, is closely connected with the fact that the Green’s function is expressed explicitly in terms of the exponential (3D case) or the Helmholtz (2D case) function. If we consider, for example, the linear acoustic problem in a medium possessing some kind of anisotropy, then the Green’s function (i.e., the solution to the basic equation with the delta function δ (x – x 0 ) on the right-hand side) is usually expressed by quadrature. Under such conditions the classical BIE method becomes less efficient, but not our approach, since the latter does not require operation with the Green’s function.

8.7. Ill-Posed Method of Auxiliary Sources in Diffraction Theory One of the well-known methods used for solving diffraction problems that has become widely used in recent years is the auxiliary sources (AS) method, in which the unknown current density is disturbed on a certain auxiliary contour (in the two-dimensional case) lying inside the body at which diffraction occurs. This approach is considered to be correct if the auxiliary contour includes all the singularities of the diffracted field, in which case the basic integral equation of the method is uniquely solvable. However, unlike the basic equation of the classical BIE method (Section 2.2), the equation of the AS method is illposed in Tikhonov’s sense. Thus, even though there is a unique solution, it remains an open question how to select an effective numerical method of solution. Here we investigate the equation of both the direct and the inverse problem of diffraction. If the solution of the diffraction problem is sought in the form of the single-layer potential with an unknown density µ(θ ) distributed over an auxiliary contour, then to find the function µ we need to solve the well-known integral equation (see Chapter 2) Z 2π µ(θ )H0(1) [kR(ϕ, θ )] dθ = f (ϕ), ϕ ∈ [0, 2π ], (8.105) 0 2 2 1/2 R(ϕ, θ ) = [ρ (ϕ) + r (θ ) – 2ρ(ϕ)r (θ ) cos(ϕ – θ )] , where ρ(ϕ) describes the true contour, r (θ ) describes the auxiliary contour, and f (ϕ) is expressed in terms of the incident wave field. For simplicity, both contours are assumed to be star-like. The diagram of the scattered wave field in the far zone is expressed in the form Z 2π F (ϕ) = µ(θ )eikr(θ) cos (ϕ–θ) dθ . (8.106) 0

The direct diffraction problem requires to solve Eq. (8.105), then the calculation of the scattered diagram (8.106) reduces to a simple integration. It is proved in Apeltsin and Kyurkchan (1990) that (8.105) has a unique solution if and only if the contour r (θ ) encloses all points at which the singularities of the wave field are concentrated, and the numerical solution can be constructed by any classical method. For instance, in the collocation method the integral in (8.105) is replaced by a finite sum; the equation is satisfied at a finite number of nodes ϕn , n = 1, 2, . . . , N , after which the problem reduces to a linear algebraic system of equations, which can be solved by the Gauss method. Strictly speaking, such an approach is not correct, because (8.105) is a first-order equation with a compact operator (in any natural functional space) and, therefore, the problem of its solving is ill-posed. Only if


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r (ϕ) ≡ ρ(ϕ), a case corresponding to the BIE method, the kernel of Eq. (8.105) has a weak (logarithmic) singularity. We know (see Voronin and Tsetsokho, 1981) that the collocation method is stable for kernels of this kind. We will therefore call this a well-posed case. Naturally, for general contour, stability of the algorithm in this method will be provided only if the auxiliary contour is very close to the true contour. The solution of the inverse problem can be found as follows. First, from the given function F (ϕ), we find the potential µ(θ ) from (8.106). Then, substituting the result for µ(θ ) into (8.105), we reduce the problem to solving the nonlinear equation (8.105) for the function ρ(ϕ). Unfortunately, there are no results on the solvability of (8.106) in the general case. After the ill-posed equations (8.105) and (8.106) have been reduced to linear algebraic systems, the properties of stability of the latter depend on the question how rapidly the eigenvalues λn of the integral operator (8.105) and (8.106) tend to zero as n → ∞ (cf. Hille and Tamarkin, 1931). The faster they decrease the smaller the modulus of the first eigenvalues of the matrices obtained by the discretization. A study of this spectral problem has been carried out by Hille and Tamarkin (1931), which shows that Eqs. (8.105) and (8.106) belong to different classes with respect to the function µ(θ ). The kernel in (8.105) is only analytic in a finite domain of the complex variables ϕ, θ and, of course, the functions ρ(ϕ) and r (θ ) themselves are regarded as analytic in this case. It follows from Theorem 10.1 of the cited work that λ n < d–n/4 , n → ∞, d = (1 + b2 )1/2 + b, where b is the dimension of the imaginary semi-axis of the ellipse inside which the kernel of the equation is analytic with respect to the variable ϕ. In the case under consideration here, analyticity of the kernel of (8.105) is broken at the points where ρ2 (ϕ) + r 2 (θ ) – 2r (θ )ρ(ϕ) cos(ϕ – θ ) = 0. Since | cos(ϕ – θ )| ≤ cosh y , ϕ = x + iy , it is easy to show that the parameter b is determined by the quantity min ln[ρ(ϕ)/r (θ )]. Thus λn → 0 like the terms of a geometric progression, and removing the true contour far from the auxiliary contour makes the quantity d–1/4 smaller. The kernel in (8.106) is an entire function with respect to both its variables ϕ, θ . It follows from Theorem 11.1 of Hille and Tamarkin (1931) that λ n → 0 here are not slower than n–1/4 . Hence Eq. (8.106), considered with respect to the function µ(θ ), is much more ill-conditioned than (8.105). The results we have obtained here for some simple cases can be made more precise. For example, if r (θ ) ≡ r , ρ(ϕ) ≡ ρ, r ≤ ρ, then the eigenfunctions for (8.105) and (8.106) are of the form cos(nθ ) and sin(nθ ). Here, respectively, n n kr 1 2i r , λn ∼ . (8.107) λn ∼ – n ρ 2 n!

Formula (8.107) confirms the rule obtained for the general case. As one might expect, the closer r to ρ, the better the conditionality of (8.105). In the limiting case corresponding to the BIE method, λn = O (1/n), leading to a stability of the calculation during the discretization (see also below Section 9.2). The conditionality of (8.106) is independent of the true contour. From these results, it follows that the basic question arising in the numerical implementation of the AS method concerns the conditionality of the corresponding linear algebraic systems. This is a more relevant factor than the solvability of (8.105) and (8.106). One of criteria for choosing the auxiliary contour proposed in Apeltsin and Kyurkchan (1990) was the absence of strong oscillations of the density µ(θ ) distributed on that contour. Yet, as specific calculations show, the auxiliary sources for the simplest test example ρ(ϕ) ≡ 3, r (θ ) ≡ 0.5, N = 18 can be uniformly concentrated around a circle r = 0.5, with densities of the order of 104 . The solution by the Gauss method of the (18 × 18) systems corresponding


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to the direct and inverse problems then gives results accurate to within 3–4 significant digits, which justifies the use of this approach. Another criterion of correctness of the auxiliary contour choice in Apeltsin and Kyurkchan (1990) was that the discrepancy between the leftand right-hand sides at points situated half-way between selected nodes should be small. It was found that this could only be the case for a contour which encloses all the singularities of the wave field. The specific calculations of the diffraction problem considered in the cited work on an ellipse with semi-axes 3 and 1.2 for small values of N (N = 18, 36, etc.) show that these discrepancies are of the same order in the following cases: (a) r (θ ) corresponds to an ellipse with semi-axis 2.8 and 1 which includes the singularities; (b) the contour r (θ ) ≡ 1 does not include the singularities; (c) the contour r (θ ) ≡ ρ(θ ) corresponds to the BIE method and is a priori admissible. In all three cases the solutions of the direct and inverse problems are with 3 or 4 significant digits. The scattered field diagram is obtained with the same accuracy. It follows from the foregoing consideration if the AS method in its discrete modification is not associated with an integral equation (which is what happens when only a small number of nodes is chosen), any interior auxiliary contour can be used, and the conditionality of the resulting system becomes a main parameter which depends on the choice of the contour. It is always true that the closer the auxiliary contour to the true contour, the better the conditionality of the corresponding algebraic system. Since the resulting systems are always ill-conditioned, some regularization method must be used to solve them. The following approach is used here: (a) the method based on orthogonal Hausdorff transformations; (b) the QR-algorithm; (c) the method of steepest descent. To verify the dependence of the conditionality number on the position of the auxiliary contour, we will find its values in a few cases. Suppose that the true contour is a circle with kρ(ϕ) ≡ 3. Let the interior contour also be a circle, with various possible values of the parameter kr . The performed calculations show for the conditionality number β = M/m of the corresponding systems the following results. When kr varies from 1 to 3, β varies from 104 to 101 in the direct problem and from 106 to 103 in the inverse problem. Suppose now that the true contour is an ellipse with semi-axes ka = 3 and kb = 1.2. The singular points of the wave field are then situated at the foci kc = ±2.75 (cf. Apeltsin and Kyurkchan, 1990). The first of the three auxiliary curves encloses the singularities, but the second and third do not. The corresponding conditionality number β is such that for the direct problem it is always of the order 102 , and for the inverse problem varies from 107 to 1010 . The following conclusions can be drawn form these results. 1. The conditionality of the inverse problem is always several orders of magnitude worse than for the direct problem. 2. The conditionality of the direct problem depends on the position of the auxiliary contour, improving as it approaches the true contour. We have found that the conditionality depends much more on the position of the auxiliary contour than on whether it encloses the singularities or not. 3. The conditionality of the inverse problem depends much less on the position of the auxiliary contour, always being poor. It should be noted that the conclusions given here might not be directly applicable to the case where there is a large number of auxiliary sources. The behavior of the computational singularities of the AS method as N increases is basically an open question and requires more thorough study. As noted above, the solution of the inverse problem can be obtained by the successive solution of (8.106) and (8.105). After substituting the function µ(θ ) found from the linear Eq. (8.106) into (8.105), this equation becomes nonlinear with respect to the function ρ(ϕ). This approach is genetically close to the method first proposed in Kirsch et al. (1988),


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in which the resulting nonlinear ill-posed problem is solved by the Tikhonov smoothing functional method. The approach we are using is based on the method of gradient descent for minimizing a functional, which is the square of the norm of the discrepancy. The gradient of this functional in this problem can be easily written out explicitly. As one might expect, the accuracy with which the contour ρ(ϕ) is reconstructed depends closely on the successful choice of the auxiliary contour r (θ ). Two examples on the applications of these results can be found in the author’s work (Bratsun and Sumbatyan, 1993). The first example is related to an elliptic domain with semi-axes 2.2 : 1.2, which was restored with auxiliary contour kr (θ ) ≡ 1.2. The second cat-face contour is described by the parametric equations kx(t) = cos t + 0.65 [cos(2t) – 1],

ky (t) = 1.5 sin t,

0 ≤ t < 2π ,

(8.108)

with kr (θ ) ≡ 0.8. Note that the auxiliary contour in the first example does not include the singularities, which are at points ±1.84. A similar situation also appears to apply in the second case. Unfortunately, more complicated boundaries ρ(ϕ) cannot be reconstructed by this method. Another drawback is that the method cannot be applied if we study the case where we know only the modulus |F (ϕ)| rather than F (ϕ) itself, then Eq. (8.106) becomes nonlinear with respect to the function µ(θ ).

Helpful remarks The method described at the present section can be used as a basis to generically study illposed integral equations of the first kind Au = f . If, for simplicity, the integral operator A is symmetric and positive, then all its eigenvalues are positive: λ n > 0, ∀n, and the respective eigenfunctions {ϕ} form anPortho-normal system. Then if we seek the solution as a series in the eigenfunctions, u = un ϕn , then un = fn /λn , where fn = (f , ϕn ), so the issue of convergence of the series depends on the issue of the asymptotics of f n and λn as n → ∞. In this sense, solvability of the equation is linked with the fact in which degree the kernel K (x, ξ ) and the right-hand side f (x) are smooth. This idea was implicitly used in our study in Section 8.6, and now in the present section. We can conclude from the consideration carried out in the last two sections that solvability takes place when the right-hand side is smoother in a sense than the kernel. This idea will also be used in Section 9.2. Some thorough results on asymptotic distribution of eigenvalues λ n , as noted above, can be found in Hille and Tamarkin (1931).

8.8. A Method of Global Random Search in Inverse Problems Here we propose an alternative approach to solve ill-posed problems which is related to the operator equation of the first kind Au = f , (8.109) where A is a compact operator acting in some Hilbert space H . The problem of finding u, when the right-hand side f is known, is therefore an inverse problem. It is known (Tikhonov and Arsenin, 1977) that Eq. (8.109) is considered as Tikhonov’s ill-posed problem in the sense that a large change in the solution u may correspond to a small change in the righthand side. This leads to instability of the basic numerical methods when they are applied to Eq. (8.109) (see the previous sections). To overcome these difficulties, special numerical methods have been developed. One of possible approaches is based on the ideas of regularizing Eq. (8.109) by means of an appropriate small perturbation of the operator A that was discussed in Section 8.2.


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In this section a method which is natural from the physical point of view is proposed. If approximations un are selected randomly, it is possible to evaluate their closeness to the exact solution u∗ , from the closeness of the left- and right-hand sides, that is, from the smallness of the discrepancy functional, M (un ) = ||Axn – f ||2 .

(8.110)

But it is known (see Sections 8.1 and 8.2) that the ill-posed equation (8.109) cannot be reduced to minimization of the functional M (u), (8.110). Instead, it is necessary to consider the smoothing functional (see Section 8.2) Mα α(u) = ||Au – f ||2 + αΩ(u),

0 < α 1.

(8.111)

To minimize functional (8.111) we will use the method of global random search (see Zhiglyavskii, 1985). Unlike direct random search, it has important properties which enable the process of finding a good approximation un to be accelerated. This is valid because 1) random sampling of values un in the neighborhood of the points X , for which the values of Mα (X ) are smaller, happens more frequently than that in the neighborhood of the points Y , where the values of Mα (Y ) are larger, and 2) domains, in which random values un are chosen, are gradually contracted to the small neighborhoods of the points with small values of M α (u). This algorithm has been tested by examples on minimization of finite-dimensional functions of a small number of variables. Here we use it to solve functional equations in the Hilbert space. We will first consider the integral equation of the first kind Z 2π Au = cos[ρ cos(ϕ – θ )] u(θ ) dθ = f (ϕ), 0 ≤ ϕ < 2π , ρ = const > 0, (8.112) 0

as a model example. Its kernel is infinitely differentiable, which implies that this equation is extremely ill-posed. We will find its solution in the form of a Fourier series with a finite number of terms u(θ ) =

M X

am cos mθ ,

0 ≤ θ < 2π .

(8.113)

m=0

For simplicity, we consider here the case when the right-hand side f (ϕ) (as well as the kernel) is an even function of ϕ. In this approach the minimization of the functional (8.111), based on the method of global random search described above, implies random sampling of sets of real numbers (a0 , a1 , . . . , am )n . It is proposed to choose a regularizing functional in the form (see above Section 8.2) Ω(u) = ||u||W22 (0,2π)

(8.114)

where W22 is Sobolev’s space, and the basic space is H = L2 (0, 2π ). The results of the application of the proposed algorithm for two right-hand sides, which are f1 (ϕ) = 2πJ0 (ρ) and f2 (ϕ) = –2πJ2 cos 2ϕ, can be found in the author’s paper (Sumbatyan, 1992), where we took the values ρ = 1, M = 4, α = 10 –1 to 10–4 . The number of random samples of the solution defining the number of calculations of the direct operator A was equal to N = 300. The method given in this section is especially effective for problems characterized by a large number of local minima of functional (8.111), as well as by large values of the


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gradients of the functional. In such problems it is practically impossible to find the global minimum by regular methods. As a detailed analysis indicates, the problems of recognizing an object from the wave field scattered by it are distinguished by these properties. When investigating the problem on reconstruction of the boundary contour, we will restrict ourselves, for simplicity, to the two-dimensional case and to the simplest model of an acoustic medium described by a single Helmholtz equation. Then it is possible to reduce the problem to a system of two nonlinear integral equations (see the previous sections) Z ik(q⋅x) e g (x) dsx = f (ϕ), 0 ≤ ϕ < 2π , (8.115) l

Z

H0(1) (k |x – y |)g (x) dsx = eik(q⋅y) ,

l

y ∈ l,

q = –{cos ϕ, sin ϕ}.

(8.116)

Here k is the wave number and H0(1) is the Hankel function. In Eqs. (8.115) and (8.116) the unknowns are the function g (x), connected with the normal derivative of the velocity, and the function x ∈ l, which defines the position of the boundary contour l (the latter may be specified, for example, in some parametric form). The results on reconstruction of two objects, an ellipse with a ratio of its semi-axes equal to 3 : 1, and a semi-circle with a diameter of 10, from the amplitude of circular back scattering, can be found in Sumbatyan (1992). As in the model example, the parametric representation of the contour l was specified in the form of a finite truncation of a Fourier series in the polar coordinate system with M = 4.

Helpful remarks 1◦ . System (8.115), (8.116) holds for the case when the amplitude of the back scattering from the object is known for the whole range of variation of the scanning angle ϕ ∈ (0, 2π ). This method of scanning, when the directions of propagation of the incident wave and the reflected wave coincide, corresponds to the echo method, widely used in ultrasonic testing, in radio-location, in underwater acoustics, and many other applications. In this method the same ultrasonic sensor device serves both as the emitter and the receiver. The amplitude of the reflected signal f (ϕ) is therefore known, but the phase, as a rule, is unknown. It is assumed that similar measurements may be carried out in principle for any angle of incidence ϕ ∈ (0, 2π ).

2◦ . The direct problem of diffraction consists in the calculation of the function f (ϕ) along the known contour l. To do this it is necessary, first, to solve the boundary integral equation (8.116) for the function g (x), x ∈ l, and then to calculate the quadrature (8.115). It is obvious that the corresponding operator Au in (8.109) is nonlinear. Calculation of the direct action of a nonlinear operator is therefore reduced to solving one linear equation and one quadrature. This situation is typical for nonlinear inverse problems. Unlike the above case the inverse operator A–1 (if it exists) is extremely nonlinear. When using regular methods to invert Eq. (8.109) it would be necessary to calculate the Frechet derivative for the operator A, which is quite difficult to do. The method described here requires only the calculation of the direct operator, which, as noted above, involves linear operations only.

8.9. Ill-Posed Problem on Reconstruction of Convex Hull of the Obstacle in Acoustic Medium Let us come back to the problem of practical realization of algorithm on reconstruction of convex parts of a smooth obstacle in acoustic medium from the far-field back-scattered


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diagram, when the known information is contained in the time-of-flight data. In this case the problem of reconstruction of the convex hull S0 can be reduced to calculations by explicit formulas (7.81), on the basis of the support function usually known from experimental measurements. Note that these formulas require operation with numerical differentiation, a process that is well known to be an ill-posed problem (see Tikhonov and Arsenin, 1977). In practical scanning there is typically a discrete array of directions where they carry out measurements with some error. We thus can conclude that direct usage of formulas is impossible in practice for two reasons. First of all, there is an error in definition of time of flight t(q). Secondly, a finite step of discrete measurements is usually not sufficiently small to provide precise and correct computations of the derivatives in those formulas. Even for convex parts of the boundary surface the obtained function p(n) is only piecewise-smooth, and so its first derivatives p0α , p0β used in reconstruction of the convex hull S0 may differ from actual values not only by their values but even by their sign. Taking into account the above discussed specifics of the studied problem, let us construct an algorithm, which permits approximation of the function p(α, β ) by a certain smooth function, for which its values and values of its derivatives are close to really obtained from measurements. In the considered situation when in practical scanning we obtain, instead of a smooth function, its approximate values in some nodes, the bases of our algorithm is an approximation of the sought function by cubic splines (see, for instance, De Boor, 1978). Let us restrict the study by the 2D case, for simplicity, and assume that we know approximate values pi , i = 1, 2, . . . , N of a certain smooth function (in our case, the function p(n)) at the points n1 , . . . , nN , and the estimate ∆pi of the mean-square deviation of pi from the real values p(ni ). In practice, ∆pi does not exceed 10%. Let us construct the function f = fδ that for a given parameter δ ∈ [0, 1] determined by the mean-square deviation minimizes the functional (see De Boor, 1978) 2 Z nN N X 00 2 pi – f i + (1 – δ ) f (θ ) dθ , (8.117) π (f ) = δ ∆p i n 1 i=1

over all functions f possessing second-order derivatives. Minimization of the functional π (f ) is a compromise between two requirements: to provide a good approximation to the given values p i , and to provide a sufficiently smooth approximating function. It is known (see De Boor, 1978) that fδ is a second-order spline with simple points of interpolation at the points n2 , . . . , nN –1 and satisfying the condition fδ00 (n1 ) = fδ00 (nN ) = 0. Let us assume, for simplicity, the step ∆n between any pair of neighbor normals to be constant, and let us introduce the notations: ai = fδ , ci = fδ00 , i = 1, 2, . . . , N . In practice, for the 2D reconstruction problems discussed below, the dependence of the function p(n) upon the direction of outward normal n is equivalent to its dependence on the angle of incidence (the same as the angle of observation). Therefore, in practice, we can put ∆n = ∆θ , i.e., the angular step when scanning around the obstacle. Then the condition of continuity of the first-order derivatives f δ0 for a closed contour takes the following form: (ci–1 + 4ci + ci+1 )(∆n)2 = 6(ai–1 – 2ai + ai+1 ),

i = 2, 3, . . . , N – 1,

2

(cN –1 + 4cN + c1 )(∆n) = 6(aN –1 – 2aN + a1 ),

(8.118)

2

(cN + 4c1 + c2 )(∆n) = 6(aN – 2a1 + a2 ). For all that, the functional (8.117) takes the following form: 2 N N X X 1–δ pi – a i + ∆n c2i + ci ci+1 + c2i+1 . π (f ) = δ ∆pi 3 i=1 i=1

(8.119)


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Further, by applying in relations (8.118) approximation of the second-order derivatives ci = fδ00 by central finite differences, with the help of values of the unknown function at three nodes, we can reduce the condition of minimum of the functional (8.117) to the following form written as a vector equality: 1–δ 2δ (p – a) + Xc = 0, – (∆p)2 12∆n (8.120) p = {p1 , . . . , pN }, a = {a1 , . . . , aN }, c = {c1 , . . . , cN }, where X is a certain sparse matrix. Nontrivial elements of this matrix form a diagonal band of five lines, and additionally two triples in the upper right and lower left corners of the matrix. Recall that ∆p is a maximum error of measurements, which is in our case around 0.1 of the representative size of the obstacle. In order to obtain a self-consistent algebraic system, we should add to system (8.120) also relations (8.118) written in the vector form (∆n)2 R c = 6 Q a, (8.121) where R and Q are some sparse matrices. By their structure, these are three-diagonal banded matrices, with two additional elements, one in the upper right and another in the lower left corners. Let us substitute (8.121) into (8.120), then we arrive at a system of linear algebraic equations, to define the vector of the second-order derivatives c: 1–δ 12δ 2δ (∆n)2 R+ QX c = Q p. (8.122) 2 (∆p) 2∆n (∆p)2 In investigation of this system it is convenient to introduce the parameter M 2 N N X (1 – δ )2 (∆p)2 X pi – f i = (X c)2i . (8.123) M= 2 ∆p (24 δ∆n ) i i=1 i=1

On the basis of the introduced parameter M we can arrange the estimate of the minimum value of the parameter δ , 0 ≤ δ ≤ 1 so as to provide the value of M less than a certain a priori given value L. As soon as the vector c is defined from system (8.122), we can determine vector a of values of the unknown function fδ at the chosen nodes, from the following relation: (1 – δ )(∆p)2 X c. (8.124) a=p– 24 δ∆n This method allows us also to determine approximate values of the first three derivatives: ci f 000 (ni ) ai+1 – ai – – δ , fδ (ni ) = ai , fδ0 (ni ) = ∆n 2∆ni 6(∆n)2 (8.125) ci+1 – ci . fδ00 (ni ) = ci , fδ000 (ni ) = ∆n √ The quantity L is proposed in De Boor (1978) to be taken somewhere between 2N and N . Such a choice permits arrangement of the function f δ to provide a sufficiently accurate approximation p(n) together with its first and second derivatives at the points n i , and to be simultaneously smooth enough. The function p(n) obtained by the thus developed method finally gives, after the substitution to formulas (7.81), the sought surface of the convex hull for the considered obstacle. Below we demonstrate the proposed reconstruction algorithm on some real cylindrical flaws made in aluminum plates. The representative size of all defects was around 10– 15 mm. The measured back-scattered diagrams were obtained with the angle of incidence changing in 5◦ increments. Scattering patterns with the respective results of reconstruction are reflected for some flaws in Figs. 8.9–8.12.


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x2 x2

x1

x1

Figure 8.9. Support function for the first obstacle and the corresponding reconstructed convex hull

x2 x2

x1

x1

Figure 8.10. Support function for the second obstacle and the corresponding reconstructed convex hull

x2

x2

x1

x1

Figure 8.11. Support function for the third obstacle and the corresponding reconstructed convex hull


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x2

x2

x1

x1

Figure 8.12. Support function for the fourth obstacle and the corresponding reconstructed convex hull

Helpful remarks It should be noted that in the considered 2D case Cartesian coordinates of the convex hull S 0 are expressed by the following formulas: x1 = p(θ ) cos θ – p0θ (θ ) sin θ , x2 = p(θ ) sin θ – p0θ (θ ) cos θ

(8.126)

(compare with Eq. (7.81) valid for 3D case), where θ is the scanning angle, equal at the same time to the angle between the normal to the boundary of the obstacle at the reflection point (x1 , x2 ) and positive direction of the axis x1 . In our calculations we used the following values: N = 360◦/5◦ = 72 and 15 ≤ L ≤ 25. For the results of reconstructions shown in Figs. 8.9–8.12, solid lines designate real boundary contour, and dashed lines the reconstructed ones. Note that we operate both with convex and nonconvex obstacles, and the shown results confirm that the proposed method to construct the convex hull of a smooth obstacle can be used for obstacles of both types.


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Chapter 9

Numerical Methods for Irregular Operator Equations 9.1. Steepest Descent Method: Stability and Improvement of the Convergence As shown in Chapter 2, a diffraction problem in acoustic media may be reduced to integral equations obtained from the Kirchhoff–Helmholtz integral formula. The Dirichlet problem is described by an equation of the first kind and the Neumann problem by an equation of the second kind, if we speak about the direct BIE method: Z ψ (y ) G(y0 , y ) dSy = F1 (y0 ), y0 ∈ S , (9.1) S

p(y0 ) – 2

Z

S

p(y )

∂G(y0 , y ) dSy = F2 (y0 ), ∂ny

y0 ∈ S .

(9.2)

Here ψ (y ) = ∂p/∂ny , p is the diffracted pressure on the boundary surface S ; F 1 and F2 are functions expressed in terms of boundary functions. To be more specific, we assume that the surface is closed and sufficiently smooth; the normal n y is directed outwards the surface S . Equations (9.1) and (9.2) describe the diffraction problem both in the 2D and 3D cases. In such cases we have, respectively G(y0 , y ) =

i (1) H (kr ), 4 0

G(y0 , y ) =

1 eikr , 4π r

r = |y – y0 |.

(9.3)

Note that equations (9.1), (9.2) may be written out for full wave field too. This changes only the form of the right-hand sides F1 and F2 . As we could see from the previous study, equations (9.1), (9.2) are investigated in detail, and a number of analytical and numerical methods have been developed to solve these equations. The exact analytical solutions are known only for canonical regions (circle, sphere, ellipse, segment, etc.). In the high-frequency range (ka 1, a is a characteristic size of the S -surface) the shortwave asymptotic methods are efficient for regions of complex shape. As regards direct numerical methods, we note once again that these methods become inefficient for extremely high frequencies. The main disadvantage of numerical methods is their isolation from the physical essence of phenomena. Besides, it is known that numerical methods lose their stability for large values of ka (see also the Preface). Here we propose a method that uses a combination of analytical and numerical techniques. The analytical solution obtained from the Kirchhoff theory is used as an initial-step approximation for a certain iterative method. It is effective for ka 1 only, and for lower values of ka this is improved by the iterative method, and we show that each step of the proposed technique improves the chosen analytical solution.


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Let us write the considered equations (9.1), (9.2) in the operator form Bx = y ,

(9.4)

and consider the linear operator B acting in some complex Hilbert space H , which may be chosen as L2 (S ). We assume the operator B to be self-adjoint and positive. We always can achieve such properties of B representing it in the form A ∗ A, where A is the operator of the initial integral equation. In this case y = A∗ f , where f is a right-hand side of the initial integral equation. Let the spectrum of the operator B lie on the interval [m, M ], 0 ≤ m < M . For the first-kind equation, m = 0 always, which follows from compactness of the integral operator. We proved in Section 8.3 that if equation (9.4) has a solution, x ∗ , then the iterative steepest descent method (SDM) xn+1 = xn –

||Lxn ||2 L n x, (Lxn , BLxn )

Lxn = Bxn – y ,

(9.5)

converges monotonically to this solution in the norm of the chosen space H . This means that ||xn+1 – x∗ || ≤ ||xn – x∗ ||. For all that, if m > 0, then n ||Lx0 || M – m ∗ ; (9.6) ||xn – x || ≤ m M +m if m = 0, then F (xn ) = B (xn – x ), xn – x ∗

∗

1 . =O n

(9.7)

It is easy to demonstrate that this result is correct if we use some allied methods, instead of SDM, which can be written as (see Section 8.3) xn+1 = xn –

||Lxn ||2 A∗ Lxn , ||A∗ Lxn ||2

or xn+1 = xn –

(A∗ A = B ),

(BLxn , Lxn ) Lxn . ||BLxn ||2

(9.8)

(9.9)

Recall that for convergence of SDM the existence of the solution x ∗ is only sufficient. Below we will use this fact in the case when the solution is not unique. The numerical realization of the considered iterative techniques implies that the integral operator is changed by a finite-dimensional one. In particular, when solving equations (9.1), (9.2) by a collocation method, we arrive at a linear algebraic system N × N . The convergence in (9.5) takes place; however in the case of an equation of the first kind we have m > 0 always, except the special cases mentioned below, hence the estimate (9.6) for the rate of convergence is always valid. At the same time, in the case when the integral equation (9.1) is approximated by a finite-dimensional operator equation with the large number of nodes N , the finite dimensional operator is a good approximation to the initial compact one, therefore m/M 1. Hence, the matrix of the system is ill-conditioned. It is easily seen that the rate of the convergence here is very slow. In the case of equation (9.2) concrete calculations at ak 1 show that the matrix of the corresponding algebraic system is poorly posed too. From this point of view, equations of the second kind differ not very much by their qualitative properties from equations of the first kind, in the case of the high-frequency process.


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In connection with this, the problem of the SDM acceleration seems to be very important. Different approaches to acceleration of iterative methods are known in the literature. Here we use an acceleration method, which is based on some asymptotic properties of SDM for large n. It is shown in Samarsky and Nikolaev (1978) that when n → ∞ the iterative process (9.5) has the following properties: ||Axn – f ||2 F (xn ) = = c2 , F (xn–1 ) ||Axn–1 – f ||2 M –m ∗ 2 ∗ 2 . xn – x = c (xn–2 – x ) c = M +m

(9.10) (9.11)

It follows from (9.11) that, when SDM approaches its asymptotics, the next iteration may be chosen in the form xn – c2 xn–2 , (9.12) x˜ = 1 – c2 where c2 may be calculated from (9.10). The usage of such an approach in concrete numerical examples demonstrates fast acceleration of the iterative process for any matrix equations, including ill-conditioned ones (m/M 1, c 2 ∼ 1). In some more details, we can observe the following properties of the iterative process. For sufficiently large n the process (9.5) approaches the worst of its estimate in inequality (9.6), and the asymptotic relation (9.10) becomes approximate as an equality. Under such conditions the difference F (x n ) at each step decreases c2 times; consequently, for ill-conditioned problem with c 2 ∼ 1 it decreases very slowly. For all that, implementation of one step (9.12) eliminates this cycling and further the process (9.5) converges more rapidly. This property is well observed from the table related to the first of the examples considered below. Recall that the first-kind equation (9.1) is ill-posed in the sense of Tikhonov, hence the iterative process (9.5) defines some family of regularizing operators R n for equation (9.1). Indeed, the iterative process at every step defines some nonlinear operator x n = Rn y and in order to calculate Rn it requires application of a finite number of continuous operators for any finite value of n. This property of Rn and convergence of the SDM process (9.5) proves the fact that Rn is a regularization operator. It is known (see Chapter 2) that for the closed surface S there exist the values of the parameter k = kj (j = 1, 2, . . . ) at which the operators corresponding to equations (9.1), (9.2) are not invertible. Such a situation takes place at frequencies that correspond to resonances of the corresponding interior problem. Some authors used various methods to overcome this difficulty. Here we show that our approach automatically avoids this obstacle. It is proved in Colton and Kress (1983) that for equation (9.1) at the critical frequencies k j the right-hand side is always orthogonal to solutions of the homogeneous equation. It is also proved that under such conditions equation (9.1) is solvable (not uniquely) for any boundary function in the Dirichlet problem, and the wave field computed on the basis of this solution is defined uniquely. Since a solution to equation (9.1) exists, the described method guarantees the iterative process to convergence to a certain solution (in case the solution is not unique). This means that the proposed method provides a stable solution to the Dirichlet problem at critical frequencies too. It is also well known that the indirect BIE method permits reducing of the Dirichlet problem to an integral equation of the second kind (see Chapter 2). But in this section we do not use such an opportunity because at the critical frequencies in this case the right-hand side is not orthogonal to solutions of the adjoint homogeneous equation. It should also be noted that if the size of the complex-valued matrix in discretization of Eqs. (9.1), (9.2) is equal to N , then solution of the matrix equation by any direct method


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requires around N 3 arithmetic operations. At the same time one step of the process (9.5) requires near N 2 operations. With large ka it is necessary to choose the number N of the nodes rather large for adequate description of the field on the boundary surface. In such situation the SDM is more stable from the computational point of view. Thus, in our approach we essentially use the following two new points: 1) acceleration of SDM; this allows us to achieve fast practical convergence; 2) the fact that SDM overcomes the classical difficulty connected with presence of critical frequencies in the Dirichlet problem. The following examples were considered to test efficiency of the proposed method. First of all, we tested the proposed approach by a certain 2D problem about diffraction of the plane wave by an acoustically hard round disk. We took ka = 5, the number of points N = 72. For such moderate wavelengths the solution on the Kirchhoff theory as the first step of iterations is ineffective. The proposed method allows us to improve it. The SDM rates of convergence without acceleration and with acceleration by Eq. (9.12) after each 10 iterations are compared in the table below. n

F (xn )

c2

F ∗ (xn )

n

F (xn )

c2

F ∗ (xn )

1 2 3 4 5 6 7 8 9 10 11 12 13 14 15

2.61 × 100 7.67 × 10–1 2.54 × 10–1 9.63 × 10–2 4.14 × 10–2 2.17 × 10–2 1.41 × 10–2 1.09 × 10–2 9.36 × 10–3 8.44 × 10–3 2.48 × 10–3 7.15 × 10–3 6.62 × 10–3 6.14 × 10–3 5.69 × 10–3

0.293 0.332 0.377 0.430 0.523 0.649 0.776 0.856 0.901 0.917 0.924 0.925 0.926 0.926 0.927

2.61 × 100 7.67 × 10–1 2.54 × 10–1 9.63 × 10–2 4.14 × 10–2 2.17 × 10–2 1.41 × 10–2 1.09 × 10–2 9.36 × 10–3 8.44 × 10–3 1.93 × 10–2 6.10 × 10–3 2.03 × 10–3 7.33 × 10–4 3.31 × 10–4

16 17 18 19 20 21 22 23 24 25 26 27 28 29 30

5.27 × 10–3 4.89 × 10–3 4.53 × 10–3 4.20 × 10–3 3.89 × 10–3 3.61 × 10–3 3.35 × 10–3 3.11 × 10–3 2.88 × 10–3 2.67 × 10–3 2.48 × 10–3 2.29 × 10–3 2.13 × 10–3 1.97 × 10–3 1.83 × 10–3

0.9271 0.9271 0.9271 0.9272 0.9272 0.9272 0.9273 0.9273 0.9273 0.9273 0.9274 0.9274 0.9274 0.9274 0.9274

1.92 × 10–4 1.42 × 10–4 1.19 × 10–4 1.07 × 10–4 9.76 × 10–5 8.06 × 10–5 4.57 × 10–6 3.35 × 10–7 5.72 × 10–8 3.30 × 10–8 2.63 × 10–8 2.31 × 10–8 2.09 × 10–8 1.91 × 10–8 1.72 × 10–8

Here c2 is calculated as in (9.10), n is the number of the iteration, F (x n ) is the value of the discrepancy functional without acceleration, F (x n ) are analogous values of the functional, by applying the proposed acceleration method. We can observe stabilization of the parameter c2 in the table, which is in agreement with the described asymptotic properties of SDM. We can see that the value of c2 is around 1 and this is connected with the poor definiteness of the matrix, which results in slow convergence of the process (without acceleration). At the same time, the acceleration makes the process more rapidly convergent: the value of the discrepancy decreases by 8 orders when compared with the first step of the iteration. Further, we continued to test efficiency of the proposed approach, by consideration of diffraction of the 2D plane acoustic wave by a hard ellipse. Here ka = 5, kb = 0.5 (a/b = 10), N = 180. It is known that efficiency of the BIE method decreases considerably for elongated bodies. It is caused by the fact that such bodies require very large number of nodes N . Besides, the numerical investigation shows that the matrix of the corresponding algebraic system becomes extremely ill-conditioned for elongated bodies. The performed computations prove that the proposed method is efficient in this case too. Then we considered the problem about diffraction of the plane wave by a soft cylinder with ka = 3.8317; this corresponds to interior resonance. We took here N = 72. The


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implementation of the SDM with the above described acceleration allows us to obtain a solution efficiently in this case also. The results are presented in more detail, with some diagrams and tables, in the author’s work (Druzhinina and Sumbatyan, 1990).

Helpful remarks We also tested some simple geometries for extremely high frequencies. The problem on comparison between approximate Kirchhoff’s solution and results of the direct numerical treatment requires a special investigation. The first problem we studied is related to a circular acoustically hard cylinder at ka = 20, N = 180. Here we can include also comparison with the solution predicted by the geometrical diffraction theory, which is given as √ πka p | sin ϕ/2|. (9.13) F (ϕ) = 2 The results of computation (see Druzhinina and Sumbatyan, 1990) show a perfect accuracy of the asymptotic solutions in the light, and poor accuracy in a shadow zone. Then we considered an example with higher frequency: ka = 50 to study further increase of the frequency parameter. To this end, to adequately describe the problem we needed to take N = 400. The 400 × 400 (complex-valued) algebraic system was solved by a direct Gauss method, which did not give us any reasonable result. The realization of SDM with 20 iterations and with acceleration after each 5 iterations decreases the discrepancy by 4 orders. The comparison with the asymptotic results demonstrates again that asymptotic theories in the shadow zone are of poor accuracy.

9.2. Galerkin Methods for Integral Equations of the First Kind with Weakly Singular Kernels As shown by the example of the problem considered in Section 8.7, the question of the stable solution of integral equations of the first kind depends upon the rate of asymptotic decrease of the eigenvalues (λn → 0, n → ∞) of the considered integral operator, as well as upon the asymptotic behavior of the coefficients in expansion of the right-hand side by eigenfunctions. If the kernel is positive and symmetric, then the spectrum {λ n } is positive and the eigenfunctions form a complete set of orthogonal functions. As follows from results of Hille and Tamarkin (1931), less regular kernels generate the spectrum {λ n }, where λn → 0 more slowly. Therefore, there is a good chance to construct a stable numerical method for the first-kind integral equations, whose kernel possesses a weak singularity. Let us start from the characteristic equation with the periodic logarithmic kernel Z π x – ξ dξ = f (x), |x| ≤ π , – u(ξ ) ln sin 2 –π (9.14) x – ξ . ∼ K0 u = f , K0 (x, ξ ) = – ln sin 2

If, for simplicity, f (x) is even, f (–x) = f (x), then it can easily be proved that the solution of equation (9.14) is even too. The spectrum of operator (9.14) can be directly defined on the basis of the following tabulated series (see Gradshteyn and Ryzhik, 1994): ∞ X x – ξ cos n(x – ξ ) = ln 2 sin , |x| ≤ π . (9.15) n 2 n=1 © 2005 by CRC Press LLC

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This shows that the set of eigenfunctions of the operator K 0 is {ϕn (x)} = {cos nx}, n = 0, 1, . . . , because for n = 0 we have (see Gradshteyn and Ryzhik, 1994) Z π Z π Z π ξ ξ x – ξ dξ = – ln sin dξ = –2 ln sin dξ K0 ϕ0 = – ln sin 2 2 2 –π 0 –π (9.16) Z π/2 π ln 2 = 2π ln 2, = –4 ln(sin ξ ) dξ = (–4) – 2 0

where we have used the following property of periodic functions: the value of the integral of any periodic function taken over the interval of the length equal to its period remains without change if we arbitrarily shift the interval of integration keeping it of the same length. For other n = 1, 2, . . . it directly follows from Eq. (9.15) that Z π Z π x – ξ x – ξ dξ = – dξ cos nξ ln sin K0 ϕn = – cos(nξ ) sin 2 2 –π –π Z π ∞ X 1 cos(nξ ) cos[m(x – ξ )] dξ =– m –π m=1 (9.17) Z ∞ X cos(nx) π cos(nξ ) cos(mξ ) dξ =– m –π m=1 π π n = 1, 2, . . . ; = cos(nx) = ϕn (x), n n hence {ϕn } are indeed eigenfunctions of the operator K0 , with the eigenvalues λ0 = 2π ln 2, λn = π/n, n = 1, 2, . . . . Now solvability of equation (9.14) depends only on analytic properties of the right-hand side f (x). If f (x) ∈ C2 (–π , π ), then Z π Z π 1 , (9.18) fn = f (x)ϕn (x) dx = f (x) cos πnx dx = O 3 n –π –π

as follows from the asymptotic estimates of Section 1.4. Then the exact solution to integral equation (9.14) can be explicitly expressed as a series in eigenfunctions ∞ ∞ X X 1 fn =O , (9.19) u(x) = un ϕn (x) = un cos(nx), un = λn n2 n=0 n=0

which is uniformly convergent over the integral x ∈ [–π , π ]. In the case of a full Fredholm integral equation of the first kind, whose periodic kernel contains a weak logarithmic singularity Z π x – ξ , (9.20) [K0 (x – ξ ) + K1 (x, ξ )] u(ξ ) dξ = f (x), |x| ≤ π , K0 (x – ξ ) = – ln sin 2 –π

where f (x) and K1 (x, ξ ) are even with respect to their arguments and K 1 (x, ξ ) is regular in some sense, we may seek a solution of equation (9.20) again in the form of expansion by eigenfunctions of the kernel K0 (x – ξ ): u(x) =

∞ X

um ϕm (x),

ϕm (x) = cos(mx).

(9.21)

m=0


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Then the standard Galerkin method (see, for example, Fletcher, 1984) imply the substitution of the representation (9.21) into Eq. (9.20): Z π ∞ X um [K0 (x – ξ ) + K1 (x, ξ )]ϕm (ξ ) dξ = f (x), |x| ≤ π , m=0 ∞ X m=0

–π

um λm ϕm (x) +

Z

π

–π

K1 (x, ξ )ϕm (ξ ) dξ = f (x),

(9.22)

|x| ≤ π ,

and the application of the scalar product of Eq. (9.22) with the functions from the same basis. As a result, we arrive at an infinite system of linear algebraic equations un +

∞ X

anm um = fn ,

n = 0, 1, 2, . . . ,

(9.23)

m=0

where

Z Z δn π π K1 (x, ξ )ϕn (x)ϕm (ξ ) dξ (n, m = 0, 1, . . . ), anm = λn –π –π Z δn π 1/2π , n = 0, f (x)ϕn (x) dx, n = 0, 1, . . . , δn = fn = 1/π , n = 1, 2, . . . . λn –π

(9.24)

Here we need to recall some classical result from the theory of regular infinite linear algebraic systems. P∞ DEFINITION 1. System (9.23) is called regular if m=0 |anm | < 1 for ∀n = 0, 1, . . . .

It is shown in Kantorovich (1958) that regular systems have a unique bounded P∞ and Krilov solution if |fn | ≤ K 1 – m=0 |anm | for all n = 0, 1, 2, . . . , and K is identical for all n. P∞ DEFINITION 2. System (9.23) is called completely regular if m=0 |anm | < θ < 1 for ∀n = 0, 1, . . . , where θ is the same constant for all n. It is proved in Kantorovich and Krilov (1958) that completely regular systems have always a unique bounded solution if fn are uniformly bounded. They can also be numerically solved by the truncation method, which is treated in the following way. Let us consider the finite-dimensional truncated system vn +

N X

anm vm = fn ,

n = 0, 1, . . . , N .

(9.25)

m=0

Then for all n = 0, 1, . . . , the solution vn of this system tends to un as N → ∞.

DEFINITION completely regular if for all n = 0, 1, 2, . . . P∞ 3. System (9.23) is called quasi P∞ we have m=0 |anm | < ∞ and ∃M such that m=0 |anm | < θ < 1 for all n ≥ M .

It is proved in Kantorovich and Krilov (1958) that such a system is uniquely solvable if the finite system (9.25), formed by the first M equations, is uniquely solvable, and if f n are uniformly bounded. Under this condition such infinite system can be numerically solved by the truncation method. Let us show that our system, obtained by the application of the Galerkin method to the first-kind Fredholm integral equation with a (weak) logarithmic singularity, is quasi completely regular if the regular part of the kernel is smooth enough. More precisely, this statement is valid if at least K (x, ξ ) ∈ C2 [(–π , π ) × (–π , π )] and f (x) ∈ C2 (–π , π ), since


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the asymptotic estimates of Section 1.4 give here: a nm = O (1/n m2 ), n, m → ∞, and P∞ fn = O (1/n), n → ∞. Indeed, with such an estimate m=0 |anm | = O (1/n2 ), which can be made less than θ , 0 < θ < 1, beginning from a certain n ≥ M . Besides, the quantities f n are uniformly bounded. From the above consideration we can finally conclude that a full Fredholm equation of the first kind (9.20) with a (weak) logarithmic singularity in the kernel can be efficiently solved by the application of the Galerkin method, which reduces the problem to a quasi completely regular infinite system of linear algebraic equations that can be solved by the truncation method. Let us pass now to equations of the first kind with logarithmic but not periodic kernels. The characteristic equation here is Z 1 – ln |x – ξ | u(ξ ) dξ = f (x), |x| ≤ 1, G0 u = f , G0 (x, ξ ) = – ln |x – ξ |, (9.26) –1

and the eigenfunctions can be constructed on the basis of the following tabulated integral (see Gradshteyn and Ryzhik, 1994): Z 1 π ln 2, n = 0, T (ξ ) √ n 2 ln |x – ξ | dξ = π (9.27) T (x), n = 1, 2, . . . , 1–ξ –1 n n

where Tn (x) = cos(n arccos x) is the Chebyshev polynomial of order n (see Abramowitz and Stegun, 1965). Hence, following our generalline, we√can state that the orthogonal 2 basis formed by the eigenfunctions is {ϕn (x)} = Tn (x)/ 1 – x , with the respective eigenvalues being equal to λn = π/n (n = 1, 2, . . . ), λ0 = π ln 2. Our statement is based on the general property of orthogonality of eigenfunctions generated by a self-adjoint operator. In our case this is given by another tabulated integral, which is the relation of orthogonality (δnm is Kronecker’s delta) Z 1 Tn (x)Tm (x) √ = εn δnm , ε0 = π/2, εn = π . (9.28) 1 – x2 –1 These relations (9.27) and (9.28) immediately give an exact representation of the solution to equation (9.26) in the form u(x) =

∞ X m=0

um ϕm (x),

Tm (x) , ϕm (x) = √ 1 – x2

(9.29)

which is in a complete agreement with the theoretical result established in Section 1.7 about the structure of the solution to Eq. (9.26). Actually, the latter contains a square root singularity at the ends of the interval (–1, 1). Substitution of the series (9.29) into (9.26), with the help of relation (9.27), leads to an explicit expression for the unknown coefficients u n : Z 1 Tn (x) fn dx. (9.30) , fn = f (x) √ un = εn λn 1 – x2 –1 Now convergence of the series (9.29) is connected with the question of the asymptotic estimate of fn as n → ∞. The latter depends on analytic properties of the right-hand side f (x). If this function is at least twice-differentiable, f (x) ∈ C 2 (–1, 1), then Z π fn = f˜(t) cos(nt) dt, x = cos t, f˜(t) = f (cos t). (9.31) 0


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If, for simplicity, we consider again only even right-hand sides, then Z 1 1 π ˜ f (t) cos(nt) dt = O , n → ∞, fn = 2 –π n3 due to the periodicity (cf. Section 1.4). For all that, as follows from Eq. (9.30), 1 , n → ∞, un = O n2

(9.32)

(9.33)

so for such functions f (x) the series is uniformly convergent on the interval (–1, 1). In the case of the full equation Z

1 –1

|x| ≤ 1,

[G0 (x – ξ ) + G1 (x, ξ )] u(ξ ) dξ = f (x),

(9.34)

where for simplicity we restrict the consideration only by even f (x) and even G 1 (x, ξ ), with respect to both its arguments, we may seek the unknown function u(x) again in the form of a series in eigenfunctions: u(x) =

∞ X

um ϕm (x),

m=0

Tm (x) . ϕm (x) = √ 1 – x2

(9.35)

Then the substitution of this representation into Eq. (9.34) and the application of the scalar product with ϕn (x) lead to a second-kind infinite linear algebraic system un +

∞ X

bnm um = fn ,

n = 0, 1, 2, . . . ,

(9.36)

m=0

where bnm

εn = λn

Z 1Z –1

1 –1

T (x) Tm (ξ ) √n √ G1 (x, ξ ) dξ , 1 – x2 1 – ξ 2

1 fn = λn

Z

1 –1

Tn (x) dx. (9.37) f (x) √ 1 – x2

By analogy to the case with periodic kernel it is directly proved that system (9.36) is quasi completely regular if f (x) ∈ C2 (–1, 1) and G1 (x, ξ ) ∈ C2 [(–1, 1) × (–1, 1)]. Therefore, the truncation method may be used to solve numerically this system.

Helpful remarks It should be noted that the method described here is an alternative to another very efficient collocation technique (see Voronin and Tsetsokho, 1981). Each of these methods has its own intrinsic merits and restrictions. Thus, the collocation technique needs to solve a finite-dimensional algebraic system, in contrast with the method discussed at the present section. However, convergence of this method for integral equations with the logarithmic singularity is a more complex problem. Here we need to solve formally an infinite algebraic system, which can be reduced again to a finite-dimensional system, within the framework of the truncation method. Convergence of such a process is substantiated in a more clear way.


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9.3. Integral Equations of the Physical Diffraction Theory in the Case of Nonconvex Obstacles In Section 2.6 we could see that in the case of short-wave diffraction Kirchhoff’s physical diffraction theory can asymptotically predict the value of the unknown pressure over the boundary surface (or contour, in the 2D problem). Here we spread some ideas of the physical diffraction theory, by applying boundary integral equations to the short-wave diffraction by obstacle with an arbitrary smooth boundary. The background of our approach is the specific feature of the incident wave interaction with convex and concave parts of the boundary. Essentially, acoustic rays incident on convex parts of the boundary cannot participate in repeated re-reflections. Vice versa, the rays incident on any concave part can re-reflect only between points of this part, and never fall to any convex part of the boundary. We show all ideas on the 2D case, but the 3D case can be treated just in the same way. Let a plane incident acoustic wave be incident on an obstacle with a smooth boundary contour l (see Fig. 9.1).

D

q C

l1

B

l2

l1

l2

x2

A

x1 Figure 9.1. Incidence of the plane acoustic wave to a nonconvex obstacle We formulate the problem simultaneously both for scalar and elastic problems. If the boundary contour is free of load, then (to be more specific) in the scalar case of acoustically soft boundary we have p|l = 0, and in the elastic case: T|l = (T1 , T2 )|l = 0, where p is the acoustic pressure in the scalar problem, and T is the stress vector in the elastic problem. Then, within the framework of the direct method the boundary integral equation in the scalar problem is (see Chapter 2) i 4

Z

H0(1) (kr )g (y ) dly = p0 (x) = eik(q⋅x) , l ∂p , r = |x – y |, x ∈ l, g (y ) = ∂ny l

(9.38)

where k is the wave number, q is the unit vector which determines direction of the incident wave, and H0(1) is the Hankel function.


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In elastic medium, instead of one equation, we have the system of two equations: Z (1) u1 (x) – 2 P1 (x, y )u1 (y ) + P2(1) (x, y )u2 (y ) dly = 2u◦1 (x), Zl (2) P1 (x, y )u1 (y ) + P2(2) (x, y )u2 (y ) dly = 2u◦2 (x), x ∈ l, u2 (x) – 2 l

∂ (k) U (x, y ) + λ ny divy U(k) (x, y ) P(k) (x, y ) = 2µ ∂ny + µ ny × roty U(k) (x, y ) , ∂ 2 (1) i (1) (1) (k) 2 ks δkj H0 (ks r ) – H0 (kp r ) – H0 (ks r ) , Uj (x, y ) = 4µks2 ∂yk ∂yj u◦ (x) = q eikp (q⋅x) , r = |x – y |, k , j = 1, 2,

(9.39)

where δkj is Kronecker’s delta, kp and ks are the longitudinal and transverse wave numbers, respectively, λ and µ are elastic moduli. Here, to be more specific, we consider incidence of the longitudinal wave. Let us construct a convex hull of the boundary (see, for example, Preparata and Shamos, 1985; and our Section 8.9), which is the minimum convex contour containing the given contour l. This implies that we need to draw tangent straight lines to the contour l, which separate convex (l1 ) and nonconvex (l2 ) parts of the boundary. Further, if we draw tangent lines parallel to the vector q, then we separate the shadow zone AB of the total boundary. The latter is situated on its back side. It is obvious that the points A and B belong to convex parts of the contour l1 . Let L be a characteristic size of the obstacle. If in the asymptotic sense, kL 1, kp L 1, ks L 1, then the solution of integral equations in the shadow zone is equal to zero. So BIE (9.38) and (9.39) are asymptotically reduced to the arc ACDB . In what follows we mean by l1 and l2 only those parts of the contour, which lie on the front part ACDB of the boundary contour. The behavior of the solution of the basic integral equations over the convex parts l 1 differs qualitatively from its behavior on the nonconvex parts l 2 . Actually, on l1 (arcs BD and AC in Fig. 9.1) there is no re-reflections of acoustic rays. By contrast, ray re-reflections play an essential role on l2 . Therefore, at high frequencies the solution of equations (9.38) and (9.39) on the convex parts BD and AC is asymptotically defined as in the case of strictly convex obstacle. This implies for Eq. (9.38) that the leading asymptotic term is ∂p0 = 2ik (q ⋅ ny ) eik(q⋅y) , y ∈ l1 , (9.40) g (y ) = 2 ∂ny in the scalar problem. In the case of elastic medium in the neighborhood of every point on the convex parts l 1 , the solution of system (9.39) coincides with the known solution of the problem about reflection of the plane wave from a free boundary of the corresponding elastic half-plane. Let us relate to every point y ∈ l1 basis unit vectors τy , ny that form a right Cartesian coordinate system where ny is an outward normal. Let ty be a unit vector that determines the direction of the transverse wave reflected from the boundary l 1 at the point y . Then, according to Brekhovskikh (1980), ks uτ (y ) = –(q ⋅ τy ) 1 + Vpp + (ty ⋅ ny )Vps eikp (q⋅y) , kp (9.41) ks un (y ) = –(q ⋅ ny ) 1 – Vpp – (ty ⋅ τy )Vps eikp (q⋅y) . kp © 2005 by CRC Press LLC

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Here Vpp and Vps are coefficients of reflection from the boundary at the point y ∈ l 1 , for longitudinal and transverse wave, respectively. As a result, components of the displacement vector u = {u 1 , u2 } on the convex parts l1 can be found without solving system (9.39), in the following form: u1 (y ) = n1 un (y ) + n2 uτ (y ),

u2 (y ) = n2 un (y ) – n1 uτ (y ),

(9.42)

where n1 and n2 are Cartesian components of the unit outward normal n y = (n1 , n2 ). As follows from the previous analysis, now we only need to define the value of the unknown functions in Eqs. (9.38), (9.39) on the nonconvex parts l 2 . To this end, we keep on the left-hand sides of the corresponding BIE only integrals over y ∈ l 2 . For y ∈ l1 the value of the unknown functions are already determined by formulas (9.40)–(9.42), and the corresponding integrals over these contours may be carried over the right-hand sides as some already known functions. Therefore, by putting x ∈ l 2 , we can reduce in the considered high-frequency case the studied BIEs to some integral equations over l 2 . Such a reducing essentially decreases required calculation time, which is of the order of M 3 if M designates the number of the collocation nodes in any quadrature formula. Let us write out, for example, an equation to which we come in the scalar case, after this reducing: Z i H (1) (k |x – y |)g (y ) dly 4 l2 0 Z (9.43) k (q ⋅ ny ) eik(q⋅y) H0(1) (k |x – y |) dly , x ∈ l2 . = eik(q⋅x) + 2 l1 Further simplification is possible in the same way. If on the illuminated part of the boundary contour there are several arcs of the type l 2 , then the key essence of the diffraction process is as follows: acoustic sources, as boundary values of the corresponding unknown functions, lying on a certain such nonconvex part, in the asymptotic sense do not influence the values of wave fields on other such nonconvex parts. By other words, acoustic rays reflected from one nonconvex part cannot hit other nonconvex parts of the boundary. Therefore, if there are J arcs l2j , j = 1, . . . , J , then it suffices to solve equation (9.43) on each contour l2j separately. At last, additional simplification may be achieved if we neglect the influence of acoustic sources lying on convex parts l1 to nonconvex ones l2 . Such an approach is quite natural since the rays reflected, for instance, from the arc AC cannot hit the arc CD neither in scalar nor in elastic case. For instance, Eq. (9.43) within the framework of such an approach can be simplified and reduced to a set of more simple independent equations Z i H (1) (k |x – y |) g (y ) dly = eik(q⋅x) , x ∈ l2j (j = 1, . . . , J ). (9.44) 4 l2j 0 To demonstrate efficiency of this approach, we consider the calculation of the backscattered diagram, which is of significant interest for many applications. If the unknown functions are defined from Eqs. (9.38), (9.39), then the back-scattered diagram is determined from the following integral representation: Z i eik(q⋅y) g (y ) dly , q = {– cos α, – sin α}, (9.45) A(α) = – 4 l 1 A(α) = 4

kp ks

4 Z

l

e

ikp (q⋅y)

λ uy ⋅ 2(q ⋅ ny ) q + ny µ

dly .

(9.46)


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Below we demonstrate the proposed method for obstacles both of convex and nonconvex shape. In all figures line 1 refers to the exact numerical solution of the full integral equation and line 2 refers to the calculations within the framework of our approximate approach. Figure 9.2 demonstrates the back-scattered diagram A(α), (9.46), for the 2D diffraction problem about elliptic void in an elastic medium. We used the values k p a = 10, a/b = 3, ks /kp = cp /cs = 5.85/3.23. For this convex obstacle implementation time required to construct lines 1 and 2 differ by 120 times. The number of nodes here was taken N = 120.

|A| 4

a

2

1 2

0

0°

30°

60°

a

Figure 9.2. Comparison of exact (line 1) and proposed (line 2) methods for the calculation of diffracted waves for an elliptic obstacle Figure 9.3 shows the back-scattered diagram for an obstacle in the form of a bagel, which is represented in the Cartesian coordinate system by the following parametric equations: kx(t) =

4 X m=0

a0 = 2.76; b1 = 8.48;

4 X

bm sin(mt),

0 ≤ t < 2π ,

a1 = 3.38; a2 = 6.15; a3 = –1.12; b2 = 2.35; b3 = 1.30; b4 = –1.74.

a4 = –0.36;

am cos(mt),

ky (t) =

m=1

(9.47)

The processor time required for calculating curve 2 is 8 times less than that required for constructing line 1. For all that, N = 64. There is reflected in Fig. 9.4 the back-scattered diagram for a three-leaf rose considered as a void in elastic medium. The equation describing boundary contour in the polar coordinate system is here kp ρ(ϕ) = 5 (2 + cos 3ϕ). Physical parameters have the same value as in the case shown in Fig. 9.2. We took here N = 150. Calculation time between lines 1 and 2 differs by 5 times. Due to evident symmetry of all considered obstacles we show all results only for a certain part of variation of the angle α ∈ (0◦ , 360◦ ). In order to evaluate more completely the accuracy of the presented results, we note that the characteristic size of the considered obstacles is around 3λ to 4λ, where λ is the wavelength (in elastic case we put λ = λp , the length of the longitudinal wave).

Helpful remarks It should be noted that some alternative approaches to extend the Boundary Integral Equation method to higher frequencies are proposed by Thiele and Newhouse (1975), as well as by Tobocman (1986, 1987).


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|A| 6

a

4

2

2

0

1

0°

60°

120°

a

Figure 9.3. Comparison of exact (line 1) and proposed (line 2) methods for the calculation of diffracted waves for a bagel obstacle

a

|A| 6

4 2

2

0

1

0°

15°

30°

45°

a

Figure 9.4. Comparison of exact (line 1) and proposed (line 2) methods for the calculation of diffracted waves for a three-leaf rose

9.4. Numerical Methods in Singular Integral Equations with the Cauchy-Type Kernel This section follows in the main the classical results of Belotserkovsky and Lifanov (1993). Let us start to expound the methods suitable for numerical treatment of singular integrals and integral equations from a quadrature formula for the Cauchy-type singular integrals. Let us introduce the so-called canonical partition of the interval (a, b) with its subdivision to n+1 equal subintervals of the length h = (b – a)/(n + 1). This implies two sets of nodes: {tj } with a = t0 , t1 , t2 , . . . , tn , tn+1 = b, tj = a + jh, j = 0, 1, . . . , n + 1, and {xi } as the


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central points of each subinterval (ti , ti+1 ), i = 0, 1, . . . , n, so that xi = a + i + 1/2 h, i = 0, . . . , n. It is proved that if g (x) ∈ C1 (a, b), then Z b n X g (tj ) ln n g (t) dt =h +O , n → ∞ ∼ h → 0. (9.48) xi – t j n(xi – a)(b – xi ) a xi – t j=1

Now let us now study a singular integral equation with a characteristic kernel Z b g (t) dt = f (x), x ∈ (a, b) , f (x) ∈ C1 (a, b) , x–t a

(9.49)

and we recall some classical results related to this equation (see, for example, Gakhov, 1966; Muskhelishvili, 1965). For equation (9.49) there exist three different classes of its solution: 1) Solution unbounded at both ends of the interval (a, b). Such a solution is not unique, and as in the theory of ordinary differential equations of the first order this is defined to within an arbitrary constant C . 2) Solution bounded at any of the two ends of the integration interval. Such a solution is unique. 3) With some additional condition for the right-hand side (i.e., not for every function f (x)) there can exist a solution bounded at both ends of the interval. We will be interested here only in case 1), where the exact analytical solution to equation (9.49) is given explicitly as Z b√ (x – a) (b – x) 1 f (x) dx . (9.50) πC – g (t) = 2 √ t–x π (t – a) (b – t) a Note that arbitrary constant C is related with the integral of the unknown function g (x): Z b C= g (x) dx. (9.51) a

Our further strategy is to construct a direct numerical collocation technique to solve characteristic Eq. (9.49) for arbitrary right-hand side, so that the constructed solution is a correct approximation for exact analytical solution (9.50). Then the method developed for the characteristic case will allow us to spread these results to a full singular integral equation where the analytical solution is not known. To this end, we apply approximation (9.48) for singular integral: Z b n X g (tj ) g (t) dt ≈h , (9.52) xi – t j a xi – t j=1 so that we arrive at some linear algebraic system. Let us prove that the solution of the system  X n g (tj )   = f (xi ), i = 1, . . . , n – 1, h    j=1 xi – tj (9.53) n X    g (tj ) = C ,  h j=1

which is indeed an n × n linear algebraic system with respect to unknown values g (t j ), j = 1, 2, . . . , n, gives a correct approximation to the exact solution (9.51) of equation (9.49), in the sense that if t = tl ∈ (a, b) is fixed, then the difference between the solution g (t l ) of


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system (9.53) and the analytical solution (9.50) tends to zero as n → ∞. The principal determinant of system (9.53) is 1 1 ... x1 – t 1 x1 – tn ... ... ... n . (9.54) D = h ∆, ∆= 1 1 ... xn–1 – t1 xn–1 – tn 1 ... 1 It is proved in Belotserkovsky and Lifanov (1993) that QQ QQ tq – t p xp – x q q

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